Identifier
stringlengths 37
82
| Paragraph
stringlengths 1.95k
9.23k
| Citation Text
list | Functions Text
list | Functions Label
list | Citation Start End
list | Functions Start End
list |
|---|---|---|---|---|---|---|
2022MNRAS.510.5302I__Fields_et_al._2014_Instance_1
|
Lithium is the only metal element produced during the Big Bang nucleosynthesis (BBN) due to the lack of stable nuclei with mass number eight (Fields, Molaro & Sarkar 2014). The element abundances predicted by the standard BBN theory for the baryonic density coming from the Planck mission agree well with those observed, except for 7Li (Fields 2011; Coc, Uzan & Vangioni 2014). Indeed, the abundance of lithium measured in the low-metallicity Galactic halo stars is A(7Li) = log[N(7Li)/N(H)] + 12 = 2.25 (Spite & Spite 1982; Sbordone et al. 2010; Bonifacio et al. 2015), which is ∼3 times below the estimate of the standard cosmological model A(7Li) = 2.72 ± 0.06 (Cyburt et al. 2016). The latter value depends on the baryon-to-photons ratio $\eta = \frac{N_b}{N_{\gamma }} \propto \Omega _B h^2$, with Ωb is the cosmological baryon density and h is the dimensionless hubble parameter (Planck Collaboration XIII 2016). This problem is also known as the Cosmological Lithium problem (Fields et al. 2014). A possible solution can be ascribed to convective diffusion in the pre-main-sequence phase as well as during the lifetime of these halo stars (Fu et al. 2015) or to new physics beyond the standard model. On the other hand, the young stellar populations in our Galaxy show Li-abundances four times greater than the SBBN estimate and more than one order of magnitude greater than the halo stars (Spite 1990; Lambert & Reddy 2004; Lodders, Palme & Gail 2009; Ramírez et al. 2012; Fu et al. 2018). The evidence of a growth requires the existence of additional lithium factories. In the last decades several astrophysical Li sources have been proposed, like Galactic cosmic-rays, AGB stars, low-mass Carbon stars, type II supernovae, and classical novae (D’Antona & Matteucci 1991; Romano et al. 1999; Prantzos 2012; Matteucci 2021). The recent detection in the outburst spectra of classical novae of 7Li and 7Be ii, an isotope whose unique decay channel is into lithium through electron capture, have confirmed these objects as Li producers. The corresponding yields inferred have placed nova explosions as the main lithium factories in the Galaxy. The time-scales involved also match, as shown by detailed Galactic chemical evolution (Izzo et al. 2015; Tajitsu et al. 2015; Molaro et al. 2016; Izzo et al. 2018; Cescutti & Molaro 2019; Grisoni et al. 2019; Molaro et al. 2020a; Matteucci 2021).
|
[
"Fields et al. 2014"
] |
[
"This problem is also known as the Cosmological Lithium problem"
] |
[
"Background"
] |
[
[
984,
1002
]
] |
[
[
920,
982
]
] |
2016ApJ...824...92H__Douglas_&_Ballai_2007_Instance_1
|
A couple of years after the closure of the solar flare myth debate, Dere et al. (1997) and Thompson et al. (1998) reported on the presence of wavefronts observed on the solar disk that propagated away from the site of the flare. They were originally termed “EIT waves” (Thompson et al. 1999) after the instrument that first detected them (EIT on board the Solar and Heliospheric Observatory (SOHO) Delaboudiniére et al. 1995) and have now been intensely studied for over 15 years. Over that time they have been assigned different terms, such as flare waves (e.g., Thompson et al. 2000; Warmuth et al. 2003), extreme-UV (EUV) waves (e.g., Patsourakos & Vourlidas 2012; Muhr et al. 2014), coronal waves (e.g., Douglas & Ballai 2007; Webb & Howard 2012), and coronal bright fronts (e.g., Gallagher & Long 2011, and references therein). While there are differences in the precise definitions of these terms, broadly speaking they describe the same phenomenon: a bright front propagating away from a point-like origin on the solar disk. In many cases the origin appears to be from the site of a solar flare associated with a CME (e.g., Thompson 1999; Thompson & Myers 2009; Wills-Davey & Attrill 2009; Nitta et al. 2013), but some workers have reported a stronger correlation of wave occurrence with CMEs rather than flares (e.g., Wills-Davey & Attrill 2009, and references therein). For simplicity, henceforth in this paper we refer to them as EIT waves although we do not use the EIT instrument for any measurements in our study. We define EIT waves with the criteria described in Section 4. Nitta et al. (2013) provide a recent review of the various physical narratives that have been assigned to EIT waves, such as the coronal manifestations of Moreton waves (e.g., Thompson 1999; Warmuth et al. 2001, 2002), the base of expanding overhead CME magnetic structures (e.g., Delanée & Aulanier 1999), and solitons (Wills-Davey et al. 2007). Other reviews on EIT waves can be found by Wills-Davey & Attrill (2009), Gallagher & Long (2011), and Patsourakos & Vourlidas (2012).
|
[
"Douglas & Ballai 2007"
] |
[
"Over that time they have been assigned different terms, such as",
"coronal waves (e.g.,"
] |
[
"Background",
"Background"
] |
[
[
708,
729
]
] |
[
[
481,
544
],
[
687,
707
]
] |
2015ApJ...815..129S__Shen_et_al._2011_Instance_1
|
The mass accretion onto the black hole is important for a better understanding of AGN evolution. The Eddington ratio, the ratio between the AGN bolometric luminosity and the Eddington luminosity (Lbol/LEdd), provides insight into the black hole growth because the bolometric luminosity reflects the mass accretion rate. We show AGN bolometric luminosity versus black hole mass for our sample of broad-line AGNs in the different redshift bins in the left panel of Figure 5. The different X-ray surveys are shown with different symbols as labeled. The dotted reference lines indicate constant Eddington ratios of 1, 0.1, 0.01, and 0.001. Our sample of broad-line AGNs covers the black hole mass range 7.0 log MBH/M⊙ 9.5 and the bolometric luminosity range 43 log Lbol 47 with a wide dispersion in the Eddington ratio distribution. For comparison, we show published observations in the same redshift range from the literature in the right panel of Figure 5 (Gavignaud et al. 2008; Merloni et al. 2010; Shen et al. 2011; Nobuta et al. 2012; Matsuoka et al. 2013). The SDSS quasar sample (gray points; Shen et al. 2011) is limited to the high-mass and high-luminosity regime because the SDSS detection limit corresponds to a luminosity of log Lbol ∼ 46 at z ∼ 1. Compared to the SDSS quasar sample, our sample of broad-line AGNs shows a wider dispersion in the black hole mass, AGN bolometric luminosity, and Eddington ratio distribution, consistent with previous studies on deep AGN samples (Gavignaud et al. 2008; Merloni et al. 2010; Nobuta et al. 2012; Matsuoka et al. 2013), which fill in the low-mass and low-luminosity region. The figure shows contours at the 1σ level, together with the literature data, except the SDSS quasar sample. The figure also reveals that only a small number of AGNs exceed the Eddington limit by a small amount. AGNs with similar black hole masses show a broad range of bolometric luminosities spanning about two orders of magnitude, indicating that the accretion rate of black holes is widely distributed. This suggests that the AGN cosmic downsizing phenomenon could be explained by some more-massive black holes with low accretion rates, which are relatively fainter than less-massive black holes with efficient accretion. Lusso et al. (2012) suggest that AGNs show higher Eddington ratios at higher redshift at any given MBH, and the Eddington ratio increases with bolometric luminosity. We confirm that there is a tendency for low-luminosity AGNs (log Lbol ≲ 45.5) with less-massive black holes (log MBH/M⊙ ≲ 8) to have lower Eddington ratios than high-luminosity AGNs (log Lbol ≳ 45.5) with massive black holes (log MBH/M⊙ ≳ 8), consistent with Lusso et al. (2012). It is important to note that, when comparing with results in the literature, one should take into account the different methods of spectral line fitting and correction for bolometric luminosities. Nevertheless, they show similar distributions of the accretion rate of black holes over a wide range, consistent with previous studies.
|
[
"Shen et al. 2011"
] |
[
"For comparison, we show published observations in the same redshift range from the literature in the right panel of Figure 5"
] |
[
"Uses"
] |
[
[
1003,
1019
]
] |
[
[
833,
957
]
] |
2018AandA...615A.148D__Weidner_et_al._(2010)_Instance_3
|
The last column in Table 1 reports the number of OB stars minus the “diffuse” population estimated from their density in the Reference field (22.5 stars per square degree): as is immediately seen, the M-star statistics is much larger than the OB star statistics. This can hardly be considered surprising, if an ordinary IMF (e.g., that from Weidner et al. 2010) is assumed for the star-formation region. In Fig. 19 we show the density ratio between M and OB stars, which provides a consistency test between our results and a plausible IMF: this ratio varies however by a large factor, close to 20, among our subregions. This might reflect differences in the respective IMFs, but also differences in completeness among the stellar samples considered for the various regions. We first note that the ratio between M and OB stars in NGC 6231 is dramatically lower than anywhere else in Sco OB1. We can indeed expect that M stars are detected less efficiently in the inner parts of NGC 6231, where the density of bright stars is very large, and their diffuse glare raises the limiting magnitude locally. As already discussed above in Sect. 4.1, this causes our sample of M stars in NGC 6231 to be highly incomplete. Moreover, we determined above that NGC 6231 is significantly more extincted, by almost half a magnitude in V, than Tr 24, and this implies a higher minimum detectable mass among NGC 6231 M stars compared to Tr 24 (see the MDA diagrams in Fig. 5); this effect reduces the completeness of the M-star sample in NGC 6231 more than in Tr 24. If Tr 24 is also slightly younger than NGC 6231, as we argue below, our M-stars in Tr 24 will reach down to lower masses than in NGC 6231, with a steep increase in the detected M-star population: adopting the IMF from Weidner et al. (2010), the predicted number of cluster M stars doubles considering the mass interval 0.25–0.5 M⊙ rather than 0.35–0.5 M⊙. If Tr 24 is younger than NGC 6231, moreover, its stars in the mass range 2.5–3 M⊙ might not have yet reached their ZAMS position as B stars, and therefore would not be counted among OB stars; this would further raisethe M/OB star ratio there by up to 30%. Therefore, the proportions of both M and OB stars that are detected in a young cluster will depend on their age and extinction, in accordance with the MDA diagrams, even for a fixed, spatially uniform photometric sensitivity. We estimated using the Weidner et al. (2010) IMF the expected range for the observed M/OB number ratio. Siess et al. (2000) predict that the latest-type B stars have a mass of ~ 3.5 M⊙ at 2 Myr, and ~ 2.2 M⊙ at 10 Myr, that is, in the range of ages expected for Sco OB1 clusters. The MDA diagrams of Fig. 5 predict that the lowest-mass stars we are able to detect using the available Sco-OB1 data have ~ 0.2 M⊙, even assuming the most favorable (and unlikely) circumstances of an age less than 2 Myr and negligible reddening. The extreme values found for the M/OB ratio are then ~ 3.8 for a minimum M-star mass as high as 0.35 M⊙ and an old age of 10 Myr, and ~ 20 for a minimum M-star massas low as 0.2 M⊙ and age of 2 Myr. These extremes are also shown as horizontal lines in Fig. 19. We note that the M/OB ratio in NGC 6231 falls well within this range; however, both Tr 24 regions are significantly richer of M stars than expected, by more than a factor of two and well above (statistical) errors. If true, then paradoxically this part of the OB association would form preferentially lower-mass stars. Of course, more detailed studies are needed to confirm this result. In the G345.45+1.50 region the M/OB ratio is highest, and far above predictions from the IMF: we may tentatively explain this since this region is very young, and some of its most massive members, like IRAS 16562-3959, are still in formation, thus decreasing the number of optically revealed OB stars. The lowest M/OB ratio in NGC 6231 is unlikely to be real, since as discussed above our M-star sample in this densest subregion is likely incomplete.
|
[
"Weidner et al. (2010)"
] |
[
"We estimated using the",
"IMF the expected range for the observed M/OB number ratio."
] |
[
"Uses",
"Uses"
] |
[
[
2409,
2430
]
] |
[
[
2386,
2408
],
[
2431,
2489
]
] |
2021MNRAS.503..815V__Schmalzing,_Buchert_&_Kerscher_1995_Instance_1
|
Level crossing statistics is a pioneering approach for characterizing stochastic processes introduced by S. O. Rice (Rice 1944, 1945). Up-, down-, and conditional crossing statistics are modifications to the primary definition of level crossing (Bardeen et al. 1986; Bond & Efstathiou 1987; Ryden 1988; Ryden et al. 1989; Matsubara 1996, 2003; Brill 2000; Shahbazi et al. 2003; Movahed & Khosravi 2011; Ghasemi Nezhadhaghighi et al. 2017). Minkowski functionals, which are also closely related to the crossing statistics, provide 1 + D functionals to quantify the morphology in D dimensions (Hadwiger 2013) and have been utilized for cosmological random fields (Mecke, Bucheri & Wagner 1994; Schmalzing, Buchert & Kerscher 1995; Schmalzing & Górski 1998; Matsubara 2003, 2010; Hikage, Komatsu & Matsubara 2006; Codis et al. 2013; Ling et al. 2015; Fang, Li & Zhao 2017). A number of critical sets including peaks (hills), troughs (lakes), saddles, voids, skeleton, genus, and Euler characteristics, are more popular in cosmology for different purposes and they have been fully explored for Gaussian stochastic fields. Some extensions for non-Gaussian and anisotropic conditions have been done in some research (Matsubara 2003; Pogosyan et al. 2009; Pogosyan, Pichon & Gay 2011; Gay et al. 2012; Codis et al. 2013). More recently, Betti numbers, Euler characteristics, and Minkowski functionals for a set of cosmological 3D fields have been examined extensively (Pranav et al. 2019). The scaling approach for investigating cosmological stochastic fields has been discussed by Borgani (1995), Movahed et al. (2011). Standard estimators like three- and four-point functions in real space, bispectra and trispectra in harmonic space, multiscaling methods such as wavelets (Planck Collaboration 2014b, 2016d; and references therein), and regeneration of stochastic processes based on the Fokker–Planck equation (Ghasemi et al. 2006) have also been considered.
|
[
"Schmalzing, Buchert & Kerscher 1995"
] |
[
"Minkowski functionals, which are also closely related to the crossing statistics",
"and have been utilized for cosmological random fields"
] |
[
"Background",
"Background"
] |
[
[
692,
727
]
] |
[
[
440,
520
],
[
607,
660
]
] |
2017AandA...602A..75R__Kaneko_&_Yokoyama_(2015)_Instance_2
|
We can see how this is consistent with the development of small scales via phase mixing by introducing local wavenumbers for the variation with α and β, (3)\begin{eqnarray} \xi \propto \exp {\rm i}\left[ \int \kappa_\alpha {\rm d}\alpha + \int \kappa_\beta {\rm d}\beta \right]\!. \label{number3} \end{eqnarray}ξ∝expi∫καdα+∫κβdβ.Here κα and κβ are the wavenumbers in α and β and have units that are the inverse of the units of their respective coordinates. These wavenumbers should be distinguished from the perpendicular components of the usual wave vector k, which has units of 1/length. The different wavenumbers may be related through the scale factors (h) that relate elemental coordinate increments to physical distances: dr = eαhαdα + eβhβdβ + eγhγdγ, where eα is a unit vector in the α direction, etc. In this notation ∇⊥ = (eα/hα)∂/∂α + (eβ/hβ)∂/∂β, and noting that ∇⊥ξ ≈ ik⊥ξ, Eq. (3) yields \begin{eqnarray} \bdelp\xi &=& {\rm i} \vec{k}_\perp \xi = {\rm i} \left( \frac{\vec{e}_\alpha}{h_\alpha}\frac{ \partial}{\partial\alpha} +\frac{\vec{e}_\beta}{h_ \beta}\frac{ \partial}{\partial \beta}\right)\xi \\ &\equiv& {\rm i}\left( \frac{\vec{e}_\alpha}{h_\alpha}\kappa_\alpha +\frac{\vec{e}_\beta}{h_ \beta}\kappa_\beta\right)\xi. \label{number4} \end{eqnarray}∇⊥ξ=ik⊥ξ=ieαhα∂∂α+eβhβ∂∂βξ≡Equating components of the second and fourth expressions in Eq. (5) gives the expected relations between the various wavenumbers, (6)\begin{eqnarray} k_\alpha = \kappa_\alpha /h_\alpha, \qquad k_\beta = \kappa_ \beta /h_ \beta. \label{number5} \end{eqnarray}kα=κα/hα, kβ=κβ/hβ.Equations (2) and (5) give a direct and elegant expression for the perpendicular wave vector as (7)\begin{eqnarray} \vec{k}_\perp \approx -(\bdel \omega_{\rm c}) t, \label{number6} \end{eqnarray}k⊥≈−(∇ωc)t,which is a generalisation to three dimensions of the results of Mann et al. (1995), (Wright et al. 1999) and Kaneko & Yokoyama (2015) for lower dimensional systems, which developed phase mixing in only one perpendicular coordinate. The above expression allows phase mixing in both perpendicular directions, giving physical phase mixing lengths (or wavelengths) in the α and β directions of (8)\begin{eqnarray} L_{{\rm ph}\alpha} = \frac{2\pi}{|k_\alpha |} \equiv \frac{2\pi h_\alpha}{|\partial \omega_{\rm c}/\partial\alpha |t}, \qquad L_{{\rm ph}\beta} = \frac{2\pi}{|k_ \beta |} \equiv \frac{2\pi h_\beta}{|\partial \omega_{\rm c}/\partial \beta |t}\cdot \label{number7} \end{eqnarray}Lphα=2π|kα|≡2πhα|∂ωc/∂α|t, Lphβ=2π|kβ|≡2πhβ|∂ωc/∂β|t·If the phase mixing lengths are expressed in the same units as α and β, rather than physical length as in Eq. (8), slightly simpler expressions are found, i.e. (9)\begin{eqnarray} \ell_{{\rm ph}\alpha} = \frac{2\pi}{|\kappa_\alpha |} \equiv \frac{2\pi}{|\partial \omega_{\rm c}/\partial\alpha |t}, \qquad \ell_{{\rm ph}\beta} = \frac{2\pi}{|\kappa_ \beta |} \equiv \frac{2\pi}{|\partial \omega_{\rm c}/\partial \beta |t}\cdot \label{number8} \end{eqnarray}ℓphα=2π|κα|≡2π|∂ωc/∂α|t, ℓphβ=2π|κβ|≡2π|∂ωc/∂β|t·The development of the phase mixing length can be pictured simply as the tendency for each field line to oscillate with its own natural frequency. Even if all the field lines start to oscillate with the same phase, they soon drift out of phase with one another as time passes. Not only does the phase mixing process generate perpendicular scales, but points of constant phase can be seen to move across field lines. This phase motion has been seen in magnetospheric data of Alfvén waves (see the review by Wright & Mann 2006) and the simulations of coronal oscillations by Kaneko & Yokoyama (2015). These studies note that the direction of motion is related to the spatial variation of ωc. The results of these papers for the perpendicular phase velocity in physical space generalise to Vph = ωc/k⊥, giving the components (10)\begin{eqnarray} V_{{\rm ph}\alpha} = \frac{-\omega_{\rm c} h_\alpha}{(\partial \omega_{\rm c}/\partial\alpha )t}, \qquad V_{{\rm ph}\beta} = \frac{-\omega_{\rm c} h_ \beta}{(\partial \omega_{\rm c}/\partial \beta )t}, \qquad \label{number9} \end{eqnarray}Vphα=−ωchα(∂ωc/∂α)t, Vphβ=−ωchβ(∂ωc/∂β)t, If the excitation occurred at a time ti, the subsequent properties are found by replacing t with t−ti in the above formulae.
|
[
"Kaneko & Yokoyama (2015)"
] |
[
"This phase motion has been seen in magnetospheric data of Alfvén waves",
"and the simulations of coronal oscillations by",
"These studies note that the direction of motion is related to the spatial variation of ωc."
] |
[
"Uses",
"Uses",
"Uses"
] |
[
[
3614,
3638
]
] |
[
[
3457,
3527
],
[
3567,
3613
],
[
3640,
3730
]
] |
2021AandA...646A.142R__Nelson_&_Melrose_1985_Instance_1
|
CME-driven shocks accelerate not just protons, but also the electrons in the solar corona (Holman & Pesses 1983; Schlickeiser 1984; Kirk 1994; Mann et al. 1995, 2001; Mann & Klassen 2005). These accelerated electron beams can be observed as type II bursts in the solar radio radiation in the metric wave range (Wild & McCready 1950; Uchida 1960). Type II bursts require electrons escaping from the shock front, and the lack of these bursts implicates the absence of accelerated electrons as type II bursts occur when 0.2–10 KeV electrons are accelerated in the shock front (see, e.g., Bale et al. 1999; Knock et al. 2001; Mann & Klassen 2005). Energetic electrons are unstable to Langmuir waves, thus they are converted into radio emission at the local plasma frequency and its harmonic (see Nelson & Melrose 1985). Therefore, type II radio bursts hold crucial information of both the shock and the surrounding ambient medium in which the CME-driven shock propagates (Gopalswamy et al. 2008a). Although almost every large SEP event is accompanied by a type II radio burst (Gopalswamy 2003; Cliver et al. 2004) that indicates CME-driven particle acceleration (Gosling 1993; Reames 1999), we have ten events in our sample that lack a type II burst: 14 August 2010, 03 August 2011, 04 March 2012, 26 May 2012, 27 May 2012, 14 June 2012, 08 September 2012, 14 December 2012, 21 April 2013, and 06 November 2013. Of the ten events, two originate in the west, three in the east, and five at the disk center. On average, these events have a maximum velocity of about 1000 km s−1 and maximum Mach number of about 0.94. The protons accelerated by these events have peak fluxes of about 23.2 cm−2 s−1 sr−1 in the > 10 MeV band, making these events the slowest and weakest SEPs in the sample. As the propagation of radio bursts does not depend on the magnetic connectivity, a possible explanation for their absence could be that the path of the radio burst did not coincide with the instrument on board the satellite or that the detection of the waves was below the range of the radio instrument, hence missing the signature. Detailed investigation is required to understand the absence of DH type II radio bursts in these events.
|
[
"Nelson & Melrose 1985"
] |
[
"Energetic electrons are unstable to Langmuir waves, thus they are converted into radio emission at the local plasma frequency and its harmonic (see"
] |
[
"Background"
] |
[
[
792,
813
]
] |
[
[
644,
791
]
] |
2019AandA...626A..49P__Springel_et_al._(2005)_Instance_1
|
To provide enough galaxies to adequately train a neural network, EAGLE galaxies from the simulation snapshots with a redshift of less than 1.0 were used. Objects with M⋆ greater than 1010 M⊙ were selected while the merging partner of the merging systems must be larger than 109 M⊙. The merging partner must also be more than 10% of the M⋆ of the primary galaxy. Galaxies were deemed to have merged when they are tracked as two galaxies in one simulation snapshot and then tracked as one galaxy in the following snapshot in the EAGLE merger trees catalogue (Qu et al. 2017). This prevents the inclusion of chance flybys that may be selected as mergers if the EAGLE galaxies were selected based on proximity. Systems that are projected to merge, using a closing velocity extrapolation, within the next 0.3 Gyr (pre-merger) or are projected to have merged, again using a closing velocity extrapolation of the progenitors, within the last 0.25 Gyr (post-merger) were selected, along with a number of non-merging systems, and gri band images were created of these systems. Springel et al. (2005) have shown that the effects of a merger are visible for approximately 0.25 Gyr after the merger event while the pre-merger stage is much longer. However, we chose to have the pre and post merger period approximately equal as tests conducted with longer pre-merger times showed no improvement, see discussion in Sect. 4.2. We note, however, that the merger timing may suffer from imprecision as a result of the coarse time resolution of the EAGLE simulation, that is the time between snapshots, which becomes coarser at lower simulation redshift. Each galaxy was imaged at an assumed distance of 10 Mpc and each image contains all material within 100 kpc of the centre of the target galaxy and is 256 × 256 pixels, where 256 pixels corresponds to a physical size of 60 kpc. There are 537 pre-merger, 339 post-merger and 335 non-merging systems, each with six random projections to increase the size of the training set. Each of the six projections are treated as individual galaxies resulting in 3222 pre-merger, 2034 post-merger and 2010 non-merging galaxy images for training. The pre-mergers and post-mergers were combined to form the merger class, keeping the pre-merger image if the same galaxy appears in both sets.
|
[
"Springel et al. (2005)"
] |
[
"have shown that the effects of a merger are visible for approximately 0.25 Gyr after the merger event while the pre-merger stage is much longer. However, we chose to have the pre and post merger period approximately equal as tests conducted with longer pre-merger times showed no improvement, see discussion in Sect. 4.2."
] |
[
"Compare/Contrast"
] |
[
[
1068,
1090
]
] |
[
[
1091,
1412
]
] |
2017MNRAS.469.4620B___2017_Instance_1
|
In Table 3, we present a detail comparison of our theoretical energy values from the layer n = 7. For this purpose, we have taken four data sets available in the literature. First, we compare with the ab initio values of Fritzsche (1995) because they have used a larger configuration space than previous ab initio computations (Vajed-Samii & MacDonald 1981; Huang et al. 1983). Secondly, theoretical values from the CHIANTI V.8 (Landi et al. 1999; Del Zanna et al. 2015) because these are considered the most appropriate data for astrophysical applications (Träbert 2014). Thirdly, with the most recent semi-empirically predicted and tentative experimental data values of Del Zanna & Badnell (2016). Fourthly, with the NIST data base (NIST Atomic Spectra Database (ver. 5.3) 2017), which is considered as the most critically analysed data base. Comparison with observed energies (Del Zanna & Badnell 2016) and with the NIST values shows that the present ab initio values are not fully converged to experimental values. In order to achieve better accuracy relative to experimental values, we suggest more balanced and extended reference configuration sets as a starting point of the orbital expansion. This, however, results in very large expansions of wave functions, which is computationally very expensive. Nevertheless, with the present set of calculations for CV correlations, results for the lowest 31 levels are better than obtained before in any ab initio investigation. A very good agreement with experiment is achieved to an accuracy better than 0.5 per cent for low lying levels and 0.8 per cent for higher lying levels, whereas this accuracy is 1–2 per cent for the mcdf calculations of Fritzsche (1995) and for superstructure calculations available at CHIANTI V.8 (Landi et al. 1999; Del Zanna et al. 2015). In some more detail, there is no experimental connection in NIST data base to the 5 levels with J = 7/2 and two levels with J = 9/2 for the 3s23p43d configuration with other levels. The present results of energies for these seven levels with less than 0.45 per cent error with new tentative experimental assignments (Del Zanna & Badnell 2016) testify them. However, we suggest different level ordering than provided by semi-empirical formalism of Del Zanna & Badnell (2016). We briefly discuss later the shift in level ordering and reasons for it. We also note that there is 2100 cm−1 difference between the latest experimental energy (Del Zanna & Badnell 2016) and NIST data base for the level no. 31 (2D3/2). The present value of 683 070 cm−1 confirms the experimental energy (Del Zanna & Badnell 2016), and we suggest a replacement in the NIST data base with new experimental value.
|
[
"NIST Atomic Spectra Database (ver. 5.3) 2017"
] |
[
"Fourthly, with the NIST data base",
"which is considered as the most critically analysed data base.",
"Comparison with observed energies",
"and with the NIST values shows that the present ab initio values are not fully converged to experimental values."
] |
[
"Uses",
"Uses",
"Similarities",
"Similarities"
] |
[
[
735,
779
]
] |
[
[
700,
733
],
[
782,
844
],
[
845,
878
],
[
906,
1018
]
] |
2022MNRAS.516.2500C__Lin_et_al._2009_Instance_1
|
Neutron star X-ray binaries are an important class of low-mass X-ray binaries to understand the radiative and dynamical configuration of the inner region of an accretion disc. Though from previous studies especially based on RXTE (Rossi X-ray Timing Explorer) data of Z sources, it was known that there must exist a corona/comptonization region to explain the observed hardtail in their X-ray spectra but the exact location and how it changes across the intensity variation is not yet properly understood. Among the two primary categories i.e. Z and Atoll sources, Z sources emit close to the Eddington luminosity (0.5–1.0 LEdd; Done, Gierliński & Kubota 2007a) and they exhibit ‘Z’ and ‘C’ shape intensity variation in the hardness intensity diagram (HID) or colour–colour diagrams (CCDs; Hasinger & Van der Klis 1989; Van der Klis 2006). The Z shape variation constitutes a horizontal branch (HB) at the top, a flaring branch (FB) at the bottom, and a normal branch (NB) connecting them diagonally. These are further classified into two broad groups, namely Sco and Cyg-like sources, due to their different appearance exhibited by the HB and FB i.e. less vertical orientation of HB and a weaker FB is seen among Cyg-like sources than in Sco-like (Kuulkers et al. 1994). The hybrid source XTE J1701–462 occupies a special place among NS LMXBs and is considered to be a remarkable source, as it displays all the characteristics exhibited by both Z and atoll sources (Homan et al. 2010, 2007; Lin et al. 2009). At the brightest state, the intensity variations were associated with HB, NB, and FB of Cyg-like and exhibited Sco-like variation at relatively lower brightness. During the decay phase, the variation closely resembles the soft state of an atoll source and later transits to the hard state of the atoll source just before going to the quiescent state. Many important results were noticed based on the spectral fitting of RXTE data of this source. The mass accretion rate was found to be constant along with the Z phase in Sco-like variation and different mechanisms were proposed to explain the spectral and timing variations during the Z phase variations (Lin et al. 2009). It was also found that mass accretion rate is the important driving parameter during the Z and all along with the atoll phases variation. Z sources are unique probes in the sense they provide a platform to understand the structure of accretion disc emitting close to Eddington luminosity because due to the radiation pressure the structure of the inner region of accretion is affected. The previous studies suggested that the interplay between the accretion disc and comptonization region mutually varies to produce the observed tracks in the HID. However, other physical components like a boundary layer (Popham & Sunyaev 2001) or a transition layer (TL) (Osherovich & Titarchuk 1999a, b; Titarchuk & Osherovich 1999) cannot be ruled out. The comptonization region can be in the form of a quasi-spherical cloud or it could be a base of a jet that causes the observed hard continuum in the X-ray spectrum (Migliari et al. 2007). But its association with dynamical features like various branch oscillations or band-limited noises is not known. The spectra of Z sources can also be explained by a structure known as the boundary layer over the NS surface but again, its association to the observed HBO, NBO, etc., is not properly understood (Popham & Sunyaev 2001; Gilfanov, Revnivtsev & Molkov 2003; Revnivtsev & Gilfanov 2006). Based on the detailed spectral modelling of GX 17 + 2, BL occupies a smaller area at the lower vertex (i.e. bottom of NB) in comparison to its area in other branches (Lin et al. 2012) and the comptonization dominates at the HB branch that fades away as source traverse to the FB. The inner disc radius was found to be moving towards the NS, as the Z track evolves from HB to FB. All these structural and radiative variations are found to be occurring at an almost constant mass accretion rate (Lin et al. 2009, 2012).
|
[
"Lin et al. 2009"
] |
[
"The hybrid source XTE J1701–462 occupies a special place among NS LMXBs and is considered to be a remarkable source, as it displays all the characteristics exhibited by both Z and atoll sources"
] |
[
"Motivation"
] |
[
[
1492,
1507
]
] |
[
[
1272,
1465
]
] |
2021MNRAS.504.3316B__than_2000_Instance_3
|
WASP-43b is the most heavily scrutinized phase curve, with four analyses of this data set already published (Stevenson et al. 2017; Mendonça et al. 2018; Morello et al. 2019; May & Stevenson 2020). Our phase curve semi-amplitude, eclipse depth, and radius are consistent with all of these works. The more contentious issue is that of the phase curve’s phase offset and nightside temperature. Stevenson et al. (2017) initially reported only a 2σ upper limit on the nightside temperature of 650 K, while all subsequent reanalyses (including ours) favour a significantly detectable nightside temperature of ∼800 K. As for the planet’s phase offset, Stevenson et al. (2017) and May & Stevenson (2020) favour a larger phase offset (21 ± 2 °E) than Mendonça et al. (2018) and Morello et al. (2019) (12 ± 3 °E and 11 ± 2 °E). May & Stevenson (2020) claimed that the differences between the retrieved phase offsets is the result of temporal binning which was not used by Stevenson et al. (2017) and May & Stevenson (2020) but was used by Mendonça et al. (2018), Morello et al. (2019), and this work. Fitting the temporally binned photometry for all 17 phase curves with each of our detector models already required more than 2000 CPU hours, and expanding this to unbinned photometry for all phase curve fits would require more than 125 000 CPU hours (or 434 d using our 12× multithreading computer) optimistically assuming all of detector models scaled linearly with the number of input data. However, we did try fitting just the WASP-43b unbinned phase curve with our preferred detector model (BLISS) and found that our phase offset and nightside temperature was unchanged. Including a linear slope in time also did not affect our phase offset or nightside temperature. Instead, we find that the phase offset inferred by our models depends on the choice of phase curve model, as our 4-parameter (v2) phase curve models are consistent with those of Stevenson et al. (2017) and May & Stevenson (2020), while our 2-parameter phase curve models (v1) are consistent with Mendonça et al. (2018) and Morello et al. (2019). Ultimately, we cannot decide between these two discrepant offsets as the ΔBIC between the two phase curve models for our preferred BLISS detector model is only 3.7 (insignificantly favouring the 20.4 ± 3.6 offset from the v2 model). For reference, Stevenson et al. (2014b) found phase offsets ranging from roughly −6 to 17 deg east in the Hubble/WFC3 bandpass.
|
[
"Stevenson et al. (2017)"
] |
[
"As for the planet’s phase offset,",
"and May & Stevenson (2020) favour a larger phase offset (21 ± 2 °E) than Mendonça et al. (2018) and Morello et al. (2019) (12 ± 3 °E and 11 ± 2 °E)."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
646,
669
]
] |
[
[
612,
645
],
[
670,
818
]
] |
2015MNRAS.446.1140T__Murray_et_al._2010_Instance_3
|
Recently, forms of feedback that are fundamentally different from SNe have been shown to be essential to galaxy formation. Murray, Quataert & Thompson (2010) analysed the dynamical effects of several forms of stellar feedback on parent molecular clouds. In their models they include momentum input from ionized gas in H ii regions, shocked stellar winds, hot gas pressure, protostellar jets and cosmic rays. Murray et al. (2010) conclude that radiation pressure (RP) on dust grains is likely to be the dominant form of feedback in star-forming galaxies. A variety of other studies have reached the same conclusions, placing the combination of RP and photoionization of gas by massive stars as the dominant mechanism for disruption of molecular clouds and internal regulation of the SF process (Indebetouw et al. 2009; Krumholz & Matzner 2009; Murray et al. 2010; Andrews & Thompson 2011; Hopkins, Quataert & Murray 2011; Lopez et al. 2011; Pellegrini, Baldwin & Ferland 2011). RP alone might also be the only mechanism that explains galactic fountains and the warm gas outflows observed in absorption in high-redshift galaxies (Murray, Ménard & Thompson 2011). In addition, recent numerical work by Krumholz & Thompson (2012) shows that radiation feedback fully accounts for the large gas velocity dispersions measured in young star clusters in the MW. There are at least three reasons why radiative feedback is an essential ingredient of the galaxy formation process. First, observations show that molecular clouds begin to disperse shortly after the O stars form and before the first SNe explode and deposit their energy into the gas (Kawamura et al. 2009). Secondly, the total energy output of a stellar cluster is dominated by radiation. The rate of radiative energy output by O and B stars is ∼200 times larger than the average power injected by SNe and stellar winds during the lifetime of the most massive stars. Thirdly, it is difficult to explain the large gas turbulence values observed in star-forming regions without including the momentum input by radiation (Murray et al. 2010).
|
[
"Murray et al. 2010"
] |
[
"There are at least three reasons why radiative feedback is an essential ingredient of the galaxy formation process.",
"Thirdly, it is difficult to explain the large gas turbulence values observed in star-forming regions without including the momentum input by radiation"
] |
[
"Background",
"Background"
] |
[
[
2072,
2090
]
] |
[
[
1353,
1468
],
[
1920,
2070
]
] |
2018MNRAS.473.4279D__Markevitch_et_al._2002_Instance_1
|
The Hubble Frontier Fields program1 (or HFF hereafter, Lotz et al. 2017) provides the most remarkably detailed examples of gravitational lensing by galaxy clusters, registering hundreds of multiply lensed galaxies for charting galaxy formation to unprecedented depths (see e.g. Jauzac et al. 2014, 2015a,b; Lam et al. 2014; Zitrin et al. 2014; Diego et al. 2015b,c, 2016; Kawamata et al. 2016; Limousin et al. 2016; Mahler et al. 2017). Furthermore, most of these HFF clusters are in a state of collision, enhancing their value for assessing the collisionality of dark matter (DM), a basic assumption of the standard particle interpretation of DM (Markevitch et al. 2002, 2004; Springel & Farrar 2007; Randall et al. 2008). Many clusters exhibit significant, but modest, offsets between the peak of the DM distribution and the centroid of the X-ray emission (Markevitch et al. 2004; Clowe et al. 2006; Mahdavi et al. 2007; Dawson et al. 2012; Menanteau et al. 2012), which is expected if DM is collisionless. These observations can provide a constraint on the DM cross-section (Markevitch et al. 2004; Randall et al. 2008). It is important that these differences are evaluated with the guidance of hydrodynamical models, as complex multibody interactions may also separate the DM from the plasma that can be explained without new physics, as is clearly evident in extreme cases of the bullet cluster (Mastropietro & Burkert 2008), and like the El Gordo cluster (Molnar & Broadhurst 2015), where high-speed collisions between pairs of clusters are ongoing. More direct evidence for collisional DM would be concluded from differences between the stellar and DM distributions as the stars behave like collisionless particles and we should expect the collisionless DM to follow the gravitational potential in the same way. Offsets between the position of the DM central peak and the peak of the luminous matter are difficult to explain with a standard Λcold dark matter but are naturally produced for reasonable values of the DM cross-section (Rocha et al. 2013). A difference of this nature has been claimed recently based on detailed lensing data in the centre of a cluster that contains four bright member galaxies (Massey et al. 2015). In case of the HFF clusters, it is interesting that our free-form analysis of MACS0146 also indicates a possible offset between the lensing-based centroids of the brightest galaxies and their luminous stellar centroids. These differences are subtle and it will be important to look at a larger sample and the model dependencies, and systematic uncertainties, in detail to support any claim of new physics.
|
[
"Markevitch et al. 2002"
] |
[
"Furthermore, most of these HFF clusters are in a state of collision, enhancing their value for assessing the collisionality of dark matter (DM), a basic assumption of the standard particle interpretation of DM"
] |
[
"Motivation"
] |
[
[
648,
670
]
] |
[
[
437,
646
]
] |
2021ApJ...916...61F__Hutsemékers_et_al._2019_Instance_1
|
There are several popular scenarios employed to explain the observational features of CL AGNs. One scenario is that the broad emission lines are obscured by the torus or moving clouds over the observer’s line of sight (Goodrich & Miller 1989), while only several CL AGNs can be explained by this scenario (Wang et al. 2019; Jaffarian & Gaskell 2020; Kokubo & Minezaki 2020). The features observed in most CL AGNs, e.g., the complex multiband spectral variabilities, and strong changes seen in the infrared or low level of polarization, strongly argue against the scattering (or obscuration) scenario (e.g., Sheng et al. 2017; Mathur et al. 2018; Stern et al. 2018; Hutsemékers et al. 2019; Kynoch et al. 2019, and the references therein). Another attractive scenario is that the CL AGNs are indeed tidal disruption events (Merloni et al. 2015; Kawamuro et al. 2016; Yang et al. 2019; Padmanabhan & Loeb 2020; Ricci et al. 2020; Zhang 2021), while it is not a general mechanism to explain all CL AGNs, such as repeating CL AGNs. The changing look phenomenon triggered by the change of the mass accretion rate of the accretion disk is a rather straightforward model (Husemann et al. 2016; Liu et al. 2019; Ruan et al. 2019; Ai et al. 2020; Sniegowska et al. 2020); however, it has a fatal problem that propagation timescale of the gas in a thin accretion disk is much longer than the observed timescale in CL AGNs, unless the viscous thin disk model is somewhat revised (Lawrence 2018). Dexter & Begelman (2019) proposed that the magnetically elevated disk model could help to explain the changing look timescale. Based on this scenario, Scepi et al. (2021) suggested that a magnetic flux inversion in a magnetically arrested disk is able to explain the CL event in 1ES 1927+654. Sniegowska et al. (2020) suggested that a narrow unstable zone between the outer thin disk and the inner ADAF could cause the periodic outbursts in repeating CL AGNs. Recently, the effects of large-scale magnetic fields on this scenario for repeating CL AGNs have been studied in detail by Pan et al. (2021), and Lyu et al. (2021) suggested that the change of broad emission lines in Mrk 1018 might be regulated by the evolution of accretion disk.
|
[
"Hutsemékers et al. 2019"
] |
[
"The features observed in most CL AGNs, e.g., the complex multiband spectral variabilities, and strong changes seen in the infrared or low level of polarization, strongly argue against the scattering (or obscuration) scenario (e.g.,"
] |
[
"Compare/Contrast"
] |
[
[
665,
688
]
] |
[
[
375,
606
]
] |
2018ApJ...866L...1S__Pecharromán_et_al._1999_Instance_5
|
It was found that the complex dielectric function from Pecharromán et al. (1999) for the sample obtained by heating bayerite at 1273 K, assuming a spheroid with depolarization parameters of (0.35, 0.003), produced an opacity with 11, 20, 28, and 32 μm features, so this component was included in the models. However, with only this component, the observed 20 μm features in the residual spectra were found to be wider than those in the models. By adding the opacity of the sample obtained by heating boehmite at 1173 K, the width of the 20 μm feature could be matched. This was done using the complex dielectric function for the sample obtained by heating boehmite at 1173 K from Pecharromán et al. (1999), assuming a spheroid with depolarization parameters of (0.35, 0.035). The complex dielectric functions of the samples obtained by heating bayerite and boehmite to various temperatures (Pecharromán et al. 1999) were derived by modeling the reflectance spectra of pellets obtained by pressing powders of these materials under great pressure. This method required Pecharromán et al. (1999) to assume an effective medium theory, such that a pellet is a mixture of one of their samples with a matrix of air. Pecharromán et al. (1999) noted that heating bayerite at 500°C eliminates the XRD pattern of bayerite, and they note that at 700°C, the infrared reflectance spectrum of the boehmite sample no longer shows OH− stretching bands. This must mean that the samples obtained from heating bayerite at 1273 K and from heating boehmite at 1173 K are no longer bayerite or boehmite, respectively. XRD performed by Pecharromán et al. (1999) of the sample of bayerite prepared at 1273 K suggests only θ-alumina was present, and their infrared and NMR spectroscopy confirms this. XRD of their sample obtained from heating boehmite to 1173 K (Pecharromán et al. 1999) suggests δ-alumina to be present, though some amounts of θ-alumina and α-alumina are present, as they deduce from XRD and infrared and NMR spectroscopy.
|
[
"Pecharromán et al. (1999)"
] |
[
"noted that heating bayerite at 500°C eliminates the XRD pattern of bayerite, and they note that at 700°C, the infrared reflectance spectrum of the boehmite sample no longer shows OH− stretching bands."
] |
[
"Compare/Contrast"
] |
[
[
1209,
1234
]
] |
[
[
1235,
1435
]
] |
2017AandA...605A..88L__Bernstein_et_al._2002_Instance_1
|
Altogether, the approximately thirty molecules recently detected have confirmed the chemical complexity in the nebula, and generated our interest for the present study. Of these species, we will focus our attention on the seventeen species listed by molecular families in Table 1. As can be seen in this table, the WHISPER survey allowed the detection of some organic molecules in the Horsehead nebula, such as formaldehyde (H2CO) and methanol (CH3OH), which constitute key species in the likely synthesis of more complex organic molecules such as some prebiotic molecules (Bernstein et al. 2002; Muñoz Caro et al. 2002; Garrod et al. 2008). Because they are detected in a wide variety of interstellar sources – in hot cores (Sutton et al. 1995; Ceccarelli et al. 2000), dark clouds (Bergman et al. 2011), shocked regions (e.g. Sakai et al. 2012; Codella et al. 2012; Tafalla et al. 2010) and even in comets (Mumma & Charnley 2011; Cordiner et al. 2015) – it is of prime importance to understand well how these precursor molecules form. H2CO is commonly thought to form both in the gas-phase and on grain surfaces, while CH3OH is believed to be only formed on grain surfaces (Garrod et al. 2006; Geppert et al. 2006). Guzman et al. (2013) reported the observations of these two molecules toward the Horsehead nebula in both the PDR and Core positions. Unable to reproduce the observed abundances of either H2CO or CH3OH at the PDR position with only pure gas-phase models, they concluded that, for this region, both species are formed on grain surfaces and then photodesorbed into the gas phase. On the other hand, at the Core position, a pure gas-phase model can reproduce the observed H2CO abundance, while photodesorption of ices is still needed to explain the observed abundance of CH3OH. Other organic molecules were reported in the Horsehead nebula as first detections in a PDR environment, including HCOOH (formic acid), CH2CO (ketene), CH3CHO (acetaldehyde), and CH3CCH (propyne) (Guzman et al. 2014). Their abundances were found to be higher at the PDR position than at the Core, revealing that complex organic chemistry is also occurring in UV-illuminated neutral gas (Guzman et al. 2014). Of these molecules, some – HCOOH, CH2CO, and CH3CHO – have now also been detected in the Orion bar PDR (Cuadrado et al. 2016, 2017).
|
[
"Bernstein et al. 2002"
] |
[
"As can be seen in this table, the WHISPER survey allowed the detection of some organic molecules in the Horsehead nebula, such as formaldehyde (H2CO) and methanol (CH3OH), which constitute key species in the likely synthesis of more complex organic molecules such as some prebiotic molecules"
] |
[
"Uses"
] |
[
[
574,
595
]
] |
[
[
281,
572
]
] |
2015AandA...579A.102B__Boselli_et_al._2009_Instance_2
|
Once corrected for dust attenuation, Hα luminosities can be transformed into star formation rates (SFR, in M⊙ yr-1) using a factor that depends on the assumed IMF and stellar model7: (10)\begin{equation} {SFR = k({\rm H}\alpha) \times L({\rm H}\alpha)_{\rm cor}} . \end{equation}SFR=k(Hα)×L(Hα)cor.We recall that this relation is valid only under the assumption that the mean star formation activity of the emitting galaxies is constant on a timescale of a few Myr, roughly comparable to the typical time spent by the stellar population responsible for the ionisation of the gas on the main sequence (Boselli et al. 2009; Boissier 2013; Boquien et al. 2014). The ionising stars are O and early-B stars, whose typical age is ≲107 yr. The stationarity condition is generally satisfied in massive, normal, star-forming galaxies undergoing secular evolution. In these objects, the total number of OB associations is significantly larger than the number of HII regions under formation and of OB stars reaching the final stage of their evolution, thus their total Hα luminosity is fairly constant with time. This might not be the case in strongly perturbed systems or in dwarf galaxies, where the total star formation activity can be dominated by individual giant HII regions (Boselli et al. 2009; Weisz et al. 2012), and the IMF is only stochastically sampled (Lee et al. 2009; Fumagalli et al. 2011; da Silva et al. 2014). The HRS sample is dominated by relatively massive galaxies undergoing secular evolution. For these objects, Eq. (10) can thus be applied. The sample, however, also includes galaxies in the Virgo cluster region, where the perturbation induced by the cluster environment might have affected their star formation rate (e.g. Boselli & Gavazzi 2006, 2014). Models and simulations have shown that in these objects the suppression of star formation occurs on timescales of a few hundred Myr (Boselli et al. 2006, 2008a,b, 2014d). These timescales are relatively long compared to the typical age of O-B stars. The recent work of Boquien et al. (2014) has clearly shown that the Lyman continuum emission tightly follows the rapid variations in the star formation activity of simulated galaxies down to timescales of a few Myrs. We can thus safely consider that the linear relation between the Hα luminosity and the star formation rate given in Eq. (10) is satisfied in the HRS sample.
|
[
"Boselli et al. 2009"
] |
[
"In these objects, the total number of OB associations is significantly larger than the number of HII regions under formation and of OB stars reaching the final stage of their evolution, thus their total Hα luminosity is fairly constant with time. This might not be the case in strongly perturbed systems or in dwarf galaxies, where the total star formation activity can be dominated by individual giant HII regions"
] |
[
"Compare/Contrast"
] |
[
[
1271,
1290
]
] |
[
[
855,
1269
]
] |
2020ApJ...888...46C__Singh_et_al._1995_Instance_1
|
The theoretical models mentioned above have made assumptions. The validity of these assumptions needs to be examined. Besides, all these models involve parameters to be determined. These problems could be better understood through numerical simulations. Hurlburt et al. (1994) performed a two-dimensional numerical simulation of compressible flow on downward overshooting. They investigated the dependence of subadiabatic extent δ on stability parameter S. They revealed a scaling law of δ ∝ S−1 in the penetrative layer and δ ∝ S−1/4 in the overshooting layer, respectively. This result is in good agreement with Zahn’s analytic model (Zahn 1991). Freytag et al. (1996) performed two-dimensional numerical simulations with a realistic description of radiation and ionization on A-type stars and DA white dwarf stars. They described the overshooting as a diffusive process, and derived an exponential decay parameter for the diffusion. Early attempts of low-resolution three-dimensional numerical simulations on overshooting were made by Singh et al. (1994; upward overshooting), Singh et al. (1995, 1998), and Saikia et al. (2000) (downward overshooting). The scaling laws derived from the numerical simulations of downward overshooting (Singh et al. 1995, 1998; Saikia et al. 2000) agree well with Zahn’s analytical model. Only a qualitative result was given in the numerical simulations of upward overshooting (Singh et al. 1994). High-resolution numerical simulations of downward overshooting across a wide range of parameters were presented by Brummell et al. (2002). They confirmed the −1/4 scaling law of the thermal adjustment overshooting layer, while the −1 scaling law of the nearly adiabatic penetrative layer was absent in the simulations. Based on a semianalytic model, Rempel (2004) argued that the absence of the nearly adiabatic penetrative layer is caused by the large energy flux specified in the numerical simulations. Numerical experiments on Boussinesq flow were performed by Korre et al. (2019). They reported steeper scaling laws of δ ∝ S−1/3 or δ ∝ S−1/2, depending on the steepness of the background radiative temperature gradient. Simulations with realistic physical variables on stellar core convection were performed by Browning et al. (2004) and Brun et al. (2005). The effects of rotation and magnetic field are considered. They found that the penetrative convection yields a prolate shape of a nearly adiabatic region. Kitiashvili et al. (2016) performed 3D radiative hydrodynamic simulations of the outer layer of a moderate-mass star (1.47 solar mass). Their result discovered a nearly adiabatic layer and a deeper subadiabatic layer. The recent work of Brun et al. (2017) simulated the differential rotation and overshooting in solar-like stars. Their result indicated that slow rotators favor a wider overshooting region near the poles and fast rotators at mid-to-low latitude. Hotta (2017) performed numerical simulations on the solar overshooting with very low energy fluxes F. He found that the overshooting distance obeys a scaling law of δ ∝ F0.31. Käpylä (2019) conducted numerical experiments on downward overshooting by considering the effect of the smoothness of the heat conduction profiles. He discovered that the power-law index of the overshooting distance on the energy flux is smaller in the smooth heat conduction profile than in the step profile. Efforts on prediction of 321D turbulent theory were made by Arnett et al. (2015) and Arnett & Moravveji (2017). They separated the overshooting region into three layers: a fully mixed layer, a partially mixed wave layer, and an extra diffusive mixing layer. With the scale analysis of turbulent plumes and eddies, Viallet et al. (2015) discussed the three possible regimes of turbulent overshooting: a diffusion-dominated regime (only mix composition), a penetrative regime (transition within the boundary layer), and an entrainment regime (mix both entropy and composition). The selection criterion of different regimes during a stellar evolution calculation is not well defined yet.
|
[
"Singh et al. (1995"
] |
[
"Early attempts of low-resolution three-dimensional numerical simulations on overshooting were made by",
"(downward overshooting)"
] |
[
"Background",
"Background"
] |
[
[
1080,
1098
]
] |
[
[
936,
1037
],
[
1132,
1155
]
] |
2019AandA...622A.106M__Lanz_et_al._(2010)_Instance_2
|
The standard single-frequency detection methods for point sources in the CMB and far IR are based on wavelet techniques (Vielva et al. 2003; Barnard et al. 2004; González-Nuevo et al. 2006) or on the matched filter (or MF hereafter, Tegmark & de Oliveira-Costa 1998; Herranz et al. 2002; Barreiro et al. 2003; López-Caniego et al. 2006, see also Herranz & Vielva 2010 for a review.). Wavelets are well suited for the detection of compact sources due to their good position-scale determination properties, whereas the MF is the optimal linear detector-estimator because it provides the maximum signal-to-noise (S/N) amplification for a source with a known shape (usually the point-spread function, or PSF hereafter, of the telescope) embedded in statistically homogeneous and spatially correlated noise. By default, these techniques are applicable only to single-frequency sky images: even for multiwavelength observatories such as the Herschel Space Observatory (Pilbratt et al. 2010) or Planck (Tauber et al. 2010), the standard detection pipelines have produced individual source catalogs for each frequency band (see e.g., Planck Collaboration VII 2011; Planck Collaboration XXVIII 2014; Planck Collaboration XXVI 2016; Maddox et al. 2018). The next logical step is to boost the signal of faint sources by combining the different bands into a single detection, that is, “multifrequency detection”. Most of the blind component separation algorithms that are used for diffuse components in microwave and far IR astronomy can not deal with the high diversity of spectral behaviors associated to the different populations of extragalactic compact sources (see for example Leach et al. 2008). However, over the last few years a number of multifrequency compact source detection techniques have been proposed in the literature (Herranz & Sanz 2008; Herranz et al. 2009; Lanz et al. 2010, 2013; Planck Collaboration Int. LIV 2018). A review on the topic can be found in Herranz et al. (2012). In particular, if the spatial profile and the spectral energy distribution (SED) of the sources are known, and if the cross-power spectrum is known, or can be estimated from the data, the optimal linear detection method is the matched multifilter (or MMF hereafter, Herranz et al. 2002). Lanz et al. (2010) also showed that the MMF can be generalized for the case where the SED of the sources is not known. This generalization outperforms the single-frequency MF in terms of S/N and can be used to infer the spectral index of synchrotron-dominated radio sources, as shown in Lanz et al. (2013). However, in this paper we will incorporate a specific SED to the MMF in order to derive a photometric redshift estimation of dusty galaxies and high-redshift star forming galaxies detected in the IR part of the spectrum1. We will do so by applying the multifrequency MMF filter to the first and second data releases of the Herschel Astrophysical Terahertz Large Area Survey (the Herschel-ATLAS or H-ATLAS, Eales et al. 2010), the largest single key project carried out in open time with the Herschel Space Observatory. We restrict our multifrequency analysis to the three wavelength bands covered by the SPIRE instrument aboard Herschel (Griffin et al. 2010), centered around 250, 350 and 500 μm. As discussed in Hopwood et al. (2010), Lapi et al. (2011), González-Nuevo et al. (2012), Pearson et al. (2013) and Donevski et al. (2018), the SPIRE bands are ideal for capturing the peak in the SED corresponding to dust emission of star-forming galaxies at z ∼ 2, that is redshifted from its rest-frame wavelength around 70–100 μm to the SPIRE wavelengths: This is the redshift range where galaxies have formed most of their stars. At higher redshifts, dusty star-forming galaxies (DSFGs) occupy the most massive halos and are among the most luminous objects found at z ≳ 4 (Michałowski et al. 2014; Oteo et al. 2016; Ikarashi et al. 2017). These high-redshift DSFGs have markedly red colors as seen by SPIRE, with rising flux densities from 250 to 500 μm (the so-called “500 μm-risers”), and have received a great deal of attention in the recent years (see for example Ivison et al. 2016; Negrello et al. 2017; Strandet et al. 2017). The DSFGs, and particularly the 500 μm risers uncovered by Herschel, are providing much insight into the early star forming history of the universe. However, sensitivity and limited angular resolution severely constrain the power of this type of objects as astrophysical probes. The sensitivity of SPIRE allows for the direct detection of only the brightest, and thus rarest objects, at the bright end of the luminosity function. By means of our multifrequency MMF technique, we intend to enhance the detectability and statistical significance of very faint red objects in the H-ATLAS source catalog and so expand the list of reliable 500 μm-riser candidates.
|
[
"Lanz et al. (2010)"
] |
[
"also showed that the MMF can be generalized for the case where the SED of the sources is not known.",
"This generalization outperforms the single-frequency MF in terms of S/N and can be used to infer the spectral index of synchrotron-dominated radio sources, as shown in Lanz et al. (2013)."
] |
[
"Background",
"Compare/Contrast"
] |
[
[
2277,
2295
]
] |
[
[
2296,
2395
],
[
2396,
2583
]
] |
2015ApJ...808...56M__Deming_et_al._2015_Instance_1
|
We have tested the pixel-ICA algorithm, i.e., a non-parametric method proposed by Morello et al. (2014, 2015) to detrend Spitzer/IRAC primary transit observations, on simulated data sets. Systematics similar to the ones present in Spitzer/IRAC data sets are obtained by combining instrumental jitter with inter- or intra-pixel sensitivity variations. A variety of jitter timeseries is used to test the pixel-ICA method with:
1.
periodic signals with different frequencies, phasing, and shape;
2.
non-stationary signals with varying amplitudes or frequencies;
3.
sudden change.
The detrending performances of pixel-ICA method have been compared with division by a polynomial function of the centroid, in this paper PCD method, and PLD method (Deming et al. 2015). Here we summarize the main results found:
1.
Pixel-ICA algorithm can detrend non-stationary signals and sudden changes, as well as periodic signals with different frequencies and phasing, relative to the transit.
2.
Inter-pixel effects are well-detrended with pixel-ICA method.
3.
Even if the instrument PSF is not entirely within the array of pixels, pixel-ICA leads to results which are consistent at ∼1σ with the input parameters.
4.
In most cases, pixel-ICA outperforms PCD method, especially if the instrument PSF is narrow, or it is not entirely within the photometric aperture.
5.
Intra-pixel effects are only detectable if the PSF is relatively small.
6.
Intra-pixel effects cannot be totally detrended by any of the three methods, but pixel-ICA, in most cases, outperforms PCD method, which is largely case-dependent. Also, pixel-ICA method provides consistent results within the error bars.
7.
It is possible to fit the astrophysical signal after detrending or together with the other components. The only differences are registered if at least one of the non-transit components has a similar shape at the time of transit, in which case the first approach is preferable, but the two results were consistent within 1σ.
8.
The PLD method, updated to include cross-term between pixel fluctuations and the astrophysical signals, lead to very similar results than pixel-ICA, particularly if the astrophysical signal is fitted together with the other components.
In conclusion, we have found in a variety of simulated cases that pixel-ICA performs as well or better than other methods used in the literature, in particular polynomial centroid corrections and PLD (Deming et al. 2015). The main advantage of pixel-ICA over other methods relies on its purely statistical foundation without the need of imposing prior knowledge on the instrument systematics, therefore avoiding a potential source of error. The results of this paper, together with previous analyses of real Spitzer/IRAC data sets (Morello et al. 2014, 2015), suggest that photometric precision and repeatability at the level of one part in 104 can be achieved with current infrared space instruments.
|
[
"Deming et al. 2015"
] |
[
"The detrending performances of pixel-ICA method have been compared with division by a polynomial function of the centroid, in this paper PCD method, and PLD method"
] |
[
"Compare/Contrast"
] |
[
[
748,
766
]
] |
[
[
583,
746
]
] |
2022MNRAS.516.5289M__Thompson_et_al._2015_Instance_2
|
Given the number densities within the mass-dissociation index plane of Fig. 8, we now ask ourselves whether known dissociated clusters, such as the Bullet cluster, are expected in L210N1024NR? The Bullet Cluster has a mass of $\sim 1.5 \times 10^{15} \, {\rm M}_{\odot }$ (e.g. Clowe et al. 2004; Bradač et al. 2006; Clowe et al. 2006) and we estimated a dissociation index of SBullet ∼ 0.335 ± 0.06. As seen in Fig. 8 there are no Bullet cluster analogues (structures of approximate mass and dissociation) in L210N1024NR, this is unsurprising as a simulation requires a significantly larger volume than that of L210N1024NR ((210cMpc h−1)3) to expect such an object (e.g. Lee & Komatsu 2010; Thompson & Nagamine 2012; Bouillot et al. 2015; Kraljic & Sarkar 2015; Thompson et al. 2015). From the distribution presented in Fig. 8, it is trivial to estimate the required cosmological volume (the effective volume, Veff) to expect structures of a given mass and dissociation index. By separating the 2D distribution on the mass-dissociation index planes into the component 1D distributions of mass and dissociation the effective volume is computed as
(12)$$\begin{eqnarray}
V_\text{eff}~^{-1} &=&\int \int \,{\rm{ d}} S \, {\rm{ d}} M \phi (S, M) \\
&=& \int _{S_\text{a}}^{S_\text{b}} \, {\rm{ d}} S \phi _S(S) \int _{M_\text{a}}^{M_\text{b}} \, {\rm{ d}} M \phi _M(M)~,
\end{eqnarray}$$where ϕS(S) is the number density function associated with S and $\phi _\mathit {M}(\mathit {M})$ is the mass function presented in Fig. 7. Assuming a probable range of S = 0.335 ± 0.06 and $1 \lt M \lt 2 \times 10^{15} \, {\rm M}_{\odot }$ we estimate a number density ∼4.92 × 10−10 Mpc−3 or that an effective volume of ∼2.03 Gpc3 would be required to observe a single Bullet-like cluster. This result is inline with the number density estimate of the order of ∼10−10 Mpc−3 by Thompson et al. (2015), which improves on previous estimates (e.g. Lee & Komatsu 2010; Thompson & Nagamine 2012; Bouillot et al. 2015) due to more sophisticated halo finding methods (e.g. Behroozi, Wechsler & Wu 2013). Conversely, it was estimated by Kraljic & Sarkar (2015) (utilizing the same halo finder as Thompson et al. 2015) that given an effective volume of ∼14.6 Gpc3, no Bullet cluster analogues are expected, however as indicated by a pairwise velocity distribution it would be expected that present binary halo–halo orbits have the potential to form a Bullet-like object.
|
[
"Thompson et al. (2015)"
] |
[
"This result is inline with the number density estimate of the order of ∼10−10 Mpc−3 by"
] |
[
"Similarities"
] |
[
[
1881,
1903
]
] |
[
[
1794,
1880
]
] |
2022ApJ...935..135B__Mathur_1990_Instance_1
|
All responses calculated in this paper only account for the direct response to a perturbing potential. In general, though, the response also has an indirect component that arises from the fact that neighboring regions in the disk interact with each other gravitationally. This self-gravity of the response, which we have ignored, triggers long-lived normal-mode oscillations of the slab that are not accounted for in our treatment. Several simulation-based studies have argued that including self-gravity is important for a realistic treatment of phase spirals (e.g., Darling & Widrow 2019a; Bennett & Bovy 2021). Using the Kalnajs matrix method (Kalnajs 1977; Binney & Tremaine 2008), we have made some initial attempts to include the self-gravity of the response in our perturbative analysis, along the lines of Weinberg (1991). Our preliminary analysis shows that the self-gravitating response is a linear superposition of two terms: (i) a continuum of modes given in Equation (12), dressed by self-gravity, that undergo phase mixing and give rise to the phase spiral; and (ii) a discrete set of modes called point modes or normal modes (see Mathur 1990; Weinberg 1991) that follow a dispersion relation. The continuum response can be amplified by self-gravity when the continuum frequencies, nΩ
z
+ kv
x
, are close to the point-mode frequencies, ν. Depending on the value of k, the normal modes can be either stable or unstable. Araki (1985) finds that in an isothermal slab the bending normal mode undergoes fire hose instability below a certain critical wavelength if σ
z
/σ ≲ 0.3, while the breathing normal mode becomes unstable above the Jeans scale. In the regime of stability, the normal modes are undamped oscillation modes in absence of lateral streaming (Mathur 1990) but are Landau damped otherwise (Weinberg 1991). For an isothermal slab with typical MW-like parameter values, the point modes are strongly damped since their damping timescale (inverse of the imaginary part of ν) is of order their oscillation period (inverse of the real part of ν), which turns out to be of order the vertical dynamical time, h
z
/σ
z
. Moreover, the normal-mode oscillations are coherent oscillations of the entire system, independent of the vertical actions of the stars, and are decoupled from the phase spiral in linear theory since the full response is a linear superposition of the two. Based on the above arguments, we conclude that self-gravity has little impact on the evolution of phase spirals in the isothermal slab, at least in the linear regime. We emphasize that Darling & Widrow (2019a), who found their phase spirals to be significantly affected by the inclusion of self-gravity, assumed a perturber-induced velocity impulse with magnitude comparable to the local velocity dispersion in the solar neighborhood; hence, their results are likely to have been impacted by nonlinear effects. Moreover, the self-gravitating response of an inhomogeneous disk embedded in a DM halo, as in the simulations of Darling & Widrow (2019a), can be substantially different from that of the isothermal slab. We intend to include a formal treatment of self-gravity along the lines of Weinberg (1991) in future work.
|
[
"Mathur 1990"
] |
[
"Our preliminary analysis shows that the self-gravitating response is a linear superposition of two terms: (i) a continuum of modes given in Equation (12), dressed by self-gravity, that undergo phase mixing and give rise to the phase spiral; and (ii) a discrete set of modes called point modes or normal modes (see",
"that follow a dispersion relation."
] |
[
"Uses",
"Uses"
] |
[
[
1145,
1156
]
] |
[
[
831,
1144
],
[
1173,
1207
]
] |
2018AandA...620A..80M__observations,_Gerin_et_al._2017_Instance_1
|
Methyl formate is the species for which the largest number of transitions were identified. In total 24 transitions from both A and E species were observed (18 unblended), with energy levels between 100 and 500 K, allowing to constrain the properties of the emitting region from the LTE model. We ran a set of models using one source component of 0.6″ in size (the angularresolution of the ALMA band 6 observations), and we modified the temperature and column density to fit the observed spectra. However, it was difficult to fit all the lines with only one component. The problem became more evident when using the deuterated substitutions, CH3OCOD and CH2DOCOH, and the vibrationally excited states νt = 1 and νt = 2. We have detected 12 transitions from CH3OCOD and four transitions from CH2DOCOH (three and two, respectively, are blended with other species), with upper energy levels between 90 and 170 K. In both cases the model gave better results when decreasing the temperature with respect to the best CH3OCOH model. On the other hand, the vibrationally excited lines (13 unblended transitions from νt = 1, and one transition from νt = 2), supported the high temperature model. Therefore, we built a model of the source with two components. One warm and compact component at 200 K and a size of 0.35″ (that of the continuum compact emission obtained by ALMA Band 7 observations, Gerin et al. 2017), and another larger (0.6″, approximately the angular resolution of the Band 6 observations) and colder component at 60 K. Figure B.4 shows two examples of the best one component models obtained using the main CH3OCOH (200 K), and the deuterated methyl formate lines (60 K), together with the final two component model. While some species give reasonable agreement with observations using either model (mostly those which observed transitions have similar energy levels), there are others (besides the ones discussed above) that fail in one of the models (e.g. HNCO and t-CH3CH2OCOH). We have checked that using other temperatures (between 50 and 250 K) in the one component models does not reproduce satisfactorily the observations. For the vibrationally excited states, we only used the compact component at 200 K, since the low temperature one does not contribute much to the line intensities. Indeed, energy levels of the observed transitions range between 300 and 500 K, supporting this scenario. We have also observed five transitions of 13CH3OCOH, with energies ≤200 K and intensities of the order of 500 mK or below. Since the lines are weak, any of the models give reasonable agreement with the observations. Acetaldehyde is the other species for which many lines are observed. We have detected 15 transitions (10 unblended) from the main species, with energies between 100 and 300 K. We have also detected one transition arising from the deuterated substitution CH3CDO, and four transitions (two of them blended) from the vibrational state νt = 1, with energy levels of ~300–400 K. We used the same model as methyl formate, except for CH3CHO νt = 1, for which only the compact and hot component was considered. The model allows to reproduce all transitions reasonably well, except for a few lines that are overestimated: two CH3OCOH transitions (Fig. B.1, second panel), three CH3OCOH νt = 1 lines (top panel in Fig. B.1, and second panel in Fig. B.3), and two lines from CH3CHO (see Fig. B.3, second and third panels). These are transitions with the lowest energies (~100 K) and the highest line strengths and Einstein coefficients, and may consequently have high opacities. We estimate the error in the fit to be ~50%.
|
[
"Gerin et al. 2017"
] |
[
"One warm and compact component at 200 K and a size of 0.35″ (that of the continuum compact emission obtained by ALMA Band 7 observations,"
] |
[
"Uses"
] |
[
[
1387,
1404
]
] |
[
[
1249,
1386
]
] |
2021MNRAS.504..146V__Vink_&_Gräfener_2012_Instance_1
|
The direct detection of the first gravitational waves from the merger of two heavy black holes (BHs) in GW 150914 confirmed one of the toughest predictions of Einstein’s theory of general relativity. But while satisfying the world of physics in general, for astrophysics this was only the beginning: many were surprised by the large BH masses of, respectively, 36 and 29 M⊙ (Abbott et al. 2016), showcasing how the new field of multimessenger astrophysics had just re-opened the field of stellar evolution in a spectacular fashion. Stellar mass BHs had previously been revealed by their interaction in binary systems (Orosz et al. 2011), but the maximum stellar BH mass in our Milky Way is not higher than roughly 15–20 M⊙ (Belczynski et al. 2010). While we know that very massive stars (VMS) above 100 M⊙ exist (Crowther et al. 2010; Vink et al. 2015), this mass is significantly diminished via stellar winds already during core hydrogen (H) burning (Vink & Gräfener 2012). The heavy nature of the BH, as measured by LIGO/VIRGO therefore supported the assumption that the gravitational wave event occurred in a part of the Universe still pristine in its enrichment with heavy elements (‘metallicity (Z)’), lowering stellar wind mass-loss (Vink, de Koter & Lamers 2001; Vink & de Koter 2005). A low-Z solution was widely accepted until the announcement of a 70 M⊙ BH in LB-1 (Liu et al. 2019), spurring stellar evolution theorists to avoid heavy mass-loss in the Milky Way (Belczynski et al. 2020; Groh et al. 2020), either by arbitrarily lowering the mass-loss rates of VMS – seemingly contradicting VMS mass-loss calibrations (Vink & Gräfener 2012) – or by invoking the presence of a strong dipolar surface magnetic field that could quench the wind (Petit et al. 2017). While such magnetic fields in some 5–10 per cent of massive OB stars do indeed exist, no B-fields have yet been detected in VMS (Bagnulo et al. 2020). The problem of the formation of a $70\, \mathrm{ M}_\odot$ BH in a solar metallicity environment apparently resolved itself when the spectral signatures of LB-1 were re-interpreted (Abdul-Masih et al. 2020; El-Badry & Quataert 2020).
|
[
"Vink & Gräfener 2012"
] |
[
"this mass is significantly diminished via stellar winds already during core hydrogen (H) burning"
] |
[
"Background"
] |
[
[
955,
975
]
] |
[
[
857,
953
]
] |
2021AandA...648A..14R__Gürkan_et_al._(2018)_Instance_1
|
To obtain a more complete picture of the physical processes that shape star formation in the early Universe, very deep radio surveys overwide areas of sky are required to complement deep submillimetre surveys. Previous work has used high-resolution radio observations, typically at 1.4 GHz, as a method of pinpointing the position of submillimetre sources detected in single-dish surveys (Ivison et al. 2002; Chapman et al. 2005); however, such work has been limited by the depth of available radio sky surveys,with dedicated deep surveys over only small regions of sky, resulting in a view biased towards the brighter radio sources. Studies have largely included limited radio spectral coverage of SMGs, focusing on the nature of the FIRC and its relation to properties such as stellar mass and redshift (e.g. Yun et al. 2001; Ivison et al. 2010; Smith et al. 2014). The Low Frequency Array (LOFAR; van Haarlem et al. 2013) has opened up new ways of studying galaxies in the radio, and a number of studies have used LOFAR’s capabilities to investigate this relationship between star formation and radio luminosity in the low-frequency regime – for example Gürkan et al. (2018), Read et al. (2018), Smith et al. (2021), and Wang et al. (2019a). However, these studies generally investigate the statistical properties of large samples of galaxies, in optically selected samples at low redshift (z ≲2), rather than probing the shapes of individual radio spectra. Thomson et al. (2019) conducted an in-depth study of high-frequency (>610 GHz) spectral curvature in SMGs, finding evidence of curved spectra that they attributed to spectral ageing of the synchrotron emission from star formation; their results implied estimated starburst ages consistent with expected SMG lifetimes. Studies at low frequencies, where we may be able to observe absorption processes affecting the shape of the spectrum, have been hampered by a lack of sufficiently deep, wide-area data. More comprehensive observations of the shape of the radio spectrum, extending to lower frequencies, can provide a probe of the physical conditions that give rise to extreme star formation in SMGs. Calistro Rivera et al. (2017) exploit LOFAR’s frequency range to investigate the spectral shapes of star-forming galaxies and AGN, finding evidence of low-frequency spectral flattening in the star-forming sample. This sample is also constrained in redshift, focusing on local galaxies rather than the peak of star formation at z > 2, and so does not probe the bulk of the highly star-forming SMG population. Chyży et al. (2018) also find weak spectral flattening in local star-forming galaxies with LOFAR but largely attribute this slight effect to synchrotron losses, predicting stronger low-frequency spectral flattening due to free–free absorption at high redshift, where galaxies with high SFRs are more common.
|
[
"Gürkan et al. (2018)"
] |
[
"The Low Frequency Array",
"has opened up new ways of studying galaxies in the radio, and a number of studies have used LOFAR’s capabilities to investigate this relationship between star formation and radio luminosity in the low-frequency regime – for example",
"Read et al. (2018), Smith et al. (2021), and Wang et al. (2019a)."
] |
[
"Background",
"Background",
"Background"
] |
[
[
1157,
1177
]
] |
[
[
868,
891
],
[
925,
1156
],
[
1179,
1244
]
] |
2022MNRAS.516.4833J__Boffin_&_Jorissen_1988_Instance_1
|
It is now clear that binary stars hold the key to understanding the formation of many planetary nebulae (PNe; Jones & Boffin 2017a). However, with much of the recent focus being placed on common envelope evolution (Boffin & Jones 2019), the role of wider binaries remains almost completely unconstrained (Tyndall et al. 2013). To date, only three wide binary central stars of PNe have confirmed periods (Jones et al. 2017) with other systems either being so wide that they are resolved (Ciardullo et al. 1999) or displaying composite spectra with giant companions (Tyndall et al. 2013). Four central star systems are known to be barium stars (Bond, Pollacco & Webbink 2003; Miszalski et al. 2012, 2013; Löbling, Boffin & Jones 2019) with only one having a known orbital period (LoTr 5; Jones et al. 2017; Aller et al. 2018). Barium stars are mostly – but not only: see e.g. Escorza et al. (2019) – giant stars of spectral type G-K that display enhanced abundances of carbon and s-process elements such as barium and strontium, and are now known to all be long-period binaries (100 d ≲ Porb ≲ 104 d; McClure, Fletcher & Nemec 1980; Jorissen et al. 2019). The chemical contamination of barium stars is believed to be due to accretion of chemically enriched material from an evolved binary companion, likely through wind (Boffin & Jorissen 1988) or wind Roche lobe overflow (WRLOF; Theuns & Jorissen 1993; Nagae et al. 2004). WRLOF will occur when the acceleration radius of the asymptotic giant branch (now a white dwarf in these barium star systems) star’s stellar wind is comparable to or greater than its Roche lobe radius (Mohamed & Podsiadlowski 2012), meaning that the wind itself is strongly influenced by the binary potential and can be accreted on to the companion via the first Lagrangian point. This accreted material chemically contaminates the surface of the companion as well as increases its spin rate due to the conservation of angular momentum (Theuns, Boffin & Jorissen 1996). Intriguingly, the amount of material required to account for the observed contamination is often in excess of that which would be expected to result in critical rotation rates in the companion, providing an indication that significant (and as yet not understood) angular momentum losses must be experienced in these systems (Matrozis, Abate & Stancliffe 2017). PNe with barium central stars offer an important window into this process as the presence of the short-lived nebula (τ ≤ 30 000 yr; Jacob, Schönberner & Steffen 2013) means that the mass transfer has occurred too recently for significant changes in the companion’s spin rate, and moreover the nebula itself traces the mass-loss history of the system as it is formed from the material that has escaped the binary potential (Jones 2018).
|
[
"Boffin & Jorissen 1988"
] |
[
"The chemical contamination of barium stars is believed to be due to accretion of chemically enriched material from an evolved binary companion, likely through wind"
] |
[
"Background"
] |
[
[
1319,
1341
]
] |
[
[
1154,
1317
]
] |
2015ApJ...803...96S__Török_et_al._2004_Instance_1
|
Since their initial discovery with AIA on board the Solar Dynamics Observatory (SDO), HCs have been generally regarded as proxies for magnetic flux ropes (MFRs; volumetric plasma structures with magnetic field lines that wrap around a central axis). This is supported by the following observational studies: (1) Cheng et al. (2014a) observed an HC that showed helical threads winding around an axis. Simultaneously, cool filamentary materials descended spirally down to the chromosphere, providing direct observational evidence of an intrinsic helical structure for the HC. (2) Cheng et al. (2011) reported that an HC can grow during an eruption, similar to the MFR growth process according to the classic magnetic reconnection scenario in eruptive flares. Song et al. (2014a) presented the formation process of an HC during a coronal mass ejection (CME) and found that the HC was formed from coronal arcades through magnetic reconnection. These works further support the idea that an HC is an MFR structure based on the relation between the HC and magnetic reconnection. (3) Cheng et al. (2014b) found that an HC was initially cospatial with a prominence. Then a separation of the HC top from that of the prominence was observed during the eruption initiated by the ideal kink instability (Török et al. 2004). It is widely accepted that a prominence/filament can exist at the dip of a flux rope (Rust & Kumar 1994). Therefore, this observation offered further important support for the idea that an HC is an MFR; beside an HC, several lines of observations in the lower corona have also been proposed as MFRs, including sigmoid structures in an active region (Titov & Démoulin 1999; McKenzie & Canfield 2008) and coronal cavities in quiescent regions (Wang & Stenborg 2010). A sigmoid has either a forward or reverse S-shape with enhanced X-ray emissions (implying an entity of high temperature) with its center straddling along the polarity inversion line of the hosting active region. Zhang et al. (2012) showed that the HC initially appeared as a sigmoidal structure and then changed to a semi-circular shape. Therefore, a sigmoid and an HC might represent the same structure, and their different shapes are likely from different perspectives and evolution phases. Both structures feature a high temperature, a possible result of flare magnetic reconnection (e.g., Song et al. 2014a, 2014b). A coronal cavity, on the other hand, which is observed as a dark circular or oval structure above the solar limb in coronal images with temperatures close to the background corona (Fuller et al. 2008; Gibson et al. 2010; Kucera et al. 2012), is also interpreted as an MFR. As mentioned, the long-studied feature of solar filaments/prominences shown best in Hα images has been interpreted as being situated along the dip in the MFR. Therefore, a prominence lying in the dip of a coronal cavity is not rare. The eruption of a coronal cavity (or filament) from a quiescent region does not show a high-temperature signature like an HC, which might be attributed to a lack of obvious heating acquired from the weak magnetic reconnection (e.g., Song et al. 2013).
|
[
"Török et al. 2004"
] |
[
"Then a separation of the HC top from that of the prominence was observed during the eruption initiated by the ideal kink instability",
"Therefore, this observation offered further important support for the idea that an HC is an MFR"
] |
[
"Differences",
"Similarities"
] |
[
[
1291,
1308
]
] |
[
[
1157,
1289
],
[
1417,
1512
]
] |
2020AandA...637A..44N__Kraus_(2018)_Instance_2
|
Among the existing IACT systems, HESS has the largest FoV and hence provides the highest sensitivity for the diffuse γ-ray flux. Its electron spectrum analysis technique could be directly used to obtain a measurement of the diffuse Galactic γ-ray flux above energies of several TeV in the Galactic Ridge (|l| 30°, |b| 2°) region; see Figs. 3 and 4. A multi-year exposure of HESS could be sufficient for detection of the diffuse emission even from regions of higher Galactic latitude. This is illustrated in Fig. 5, where thick blue and red data points show mild high-energy excesses of the electron spectra derived by Kraus (2018), Kerszberg et al. (2017), Kerszberg (2017) over broken power-law models derived from the fits to lower energy data. Comparing these excesses with the level of the IceCube astrophysical neutrino flux and with the Fermi/LAT diffuse sky flux from the region |b| > 7° (corresponding to the data selection criterium of HESS analysis Kerszberg et al. 2017; Kerszberg 2017) we find that the overall excess flux levels are comparable to expected diffuse γ-ray flux from the sky region covered by the HESS analysis (the quoted systematic error on the electron flux is Δlog(EFE) ≃ 0.4). The overall excesses within 805 and 1186 h of HESS exposures (Kraus 2018; Kerszberg 2017) are at the levels of >4σ for the analysis of Kraus (2018) and 1.7σ for the analysis of Kerszberg (2017). A factor-of-ten longer exposure (which is potentially already available with HESS) could reveal a higher significance excess at the level of up to 5σ. Such an excess is predicted in a range of theoretical models including interactions of cosmic rays injected by a nearby source (Andersen et al. 2018; Neronov et al. 2018; Bouyahiaoui et al. 2019) or decays of dark matter particles (Berezinsky et al. 1997; Feldstein et al. 2013; Esmaili & Serpico 2013; Neronov et al. 2018) or a large-scale cosmic ray halo around the Galaxy (Taylor et al. 2014; Blasi & Amato 2019).
|
[
"Kraus 2018"
] |
[
"The overall excesses within 805 and 1186 h of HESS exposures"
] |
[
"Uses"
] |
[
[
1272,
1282
]
] |
[
[
1210,
1270
]
] |
2018AandA...612A..34D__Dexter_et_al._2010_Instance_1
|
Sagittarius A* (Sgr A*) is a supermassive black hole system that allows one to observationally test the aforementioned GRMHD models of accretion flows (Goddi et al. 2017). Millimeter-Very Long Baseline Interferometry (mm-VLBI) is capable of resolving the shadow of the event horizon (Falcke et al. 2000), making this an ideal laboratory not only to tests Einstein’s General Theory of Relativity (GR) but also to investigate electron acceleration in the vicinity of a black hole. Most of the radiative models for Sgr A*, which are based on post-processing GRMHD simulations, assume that electrons have a thermal, relativistic (Maxwell–Jüttner) distribution function, and that the proton-to-electron temperature ratio is constant across the simulation domain (Goldston et al. 2005; Noble et al. 2007; Mościbrodzka et al. 2009; Dexter et al. 2010, 2012a; Shcherbakov et al. 2012). When the proton-to-electron temperature is constant, the disk dominates the images and spectra since most of the matter resides there. We have recently extended these radiative models by making the temperature ratios a function of the plasma β parameter, where
$\beta = \frac{P_{\textrm{gas}}}{P_{\textrm{B}}}$
β=
P
gas
P
B
is the ratio of gas to magnetic pressures. In these extended models, the electrons are hotter in the more magnetized plasma, which is usually outflowing from the system. The reason for this is that the previously mentioned models do not recover the flat radio spectra. The β parameterization enforces that the disk emission is suppressed by significantly decreasing the temperature of the electrons in those regions. As a consequence of this, the jet will be the dominant source of emission. These modifications to the electron temperature model allowed us to recover some basic observational characteristics of Sgr A* (a roughly flat radio spectral slope and a size vs. wavelength relationship that is in agreement with observations) (Mościbrodzka & Falcke 2013; Mościbrodzka et al. 2014; Chan et al. 2015b,a; Gold et al. 2017). Our model for the electron temperatures as a function of the β plasma parameter is now roughly confirmed with extended-GRMHD simulations that self-consistently take into account the evolution of the electron temperatures (Ressler et al. 2015, 2017). Moreover, GRMHD simulations with the new electron temperatures can naturally explain the symbiosis of disks and jets observed in many accreting black hole systems (Falcke & Biermann 1995; Mościbrodzka et al. 2016a).
|
[
"Dexter et al. 2010"
] |
[
"Most of the radiative models for Sgr A*, which are based on post-processing GRMHD simulations, assume that electrons have a thermal, relativistic (Maxwell–Jüttner) distribution function, and that the proton-to-electron temperature ratio is constant across the simulation domain"
] |
[
"Background"
] |
[
[
825,
843
]
] |
[
[
479,
756
]
] |
2019ApJ...882..131M__Clements_et_al._2018_Instance_1
|
The ice layer covering refractory grains in dense MCs consists mainly of amorphous water ice (Tielens & Allamandola 1987). A key property of the icy mantle is its porosity, which determines its ability to adsorb, desorb, and trap atoms and molecules. The actual degree of porosity of interstellar ices is still debated. There are indeed indications that the buildup of the ice in cold environments results in the formation of pores as shown by laboratory measurements (e.g., Dohnálek et al. 2003; Bossa et al. 2014) and simulations (Cuppen & Herbst 2007; Clements et al. 2018). These two studies, based on Monte Carlo simulations, suggest a significant level of porosity at the surface/subsurface, in particular in cold and dense environments. On the other hand, UV photons, exothermic reactions, and energetic ions tend to compact astrophysical ices. Estimates of the timescale for mantle compaction calculated from experiments range from a few up to 50 Myr (Raut et al. 2008; Palumbo et al. 2010; Accolla et al. 2011). It has also been suggested that the compaction of icy mantles may not be completed within the typical lifetime of MCs if they consist of ice mixtures, since compaction is slower for mixtures than for pure ices (Palumbo 2006). Transient events like the impact of energetic ions can also generate cavities in the ice surface and/or subsurface. Such energetic events induce a strong local increase of the ice temperature, produce fragments, and are associated with sputtering and evaporation along the ion track. Such cavities are short-lived since the heated ice rearranges while relaxing (Mainitz et al. 2016). The missing O–H dangling features near 3700 cm−1 in astronomical spectra have been taken as an indication of a low level of porosity of interstellar ices. However, laboratory data and simulations indicate that the absence of the O–H dangling modes does not necessarily imply the complete absence of porosity (Raut et al. 2007; Isokoski et al. 2014).
|
[
"Clements et al. 2018"
] |
[
"There are indeed indications that the buildup of the ice in cold environments results in the formation of pores as shown b",
"and simulations",
"These two studies, based on Monte Carlo simulations, suggest a significant level of porosity at the surface/subsurface, in particular in cold and dense environments."
] |
[
"Background",
"Background",
"Background"
] |
[
[
555,
575
]
] |
[
[
320,
442
],
[
516,
531
],
[
578,
743
]
] |
2022MNRAS.514.1169A__Fisher_et_al._1995_Instance_1
|
SHT has some inherent geometrical advantages, especially for large-angle and deep surveys. A significant advantage is that the spherical coordinates apply to wide-angle surveys like BOSS without any flat-sky approximation. A traditional P(k) analysis relies on the flat-sky approximation to distinguish between transverse and line-of-sight modes, which is critical for accurately representing RSDs and distinguishing continuum foregrounds from line signal. By contrast, in the SHT formalism, both foregrounds and linear RSDs take exact, simple forms. This geometric advantage is shared with the related analysis technique of spherical Fourier-Bessel (SFB) decomposition (Fisher et al. 1995; Heavens & Taylor 1995; Leistedt et al. 2012; Rassat & Refregier 2012; Yoo & Desjacques 2013; Lanusse et al. 2015; Liu, Zhang & Parsons 2016; Grasshorn Gebhardt & Doré 2021), which, in addtion to spherical harmonic decomposition, involves a further Fourier-Bessel transform and data-compression along the line-of-sight. However, SFB does not share another important feature of SHT, which is its ability to capture redshift-dependent change over cosmological time in deep surveys. For deep surveys, structure growth and changes in star formation rate break the assumption of translational invariance in the line-of-sight direction, rendering the P(k) or SFB statistic insufficient. However, since Cℓ(z, z′) does not compress data along the line-of-sight direction, it describes redshift evolution. A study (Mondal et al. 2022) of simulated 21-cm data from the EoR found that an implementation of the SHT technique, MAPS (Mondal, Bharadwaj & Datta 2018; Mondal et al. 2019, 2020), obtains ∼2 times more stringent error bars on model parameters than techniques that fail to capture redshift evolution due to data compression along the line-of-sight, such as P(k) or SFB. A final geometric advantage is that Cℓ(z, z′) describes the data in observing coordinates of angle and frequency (or, equivalently, redshift) rather than re-gridding the data on to cosmological distances in an assumed cosmological model. An MCMC likelihood analysis that constrains cosmological parameters would therefore not need to recompute the data statistic at each step, which in principle would be needed for a P(k) or SFB analysis.
|
[
"Fisher et al. 1995"
] |
[
"This geometric advantage is shared with the related analysis technique of spherical Fourier-Bessel (SFB) decomposition",
"which, in addtion to spherical harmonic decomposition, involves a further Fourier-Bessel transform and data-compression along the line-of-sight.",
"However, SFB does not share another important feature of SHT, which is its ability to capture redshift-dependent change over cosmological time in deep surveys."
] |
[
"Similarities",
"Background",
"Differences"
] |
[
[
671,
689
]
] |
[
[
551,
669
],
[
865,
1009
],
[
1010,
1169
]
] |
2022MNRAS.516.3900A__Cazaux_et_al._2022_Instance_1
|
Sudden outbursts of NH3 simultaneously with H2S detected with the ROSINA-DFMS instrument on the Rosetta S/C point to the presence of abundant ammonium hydrosulphide in or on carbonaceous grains from comet 67P/Churyumov-Gerasimenko. There seems to be a clear distinction between the nucleus ice, where H2S and NH3 exist independently and grains, where they desorb together. S2 is much more abundant on grains compared to water than in the ice of the comet, while S3 is found only in grain impacts. This higher abundance points to radiolysis in these grains, which means they must have been exposed to energetic particles over an extended time. While for operational reasons, S4 could not be measured close to the dust impacts, S4 was clearly identified in periods where the coma was very dusty (Calmonte et al. 2016). Longer sulphur chains very likely are refractory, not sublimating at temperatures reached in the instrument or on grains in the coma. While Sn can also be formed from pure H2S ice by photo processing (Cazaux et al. 2022), the fact that S3 is clearly related to dust and is not found in the normal nucleus ice, where H2S is quite abundant, indicates that S3 is a product of radiolysis of the ammonium salt. In addition, photo processing of H2S results not only in Sn, but also in H2S2 (Cazaux et al. 2022), a species not detected in the DFMS m/z 66 and m/z 65 (HS2) spectra. This exposure rules out a contemporary formation of the salt on the surface or in the interior of the comet or a formation of the salt in the mid-plane of the protoplanetary disc, while the comet accreted. A pre-stellar formation is therefore likely. The salt is semivolatile, less volatile than water and could probably have survived quite high temperatures. It seems that on these grains, acids and ammonia are all locked in salts, be it sulphur, halogens, or carbon bearing acids like HOCN. If indeed, a relatively large part of sulphur and nitrogen is therefore in a semivolatile state in these grains, then the depletion of nitrogen in comets and of sulphur in star-forming regions could probably be explained, primarily because salts escape detection unless they experience temperatures above water sublimation. With the JWST S/C in orbit, there is hopefully the possibility to detect salts, or at least several of the acids in ices, which are supposed to be part of ammonium salt, like HOCN, H2CO, and formamide while looking for ammonium salts in star-forming regions and possibly comets.
|
[
"Cazaux et al. 2022"
] |
[
"While Sn can also be formed from pure H2S ice by photo processing",
"the fact that S3 is clearly related to dust and is not found in the normal nucleus ice, where H2S is quite abundant, indicates that S3 is a product of radiolysis of the ammonium salt."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
1018,
1036
]
] |
[
[
951,
1016
],
[
1039,
1222
]
] |
2018AandA...620A..31M__Palla_&_Stahler_1992_Instance_1
|
Any massive stellar source from ~8 M⊙ can burn hot enough to completely ionise the surrounding molecular material to form an H II region (Wood & Churchwell 1989; Churchwell 1990; Kurtz 2005) and destroy any complex chemical tracers that are traditionally used to understand the kinematics of their natal environments and their accretion discs. Initial arguments by Walmsley (1995) argued that high accretion rates will cause a very dense H II region that is optically thick to radio emission and thus the H II would not be seen. However, for heavily accreting massive YSOs the onset of the H II region could be delayed via stellar bloating (Palla & Stahler 1992; Hosokawa & Omukai 2009; Hosokawa et al. 2010; Kuiper & Yorke 2013) where the effective temperature of the star is much cooler than it would be considering a main sequence star of the same luminosity. In these models, a halt or considerable reduction in accretion (≪10−3 M⊙ yr−1), or growth beyond ~30−40 M⊙, will result in the YSO contracting to a “main-sequence” configuration, heating significantly, and being able to create an H II region. The fine details of this transition are still somewhat unclear since they depend on the assumed accretion law and the initial conditions chosen for the stellar evolution calculations (Haemmerlé & Peters 2016). Observations in the infra-red (IR) have been made in search of cool stellar atmospheres that point to bloated stars, however these studies remain inconclusive (Linz et al. 2009; Testi et al. 2010). Alternative scenarios are that H II regions can be gravitationally trapped at very early ionisation stages (see Keto 2003, 2007), or flicker due to chaotic shielding of the ionising radiation by an accretion flow, leading to a non-monotonous expansion (Peters et al. 2010b). Hyper Compact (HC) H II regions (0.03 pc) are thought to be the earliest ionisation stage and therefore could relate to the halt of accretion, be a marker of a transition phase.
|
[
"Palla & Stahler 1992"
] |
[
"However, for heavily accreting massive YSOs the onset of the H II region could be delayed via stellar bloating",
"where the effective temperature of the star is much cooler than it would be considering a main sequence star of the same luminosity. In these models, a halt or considerable reduction in accretion (≪10−3 M⊙ yr−1), or growth beyond ~30−40 M⊙, will result in the YSO contracting to a “main-sequence” configuration, heating significantly, and being able to create an H II region."
] |
[
"Background",
"Background"
] |
[
[
641,
661
]
] |
[
[
529,
639
],
[
730,
1105
]
] |
2019AandA...625A.121M__Beaugé_&_Nesvorný_2012_Instance_2
|
The final location of close-in giant planets in our models reflects the strength of the tides that we include in our modeling, which play a very important role in the decay of planetary orbits. These are dynamical tides (e.g., Lai 1997; Ivanov & Papaloizou 2004, 2007, 2011) and in our simulations we used a formulation given by Ivanov & Papaloizou (2007) as described inSect. 2. However, the impulse approximation used in the evaluation of dynamical tides becomes a poor approximation when the circularization proceeds (e.g., Mardling 1995a,b) and the eccentricity becomes low. Equilibrium tides become then effective (e.g., Beaugé & Nesvorný 2012 and references therein) and the tidal evolution may occur on a longer timescale. In short, at the beginning of the orbital evolution that leads to the formation of hot/warm Jupiters, dynamical tides are important in forcing the decay of the orbit. In the last part of the dynamical evolution when the eccentricity has become low, equilibrium tides are more important in determining the location where the planet stops. Unfortunately, at present it is not known when and how the two tides switch. When we change the magnitude of two tides, the final location of the planets can be adjusted (Beaugé & Nesvorný 2012). However, rather than introducing artificial effects, we continue to use dynamical tides in our simulations even for low eccentricities but we stop our simulations when the energy, decreasing from the tide at the pericenter, overcomes the orbital energy leading to a clustering of tidally circularized planets around 0.02 au. However, the final distribution of the inclination of the planets does not depend on this choice and highly misaligned planets would be produced anyway. Since the tidal evolution of planets with arbitrary inclinations is still not well known, we assume that planetary inclination is not significantly changed during tidal evolution (Barker & Ogilvie 2009). Thus, the planets maintain the inclination they have when the circularization begins.
|
[
"Beaugé & Nesvorný 2012"
] |
[
"When we change the magnitude of two tides, the final location of the planets can be adjusted"
] |
[
"Uses"
] |
[
[
1239,
1261
]
] |
[
[
1145,
1237
]
] |
2020MNRAS.491.5073P__Sutherland_&_Saunders_1992_Instance_1
|
The catalogue used in this work is based on a far-IR sample selected in the ∼2 deg2-wide COSMOS field and obtained within the Herschel-PEP survey (Lutz et al. 2011). We consider the latest released blind catalogue selected at 160-μm (DR1, 7047 sources) with >3σ flux density, corresponding to a flux limit of ∼9.8 mJy. The choice of considering as parent sample a far-IR catalogue is guided by the necessity of having several detections at different wavelengths for each system to constrain the dust masses, and yet with a very simple selection function (see Section 3.1). From the original 160-μm selection, we built a multiband catalogue taking advantage of the extensive multiwavelength coverage in the COSMOS field. Concerning the other far-IR PACS band (100 μm) and the mid-IR 24-μm band, we use the association available in the DR1 release and based on the maximum likelihood technique (Sutherland & Saunders 1992; Ciliegi et al. 2001). For the cross-match with the SPIRE far-IR bands (250, 350, and 500 μm), we used the same catalogue considered in previous PACS-based works (i.e. Gruppioni et al. 2013; Delvecchio et al. 2015), the ones provided by the HerMES collaboration (Roseboom et al. 2010) using the Spitzer-MIPS 24-μm positions as priors to extract the SPIRE fluxes. Finally, the IRAC/optical/UV fluxes taken from the COSMOS2015 catalogue (Laigle et al. 2016) were merged to the PACS-160-μm sample by matching the 24-μm counterparts listed in both. The COSMOS2015 is NIR selected, where objects have been detected from the sum of the UltraVISTA-DR2 YJHK and z++ images. By construction, in comparison to the previous i-selected catalogue, this catalogue is missing a fraction of blue, faint, star-forming galaxies (Laigle et al. 2016). For this reason, we decided to cross-match the far-IR sources with no counterparts in the COSMOS2015 catalogue with the Ilbert et al. (2009) i-selected catalogue. Totally, among the 160-μm selected sources (7047), 6002 are with 24-μm counterparts (${\sim }86{{\ \rm per\ cent}}$), of which 5993 with available NIR or optical counterparts (${\sim }99.9{{\ \rm per\ cent}}$, 5783 in the COSMOS2015 and 210 in the Ilbert et al. 2009 catalogue). While the cross-matching with the optical/NIR bands does not involve almost any source loss, a moderate (14${{\ \rm per\ cent}}$) but not negligible fraction of the far-IR sources does not have 24-μm counterparts. A fraction of these sources are likely spurious sources, as shown by the simulations done for the DR1 release PEP catalogue (∼5 per cent at the 3σ flux level). In Section 3.1, we will describe our method to correct for incompleteness and for the presence of spurious systems.
|
[
"Sutherland & Saunders 1992"
] |
[
"Concerning the other far-IR PACS band (100 μm) and the mid-IR 24-μm band, we use the association available in the DR1 release and based on the maximum likelihood technique"
] |
[
"Uses"
] |
[
[
893,
919
]
] |
[
[
720,
891
]
] |
2020AandA...638A..16T__Barnes_(2017)_Instance_1
|
Figure 12 shows the results of our tidal evolution calculations. The left panel of Fig. 12 shows the planetary rotational evolution of GJ 1148 b due to star–planet tides. After ~850 Myr, GJ 1148 b reaches a rotation period that is 2∕3 of the orbital period, and remains there with Prot = 27.5 d. During the integration the planetary semi-major axis and eccentricity are mostly unaffected. An asymptotic rotation period that is shorter than synchronous and 2/3 of the orbital period is expected for eb ≳0.24 in the constant Q tidal model (Goldreich & Peale 1966; Cheng et al. 2014). The time for GJ 1148 b to reach asymptotic rotation is inversely proportional to the initial Prot, as long as the initial Prot is much less than 27.5 d, and it depends on the other parameters of GJ 1148 b according to Eq. (3) of Barnes (2017) and Eq. (15) of Cheng et al. (2014). The rotational period of GJ 1148 b is thus very likely much longer than the orbital periods of the hypothetical exomoons, which could be dynamically stable only with orbital periods between 0.7 and 2 d. The right panel of Fig. 12 shows that the longer rotational period of GJ 1148 b (Prot = 27.5 d) leadsto strong orbital decay of the stable exomoon orbits due to tidal interactions with the planet. An exomoon eventually reaches the Roche limit where it is tidally disrupted by the gas giant. Not even one hypothetical “stable” exomoon in the context of Sect. 5.3.1 had survived this test. The maximum time a Mars-like exomoon could survive is ~55 M yr, while for Titan-like moons the maximum survival time is longer, ~255 M yr. The latter is longer by roughly the mass ratio of Mars to Titan, which can be understood from Eq. (2) of Barnes (2017) and Eq. (16) of Cheng et al. (2014). These timescales are optimistic since the orbital decay would start before the planet reaches the asymptotic spin state. In both cases the survival times are much shorter than the age of the system. Therefore, given the relatively fast orbital decay in the small stable region around the planet, we conclude that exomoons around GJ 1148 b are unlikely to exist.
|
[
"Barnes (2017)"
] |
[
"The time for GJ 1148 b to reach asymptotic rotation is inversely proportional to the initial Prot, as long as the initial Prot is much less than 27.5 d, and it depends on the other parameters of GJ 1148 b according to Eq. (3) of"
] |
[
"Uses"
] |
[
[
811,
824
]
] |
[
[
582,
810
]
] |
2019ApJ...883...76R__Strateva_et_al._2005_Instance_1
|
Previous observations of AGN that investigate correlations between αOX and Eddington ratio have revealed some similarities with X-ray binary outbursts at high Lbol/LEdd, but these comparisons have not been possible below the critical Lbol/LEdd ≲ 10−2, where an inversion in this correlation is predicted to occur. At higher Eddington ratios of Lbol/LEdd ≳ 10−2, single-epoch X-ray and UV observations of large samples of AGN have previously revealed a hardening of αOX as Lbol/LEdd drops from ∼1 to ∼10−2 (e.g., Vignali et al. 2003; Strateva et al. 2005; Steffen et al. 2006; Just et al. 2007; Grupe et al. 2010; Jin et al. 2012; Wu et al. 2012; Trichas et al. 2013; Vagnetti et al. 2013). This correlation was also observed in multi-epoch UV/X-ray observations of the fading of Mrk 1018 (Noda & Done 2018), which confirms this behavior in an individual AGN. However, the predicted softening of αOX below Lbol/LEdd ≲ 10−2 (thus causing an inversion in the correlation between αOX and Lbol/LEdd) has not been previously observed. This is primarily due to the difficulty of robustly measuring both αOX and Lbol/LEdd for AGN below Lbol/LEdd ≲ 10−2, for three main reasons. First, at low Eddington ratios, AGN are often dust-obscured (Fabian et al. 2008), and thus measuring their intrinsic UV luminosities (and αOX) is difficult. Second, broad emission lines often disappear in low-luminosity AGN below Lbol/LEdd ≲ 10−2, making it difficult to measure MBH (and LEdd). Third, using a sample of AGN with a wide range of Eddington ratios to trace how αOX changes as a function of Lbol/LEdd can be hampered by the
scaling of the thin disk temperature with MBH at a fixed Eddington ratio. If the AGN sample has a large range in MBH, this can cause an additional scatter in αOX. Thus, we would ideally use a sample of AGN with a narrow range in MBH, but the difficulty of measuring MBH at Lbol/LEdd ≲ 10−2 also hampers the construction of such a sample. In this paper, we will use a new method to bypass all of these issues, with the goal of extending this spectral comparison between X-ray binaries and AGN to Lbol/LEdd ≲ 10−2.
|
[
"Strateva et al. 2005"
] |
[
"At higher Eddington ratios of Lbol/LEdd ≳ 10−2, single-epoch X-ray and UV observations of large samples of AGN have previously revealed a hardening of αOX as Lbol/LEdd drops from ∼1 to ∼10−2 (e.g.,"
] |
[
"Background"
] |
[
[
533,
553
]
] |
[
[
314,
511
]
] |
2022MNRAS.513.5245A__Done_&_Jin_2016_Instance_1
|
We assume the time-scales we observe here are generated in the corona itself (and note that longer time-scale changes will be driven by the disc outside of the corona) and are made visible by a changing electron temperature and density as a result of local turbulence and coupling to mass accretion rate propagations through the flow from rout to rin. We note that we can discount variations in the seed photon population as the driver for changes in the power spectrum, as the UV emission from the disc is established to be considerably less variable than the corona in NLS1s (Leighly 1999; Smith & Vaughan 2007; Ai et al. 2013; Alston, Vaughan & Uttley 2013; Done & Jin 2016). We assume that the variability generated locally at each radius (rν) is at the viscous frequency (see Churazov et al. 2001) such that
(9)$$\begin{eqnarray}
r_{\nu } = \left[ \frac{2\pi \nu }{\alpha }\left(\frac{H}{R}\right)^{-2}\right]^{-2/3} ,
\end{eqnarray}$$where the frequency is in units of c/Rg (see e.g. Kato, Fukue & Mineshige 1998; Arévalo & Uttley 2006). In the above, α and $\frac{H}{R}$ are the dimensionless viscosity parameter of Shakura & Sunyaev (1973) and scale height of the accretion disc, respectively. We assume that our frequency range of interest, 0.01−1 mHz (i.e. the range over which we can practically fit to the data) corresponds to radii between the ISCO (rin) and some radius within the true outer edge of the corona (i.e. rout ≤ rcorona). The actual frequencies generated in our model therefore depend on the SMBH spin and the combination $\left(\frac{H}{R}\right)^2 \alpha$ (and somewhat on the SMBH mass – although here the range is small). Given the reported high spin values for these bright AGNs (Ogle et al. 2004; Fabian et al. 2013; Done & Jin 2016; Kara et al. 2017; Buisson et al. 2018b), we expect the ISCO to sit at ∼1.25Rg. For the corona at the ISCO to produce variability above our upper frequency limit of 1 mHz requires $\left(\frac{H}{R}\right)^2 \alpha \gtrsim 0.01$. We note that for the mass range subtended by our AGN sample (from 106.00 to 106.63 M⊙, see Table 2), should we instead assume zero spin (rin = 6Rg), the viscous frequency at rin is lower and we have strong curvature in our observed bandpass.
|
[
"Done & Jin 2016"
] |
[
"We note that we can discount variations in the seed photon population as the driver for changes in the power spectrum, as the UV emission from the disc is established to be considerably less variable than the corona in NLS1s"
] |
[
"Uses"
] |
[
[
661,
676
]
] |
[
[
352,
576
]
] |
2018MNRAS.478...95K__Toalá_et_al._2012_Instance_1
|
We subsequently considered the time evolution of a number of selected cores (based on the requirement that we have sufficient data to follow at least 200 kyr of evolution), finding a remarkably similar chemical evolution in all cores, including one that is strongly stabilized by turbulence and magnetic field and therefore not going through gravitational collapse. Assuming a more or less chemically homogeneous initial condition, the chemical evolution of the cores, particularly regarding deuteration effects, is thus very similar. Of course, environmental effects such as the local cosmic ray ionization rate could induce potential differences, by changing the ionization degree. Similarly, metallicity effects at lower densities may depend on the local conditions. At least within one filament, it is however plausible that the deuteration fraction is indeed indicative of chemical age. We find here that about two free-fall times (as defined for cylindrical systems, see Toalá et al. 2012) are sufficient to reach core deuteration fractions of ≳0.1. We finally investigated also the radial structure of the core, finding overall similar properties as in the isolated collapsing cores studied by Körtgen et al. (2017). The H2 ortho-to-para ratio appears to be approximately flat and only weakly dependent on radius. Both the deuteration fraction and the gas surface density show a peak on scales of about 1000–2000 au, which is particularly pronounced in the case of parallel magnetic fields. As found previously, this peak moves outward with increasing turbulent Mach number, indicating the amount of support against gravity. The difference in the visibility of the peak may result from the difference in the fragmentation mode in both cases, as previously described by Seifried & Walch (2015), thus potentially affecting the structure of the resulting cores. Overall, our results have shown that the observed high deuteration fraction in prestellar cores can be readily reproduced in simulations of turbulent magnetized filaments. We further found that deuteration fractions of order 0.1 can be produced independent of the specific history of the cores, both for high and low virial parameters. The latter suggests that deuteration is potentially very efficient.
|
[
"Toalá et al. 2012"
] |
[
"We find here that about two free-fall times (as defined for cylindrical systems, see",
"are sufficient to reach core deuteration fractions of ≳0.1."
] |
[
"Uses",
"Uses"
] |
[
[
977,
994
]
] |
[
[
892,
976
],
[
996,
1055
]
] |
2021MNRAS.504.5074S__Maraschi,_Ghisellini_&_Celotti_1992_Instance_1
|
The broad-band spectral energy distribution (SED) of blazars is characterized by two broad humps, one at optical/UV/X-ray bands and the other in the HE γ-ray band (see Padovani et al. 2017 for a recent review). It is believed that the first peak (low-energy component) is mostly due to synchrotron emission from relativistic electrons, whereas the origin of the second component is highly debatable. Within conventional leptonic scenarios, this component is produced when the synchrotron-emitting electrons inverse Compton up scatter the photons of internal (synchrotron self-Compton, SSC; Ghisellini, Maraschi & Treves 1985; Maraschi, Ghisellini & Celotti 1992; Bloom & Marscher 1996) or external (external inverse Compton, EIC; Sikora, Begelman & Rees 1994; Błażejowski et al. 2000; Ghisellini & Tavecchio 2009) origin. The nature of the external photon fields depends on the distance of the emitting region from the central black hole (Sikora et al. 2009) and can be dominated either by the photons directly emitted from the accretion disc (Dermer, Schlickeiser & Mastichiadis 1992; Dermer & Schlickeiser 1993) or disc photons reflected from the broad-line region (BLR; Sikora et al. 1994) or IR photons emitted from the dusty torus (Błażejowski et al. 2000). Recently, after associating TXS 0506+056 with the IceCube-170922A neutrino event (IceCube Collaboration et al. 2018a,b; Padovani et al. 2018), it is more evident that the HE component could be initiated by the interaction of energetic protons when they are effectively accelerated in the blazar jets. The HE component can be either from proton synchrotron emission (Mücke & Protheroe 2001) or from secondary particles from pion decay (Mannheim & Biermann 1989; Mannheim 1993; Mücke & Protheroe 2001; Mücke et al. 2003; Böttcher et al. 2013). In the latter case, blazars are also sources of very high energy neutrinos (Ansoldi et al. 2018; Keivani et al. 2018; Murase, Oikonomou & Petropoulou 2018; Padovani et al. 2018; Sahakyan 2018; Cerruti et al. 2019; Gao et al. 2019; Righi, Tavecchio & Pacciani 2019; Sahakyan 2019).
|
[
"Maraschi, Ghisellini & Celotti 1992"
] |
[
"Within conventional leptonic scenarios, this component is produced when the synchrotron-emitting electrons inverse Compton up scatter the photons of internal (synchrotron self-Compton, SSC;"
] |
[
"Compare/Contrast"
] |
[
[
626,
661
]
] |
[
[
400,
589
]
] |
2015MNRAS.446.2468E__Kauffmann_&_Haehnelt_2000_Instance_1
|
In the following, all measurements and maps derived from snapshots prior to t = 800 Myr are therefore using the original simulation (R+13), with the others extracted from the simulation at 1 pc resolution (for the gaseous component). Note that star formation was turned on at t ∼ 745 Myr in the simulation as to avoid gas being prematurely consumed. Details on the implemented recipes for star formation, stellar feedback (photoionization, radiative pressure, supernova explosions) are described in R+13. Adding on AGN feedback would significantly impact on the distribution, kinematics and physical status of the gas, specifically for the close environment of the black hole (Ciotti & Ostriker 1997; Haehnelt, Natarajan & Rees 1998; Silk & Rees 1998; Kauffmann & Haehnelt 2000). For the present simulations, however, we do not include the potential feedback from an AGN, thus focusing on a time window (t ∼ 750–830 Myr) when we consider that the AGN itself is quiet (or in an ‘off-state’). This is partly justified by the assumption that AGN have low duty cycle at low redshift and for black holes of a few 106 M⊙ (Haehnelt & Rees 1993; Wang et al. 2009; Shankar et al. 2010; Shankar, Weinberg & Miralda-Escudé 2013) and by the short time range we are considering. More importantly, it allows us to narrow down our study to probe the interplay between the dynamical evolution and the effect of star formation (similarly to e.g. Levine et al. 2008; Hopkins & Quataert 2010a). Turning on the AGN in such a simulation would be paramount to understand any potential fuelling cycle starting from the large-scale down to the vicinity of the black hole, and such an implementation has already been included in ramses by a direct calculation of the Bondi accretion rate (see Teyssier et al. 2011; Gabor & Bournaud 2014, and references therein). It would nevertheless require to probe various feedback schemes, and triggering mechanisms, which is beyond the scope of the present paper.
|
[
"Kauffmann & Haehnelt 2000"
] |
[
"Adding on AGN feedback would significantly impact on the distribution, kinematics and physical status of the gas, specifically for the close environment of the black hole"
] |
[
"Future Work"
] |
[
[
752,
777
]
] |
[
[
505,
675
]
] |
2018MNRAS.476L...6R__Cyr_et_al._2000_Instance_1
|
The coronal mass ejections (CMEs) are frequent discharge of huge energy and massive magnetized plasma from the solar corona into the heliosphere. They are of paramount importance in space physics for their key role in extreme space weather and geo-effectiveness, e.g. (Gosling 1993; Low 2001; Schrijver & Siscoe 2010; Cannon et al. 2013). In last few decades, the understanding of CMEs improved significantly because of space and ground-based observational data with the help of various modelling efforts. The studies are focused on the morphological and kinematic evolution of CMEs in the heliosphere, e.g. (Lindsay et al. 1999; St Cyr et al. 2000; Zurbuchen & Richardson 2006; Chen 2011; Webb & Howard 2012; Wang et al. 2016; Lugaz et al. 2017). By considering the number of CMEs emitted from the Sun during solar maximum and variations in their respective speeds, the interaction between multiple CMEs in the heliosphere is expected to be more frequent. The collision of multiple CMEs highly affect their dynamic evolution properties and contribute to enhanced geo-effectiveness, e.g. (Wang, Wang & Ye 2002; Farrugia & Berdichevsky 2004; Lugaz, Manchester IV & Gombosi 2005; Wang et al. 2005; Xiong et al. 2007; Shen et al. 2011, 2012; Lugaz et al. 2012; Temmer et al. 2012). To predict space weather effects near the Earth, an accurate estimation of arrival time of CMEs at the Earth is crucial (Mishra & Srivastava 2014). Besides this, the study of CME–CME and CME–solar wind interactions provide unique observational evidences to understand energy dissipation of large-scale magnetic clouds in interstellar medium and authenticate the physical processes predicted theoretically. Therefore, interaction of multiple CMEs needs to be examined in detail. The various results obtained from studies have justified CME–CME collision as an inelastic/elastic collision or superelastic collision, e.g. (Lugaz et al. 2012; Shen et al. 2012, 2016; Lugaz et al. 2017; Mishra et al. 2017). The magnetohydrodynamics (MHD) numerical simulations have striven to understand the physical mechanism involved in CME–CME interaction, CME–CME driven shock interactions and their consequences, e.g. (Niembro et al. 2015; Jin et al. 2016; Shen et al. 2016; Wu et al. 2016).
|
[
"St Cyr et al. 2000"
] |
[
"In last few decades, the understanding of CMEs improved significantly because of space and ground-based observational data with the help of various modelling efforts. The studies are focused on the morphological and kinematic evolution of CMEs in the heliosphere, e.g."
] |
[
"Background"
] |
[
[
630,
648
]
] |
[
[
339,
607
]
] |
2020MNRAS.499.4666F__Michałowski_2015_Instance_1
|
An example of these implications is the so-called ‘dust budget crisis’ introduced in Section 4.4: the dust masses currently estimated at z > 5 are not compatible with standard dust production channels and require an overhaul in our models of the initial mass function for star formation, of supernova production rates, or of dust growth in the ISM. Overall, the dust production rate would need to increase by one to two orders of magnitudes, as shown by Rowlands et al. (2014). The growth of dust grains through accretion in the ISM has been proposed as a solution (e.g. Mancini et al. 2015; Michałowski 2015; Popping et al. 2017), but there are doubts on the efficiency of accretion at high z, where high dust temperatures due to the CMB (see Section 3.3) keep the desorption time-scale for accreted materials short (Ferrara et al. 2016). The dust budget crisis is not only a problem at high redshift; it is observed, e.g. in the Magellanic Clouds (SMC, LMC). As explained in Srinivasan et al. (2016) using the dust mass fits by Gordon et al. (2014), the dust replenishment time-scale in the SMC from stellar sources alone is expected to be larger than the dust destruction time-scale and, in the worst-case scenario, longer than the lifetime of the Universe. Similarly, the ratio between the best LMC dust mass estimate by Gordon et al. (2014) and the dust injection estimates by Riebel et al. (2012) results in an LMC replenishment time-scale of 34 ± 8 Gyr, exceeding the age of the Universe. Both the high redshift and the local Universe, therefore, show a dust budget crisis that could be alleviated – and, in the best case scenario, fully resolved – if the actual dust masses turned out to be lower than currently estimated, as our results suggest. More specifically, Rowlands et al. (2014) mention that dust opacity needs to be increased by just a factor of 7 to solve the high-redshift crisis (provided dust destruction by SNe is not efficient); in the LMC, the aforementioned replenishment time-scale would decrease to less than 2 Gyr if the dust mass were decreased by a factor of 20.
|
[
"Michałowski 2015"
] |
[
"The growth of dust grains through accretion in the ISM has been proposed as a solution (e.g."
] |
[
"Motivation"
] |
[
[
592,
608
]
] |
[
[
478,
570
]
] |
2022ApJ...934..103H__Moore_et_al._2001_Instance_1
|
Compared to various studies based on in situ measurements of MFR structures after eruption, the origination of CME-MFRs before and during eruptions remains elusive due to the complex environment in the solar source region and limited observations. At the present time, there are certain hypotheses on the formation process of MFRs. Some studies indicate that MFRs could exist prior to the eruption. For example, both Fan (2001) and Magara (2004) reported findings from magnetohydrodynamic (MHD) simulation that a twisted MFR initially formed below the photosphere can partially emerge into the low corona by magnetic buoyancy. Other studies suggest that the presence of preeruptive MFRs is not necessary and MFRs could be built up in the corona via magnetic reconnection processes associated with flares (Amari et al. 2003; Moore et al. 2001; Antiochos et al.1999; Jiang et al. 2021a, 2021b). To understand the physical processes more precisely for flare–CME events, extensions of the standard 2D flare model have been proposed to account for much broader ranges of quantitative measurements with three-dimensional (3D) features intrinsic to realistic solar eruptions (Longcope et al. 2007; Aulanier et al. 2012; Priest & Longcope 2017; Aulanier & Dudík 2019). For example, quasi-3D models have been developed with a nonvanishing magnetic field component along the axis of the MFR and to illustrate the scenario that sequential reconnection along the magnetic polarity inversion line (PIL) forms the MFR in the first place (van Ballegooijen & Martens 1989; Longcope et al. 2007; Schmieder et al. 2015). This scenario has been widely applied to infer and interpret magnetic reconnection properties based on the observed flare ribbon morphology (Qiu et al. 2002, 2004, 2010; Hu et al. 2014; Kazachenko et al. 2017; Zhu et al. 2020). From such analyses, Qiu et al. (2004) illustrated that there is a temporal correlation between the magnetic reconnection rate and the acceleration of the CME (considered as the eruptive MFR) in the low corona. Such a correlation has been further established by Zhu et al. (2020) based on a statistical study of ∼60 events. In addition, Qiu et al. (2007) and Hu et al. (2014) showed a correlation between the magnetic reconnection flux and the flux contents of the corresponding ICME/MC flux ropes based on modeling results employing in situ spacecraft measurements. These results support the hypothesis that CME-MFRs can be formed by magnetic reconnection during the corresponding flare process. Recent simulation results also indicate clearly that the reconnection flux contributes to the axial (toroidal) flux of the CME-MFR in the early stage (Jiang et al. 2021a; Inoue et al. 2018).
|
[
"Moore et al. 2001"
] |
[
"Other studies suggest that the presence of preeruptive MFRs is not necessary and MFRs could be built up in the corona via magnetic reconnection processes associated with flares"
] |
[
"Motivation"
] |
[
[
824,
841
]
] |
[
[
627,
803
]
] |
2021MNRAS.508.4767S__Chiaki,_Yoshida_&_Hirano_2016_Instance_1
|
Although it is still unknown why the disc fragmentation for the primordial cases is well described by the simple relation such as equation (1), an important fact is that a barotropic EOS with γeff ≃ 1.1 approximately represents the gas thermal evolution during the cloud collapse for $n \lesssim 10^{19}\, \mathrm{cm}^{-3}$ (Omukai & Nishi 1998). In fact, several simulations study the disc fragmentation assuming the same barotropic EOS with γeff = 1.1 for n ≤ nth, resulting in the evolution described by equation (1) (Susa 2019). Since nth is only the characteristic quantity for this case, the simple scaling of equation (1) may be convincing. This suggests that the disc fragmentation with a different EOS should provide different evolution of Nc,b. For instance, it is well known that adding a tiny amount of heavy elements and dust grains alters the EOS of a collapsing cloud (e.g. Omukai 2000; Bromm et al. 2001; Omukai et al. 2005; Schneider et al. 2006, 2012a; Smith, Sigurdsson & Abel 2008; Jappsen et al. 2009; Safranek-Shrader, Milosavljević & Bromm 2014; Chiaki et al. 2015; Chiaki, Yoshida & Hirano 2016). While previous studies demonstrate that the dust cooling enhances the fragmentation during the cloud collapse (e.g. Meece, Smith & O’Shea 2014; Smith et al. 2015; Chiaki & Wise 2019), its effect on the disc fragmentation remains to be further explored. Tanaka & Omukai (2014) investigate the evolution of the circumstellar disc in metal-poor environments developing 1D semi-analytical models. They predict that the discs with $Z \sim 10^{-5}\!-\! 10^{-3}\, \mathrm{Z}_{\odot }$ are subject to the efficient dust cooling and are more unstable than those for the primordial cases. Machida & Nakamura (2015) consider the disc fragmentation with various metallicities $0 \le Z \le 1\, \mathrm{Z}_{\odot }$, performing a suite of 3D numerical simulations. They find qualitative differences between the cases with $Z \lesssim 10^{-4}\, \mathrm{Z}_{\odot }$ and $Z \gtrsim 10^{-4}\, \mathrm{Z}_{\odot }$; the vigorous disc fragmentation only occurs for the former metal-poor cases. Whereas Machida & Nakamura (2015) use the metallicity-dependent barotropic EOS, Chiaki & Yoshida (2020) recently report 3D simulations of the disc fragmentation solving the energy equation with relevant thermal processes coupled with a non-equilibrium chemical network. They find that for the metal-poor cases with $Z \le 10^{-3}\, \mathrm{Z}_{\odot }$, the disc fragmentation does not necessarily prevent the mass growth of the most massive protostar as many clumps are short-lived owing to the frequent merger or tidal disruption events. Although the above studies suggest the metallicity-dependence of the disc fragmentation, they both only follow the short-term evolution for $\sim 100\, \mathrm{yr}$ since the first emergence of a protostar.
|
[
"Chiaki, Yoshida & Hirano 2016"
] |
[
"For instance, it is well known that adding a tiny amount of heavy elements and dust grains alters the EOS of a collapsing cloud (e.g."
] |
[
"Background"
] |
[
[
1089,
1118
]
] |
[
[
755,
888
]
] |
2020AandA...641A.118F__Nakajima_et_al._(2018b)_Instance_1
|
Comparison with z ≈ 2−4galaxies. High-ionization UV lines have been detected in the spectra of z ≳ 2 galaxies through gravitational lensing (e.g., Stark et al. 2014; Patrício et al. 2016; Vanzella et al. 2016, 2017; Berg et al. 2018), spectral stacking (e.g., Nakajima et al. 2018b; Rigby et al. 2018; Saxena et al. 2020), and deep spectroscopic observations (e.g., Erb et al. 2010; Maseda et al. 2017; Amorín et al. 2017; Nanayakkara et al. 2019). The EWs of He II, O III]λ1666, and C III] from these works are larger than those measured in the average spectra of our LAEs, with the exception of some of the MUSE sources at z ≳ 3 studied by Patrício et al. (2016) and Nanayakkara et al. (2019), some of the stacks from Nakajima et al. (2018b), and some of the VANDELS He II emitters from Saxena et al. (2020). The EWs from Nakajima et al. (2018b) are computed from average spectra of a z ≈ 3 LAE population whose median UV luminosity is about two orders of magnitude brighter than ours. Most of the 2.4 z 3.5 UV-selected, low-luminosity galaxies from Amorín et al. (2017) have higher He II and C III] EWs, most likely because of their higher SFR (see Fig. 2). The line measurements from Saxena et al. (2020) are performed on the stacked spectra of UV continuum, bright He II emitters (−19 MUV −22). The strong line emitters from Erb et al. (2010) and Berg et al. (2018) are brighter and more massive than the median value of our LAEs. The lensed galaxies from Vanzella et al. (2016, 2017) are among the least massive, most metal-poor, young, and faintest systems observed at z ∼ 3. With MUV > −16, the source ID14 from Vanzella et al. (2017) is roughly one order of magnitude fainter that the faintest LAE in our sample, while the source ID11 from Vanzella et al. (2016) has an remarkable blue UV slope (β = −2.95). Recently, Du et al. (2020) measured C III] EWs in z ∼ 2 analogs of galaxies in the reionization era, obtaining values from 13.2 Å down to 1 Å, reaching values as low as those of our LAEs. The authors also found differences in the C III] EW depending on the selection criteria, with higher values of C III] EW for emission lines rather than for continuum-selected galaxies. Mainali et al. (2020) detected two targets with C III] emission reaching EW ≈ 17 − 21 Å in the spectra of z ∼ 2 galaxies selected for their strong rest-optical line ([O III]λ5007+Hβ) EWs. These values are above the ones of our LAEs and are similar to those observed at z > 6 (e.g., Stark et al. 2015b, 2017; Hutchison et al. 2019).
|
[
"Nakajima et al. 2018b",
"Nakajima et al. (2018b)",
"Nakajima et al. (2018b)"
] |
[
"High-ionization UV lines have been detected in the spectra of z ≳ 2 galaxies through",
"spectral stacking",
"The EWs of He II, O III]λ1666, and C III] from these works are larger than those measured in the average spectra of our LAEs, with the exception of",
"some of the stacks from",
"The EWs from",
"are computed from average spectra of a z ≈ 3 LAE population whose median UV luminosity is about two orders of magnitude brighter than ours."
] |
[
"Background",
"Background",
"Similarities",
"Similarities",
"Differences",
"Differences"
] |
[
[
260,
281
],
[
720,
743
],
[
824,
847
]
] |
[
[
33,
117
],
[
235,
252
],
[
449,
596
],
[
696,
719
],
[
811,
823
],
[
848,
987
]
] |
2021ApJ...908...45B__Eyink_&_Sreenivasan_2006_Instance_1
|
Here we shall set the potential
and make the change of variable
; hence, the vortical parts of the footpoint equations of motion are cast into a Hamiltonian form. The Hamiltonian describing the dynamics of magnetic footpoints driven by the vortical component of the turbulence on the photosphere is
50
The footpoint Hamiltonian
has the physical dimension of a frequency (s−1). From Figure 2 in Rincon et al. (2017), we further notice that the vortical component of the photospheric surface velocity field has a structure reminiscent of 2D Euler turbulence, which admits a vortex point representation (Eyink & Sreenivasan 2006) also on the sphere (Pavlov et al. 2001). There are three important parameters that can be extracted from the vortical component of photospheric turbulence. These parameters are the typical amplitude of the velocity fluctuations
, the correlation length
, and the correlation time
of the turbulence, where
is the rms amplitude of the fluctuating stream function Φ and
is the angular size of the turbulent eddies on the source surface. These three independent parameters
,
, and
can all be derived from a more fundamental quantity, which is the covariance function
, i.e., the autocorrelation function of the turbulent stream function (Creasey & Lang 2018). The covariance function is sufficient to characterize the vortical component of the surface flows provided that the stream function Φ does not significantly deviate from a stationary isotropic Gaussian random field on the sphere. Hamilton equations with the Hamiltonian (50) determine a genuine problem of chaotic mixing (Ott 1993; Morrison 1998) on the sphere. We emphasize that mixing of footpoints on the photosphere relies on the time dependence of the Hamiltonian. In fact, the footpoint dynamics is regular (quasi-periodic) when the Hamiltonian is a constant of time, independently of the number of modes involved in its decomposition. The reason is that the dynamics generated by 1 degree of freedom Hamiltonians, here for the pair of canonically conjugate variables μ and ϕ, is regular. A simple mathematical proof of this statement is given by Morrison (1998). In one version of the protocol used by Giacalone & Jokipii (2004) to model the magnetic footpoint dynamics, the time dependence of the flow results from the (random) phase of the (complex) toroidal coefficients being renewed at random times, following the work of Simon et al. (1995).
|
[
"Eyink & Sreenivasan 2006"
] |
[
"From Figure 2 in Rincon et al. (2017), we further notice that the vortical component of the photospheric surface velocity field has a structure reminiscent of 2D Euler turbulence, which admits a vortex point representation",
"also on the sphere"
] |
[
"Uses",
"Uses"
] |
[
[
626,
650
]
] |
[
[
402,
624
],
[
652,
670
]
] |
2016AandA...592L..11S__Vasyunin_&_Herbst_(2013)_Instance_1
|
Methanol is believed to be formed on dust grains (Watanabe & Kouchi 2002) by subsequent hydrogenation of carbon monoxide, and its detection towards prestellar cores is already a challenge for current models given the absence of efficient desorption processes in these sources. Thermal desorption is out of question because of the low dust temperature. Recent laboratory studies showed that the photo-desorption of methanol from ices is likewise negligible (Bertin et al. 2016; Cruz-Diaz et al. 2016). The main desorption products when irradiating pure and mixed methanol ices are photo-fragments of methanol. An alternative route to explain the presence of methanol in the gas phase is the reactive/chemical desorption that has been theoretically proposed by Garrod et al. (2007) and Vasyunin & Herbst (2013) and experimentally studied by Dulieu et al. (2013) and Minissale et al. (2016). On the other hand, c-C3H2 is mainly formed in the gas phase (e.g. Spezzano et al. 2013) and is expected to preferentially trace dense and chemically young gas, that is, gas where C atoms have not yet been mainly locked into CO. This gas rich in C atoms is expected in the outer envelope of an externally illuminated dense core (e.g. Aikawa et al. 2001). However, towards L1544, c-C3H2 only appears to trace one side of the core, the one closer to the sharp N(H2) edge and away from the CH3OH peak. This indicates that photo-chemistry is not uniformly active around L1544, most likely because the distribution of the envelope material (belonging to the filament within which L1544 is embedded) is not uniform, as is clearly shown by the Herschel map in Fig. 1. This figure shows that methanol traces a region farther away from the southern sharp edge of the N(H2) map, which is possibly more shielded from the ISRF and where most of the carbon is locked in CO. CH3OH is preferentially found at the northern edge of L1544 because here photochemistry does not play a major role (so C is locked in CO) and the density is low enough to maintain a higher fraction of CH3OH in the gas phase, but above the threshold value for CO freeze-out, a few ×104 cm-3 (Caselli et al. 1999). Based on the Keto & Caselli (2010) model (that was updated by Keto et al. 2014), the volume density at the distance of the methanol peak is predicted to be 8 × 104 cm-3, which is just above the threshold value. In contrast, cyclopropenylidene has the most prominent peak towards the southern sharp edge of the H2 column density and extends along the semi-major axis of the core, almost parallel to the south-west edge of the N(H2) map. This behaviour is also clearly shown in Fig. 2, where the integrated intensities of both methanol and cyclopropenylidene are plotted against N(H2). c-C3H2 is also present at values of N(H2) that are lower than those of methanol, and it maintains a flat intensity profile, suggestive of a layer-like structure, with no significant increase towards the core centre. The CH3OH intensity instead shows a sharp rise up to column densities of about 1.6 × 1022 cm-2, and it declines at higher values.
|
[
"Vasyunin & Herbst (2013)"
] |
[
"An alternative route to explain the presence of methanol in the gas phase is the reactive/chemical desorption that has been theoretically proposed by Garrod et al. (2007) and",
"and experimentally studied by Dulieu et al. (2013) and Minissale et al. (2016).",
"On the other hand, c-C3H2 is mainly formed in the gas phase",
"and is expected to preferentially trace dense and chemically young gas, that is, gas where C atoms have not yet been mainly locked into CO.",
"This gas rich in C atoms is expected in the outer envelope of an externally illuminated dense core",
"However, towards L1544, c-C3H2 only appears to trace one side of the core, the one closer to the sharp N(H2) edge and away from the CH3OH peak. This indicates that photo-chemistry is not uniformly active around L1544, most likely because the distribution of the envelope material (belonging to the filament within which L1544 is embedded) is not uniform, as is clearly shown by the Herschel map in Fig. 1."
] |
[
"Background",
"Background",
"Background",
"Background",
"Background",
"Differences"
] |
[
[
784,
808
]
] |
[
[
609,
783
],
[
809,
888
],
[
889,
948
],
[
977,
1116
],
[
1117,
1215
],
[
1243,
1648
]
] |
2016MNRAS.456..512C__Kronberg_et_al._2004_Instance_3
|
Extended radio emission in galaxies is associated with both radio jets and lobes and with outflows, seen often as aligned radio sources in the opposite directions with respect to the central compact radio core. Giant radio galaxies (GRG) are extreme cases of this phenomenology with jets and lobes extending on ∼ Mpc scales suggesting that they are either very powerful or very old site for electron acceleration. In this respect, GRGs have a crucial role in the acceleration of cosmic rays over large cosmic scales (e.g. Kronberg et al. 2004), in the feedback mechanism of AGNs into the intergalactic and intracluster medium (e.g. Subrahmanyan et al. 2008) and in the seeding of large-scale magnetic fields in the universe (e.g. Kronberg et al. 2004) and they are excellent sites to determine the total jet/lobe energetics in AGN-dominated structures (see e.g. Colafrancesco 2008, Colafrancesco & Marchegiani 2011). To date our knowledge of GRGs (see e.g. Ishwara-Chandra & Saikia 1999, 2002; Lara et al. 2001; Machalski, Jamrozy & Zola 2001; Schoenmakers et al. 2001; Kronberg et al. 2004; Saripalli et al. 2005; Malarecki et al. 2013; Butenko et al. 2014) is limited by their sparse numbers and by the difficulty of detecting them over large areas of the sky. Low-frequency radio observations have an enhanced capacity to detect the extended old electron population in these objects (see e.g. the recent Low Frequency Array – LOFAR – observation of the GRG UGC095551), but high-frequency radio observations are less efficient in this task due to the steep-spectra of giant radio lobes. In this context these sources will be ideal targets for the next coming deep, wide-field surveys like, e.g. the ATLAS survey of the Australia Telescope Network Facility (ATNF; see Norris et al. 2009) or the Square Kilometre Array (SKA) deep surveys that will have the potential to study their population evolution up to high redshifts and thus clarifying their role on the feedback for the evolution of non-thermal processes in large-scale structures.
|
[
"Kronberg et al. 2004"
] |
[
"To date our knowledge of GRGs (see e.g.",
"is limited by their sparse numbers and by the difficulty of detecting them over large areas of the sky. Low-frequency radio observations have an enhanced capacity to detect the extended old electron population in these objects (see e.g. the recent Low Frequency Array – LOFAR – observation of the GRG UGC095551), but high-frequency radio observations are less efficient in this task due to the steep-spectra of giant radio lobes.",
"In this context these sources will be ideal targets for the next coming deep, wide-field surveys like, e.g. the ATLAS survey of the Australia Telescope Network Facility (ATNF",
"or the Square Kilometre Array (SKA) deep surveys that will have the potential to study their population evolution up to high redshifts and thus clarifying their role on the feedback for the evolution of non-thermal processes in large-scale structures."
] |
[
"Background",
"Background",
"Future Work",
"Future Work"
] |
[
[
1070,
1090
]
] |
[
[
917,
956
],
[
1159,
1588
],
[
1589,
1763
],
[
1789,
2040
]
] |
2015MNRAS.448.1847H__Kim_et_al._2003_Instance_1
|
We assume that gas and dust are well mixed along each line of sight in which case the gas surface density (Σgas) is proportional to the dust surface density (Σdust) and the proportionality factor is the gas-to-dust mass ratio (rgd). The dust masses derived by SED fitting methods vary significantly and systematically (e.g. Galliano et al. 2011; Gordon et al. 2014) depending on the applied method and assumed dust optical properties. It is therefore important to use a value of rgd consistent with the used SED fitting methodology, i.e. the value of rgd that gives the correct gas mass given the derived dust mass. We have calibrated the rgd using the available gas tracers on a large scale. Within a 500 arcsec radius centred on NGC 346, we find an atomic gas mass of 2.7×106 M⊙ using the Parkes+ATCA H i map (Kim et al. 2003) and a spin temperature of 60 K (Bernard et al. 2008). The molecular gas mass determined from CO(J = 1–0) within the same aperture is 7.4×104 M⊙ when using a CO-to-H2 column density conversion factor (XCO) of 1021 [cm−2 (K km s−1)−1] (Bolatto, Wolfire & Leroy 2013). If we use an XCO of 1022 [cm−2 (K km s−1)−1], we find a molecular gas mass of 7.5×105 M⊙. Such an elevated XCO may be more appropriate given the harsh radiation field that is prevalent in the region, which may induce enhanced photodissociation of the CO molecules. Thus, we derive a total gas mass in the range of 2.7×106–3.4×106 M⊙. The dust mass derived using the dust continuum photometry in this aperture and applying the same SED fitting routine we use for the pixel-by-pixel analysis yields 2.7×103 M⊙. The derived rgd ranges from ∼1000 to ∼1250. In the following, we use a fiducial value for rgd of 1250, which is much higher than the Galactic value (100; Draine & Li 2007) due to the lower metallicity of the SMC. Our value for rgd of 1250 is lower than the total SMC integrated value of 1740 from Gordon et al. (2014). The global value includes a significant amount of atomic gas at low column density in the outskirts of the SMC which is not detected in dust emission. This drives up the global rgd.
|
[
"Kim et al. 2003"
] |
[
"Within a 500 arcsec radius centred on NGC 346, we find an atomic gas mass of 2.7×106 M⊙ using the Parkes+ATCA H i map"
] |
[
"Uses"
] |
[
[
812,
827
]
] |
[
[
693,
810
]
] |
2021MNRAS.503.5179N__Blanton_et_al._2004_Instance_2
|
Here, we report on molecular gas observations of NGC 0708, the BCG in the low-mass galaxy cluster Abell 262, itself part of the Perseus–Pisces galaxy supercluster. NGC 0708 lies 58.3 ± 5.4 Mpc away (estimated using infrared surface brightness fluctuations; Jensen et al. 2003). It is a giant elliptical galaxy with a weak dust lane (Ebneter & Balick 1985; Wegner et al. 1996) and an effective radius of 33 arcsec ($\approx \, 9.3$ kpc; Wegner et al. 2012). See Fig. 1 for an HST image of NGC 0708. Abell 262 was identified as having an X-ray emitting ICM by Jones & Forman (1984), and Stewart et al. (1984) measured the cooling time to be 1.3 × 109 yr, smaller than the age of the Universe so that the cluster is expected to form a cooling flow. The 20-cm observations of Parma et al. (1986) revealed a double-lobed, ‘S’-shaped jet and led to the classification of NGC 0708 as a weak Fanaroff–Riley Class I radio source (Blanton et al. 2004). The top panel of Fig. 1 also has 330 MHz continuum observations from Clarke et al. (2009) overlaid (blue contours) to show the shape and orientation of the large-scale jet. Analysis of Chandra observations revealed a hole or bubble within the ICM, cospatial with the eastern lobe of the jet (Blanton et al. 2004). Clarke et al. (2009) found additional 3–6 kpc radius cavities at differing position angles within the X-ray gas, and at a range of radial distances from the BCG (8–29 kpc), indicating multiple episodes of AGN activity from a precessing SMBH jet. They concluded that the total AGN emission should be capable of counteracting the cooling flow over several outbursts. Using their multifrequency observations of NGC 0708, Clarke et al. (2009) also calculated the radio spectral index (α) from 235 to 610 MHz, finding the spectrum to be flat in the core (α = −0.5), typical of new particles in a jet. They also estimated a lower limit on average outburst repetition time-scales in Abell 262 to be τrep ≥ 28 Myr.
|
[
"Blanton et al. 2004"
] |
[
"Analysis of Chandra observations revealed a hole or bubble within the ICM, cospatial with the eastern lobe of the jet"
] |
[
"Background"
] |
[
[
1235,
1254
]
] |
[
[
1116,
1233
]
] |
2020AandA...633A.163C__Aalto_et_al._2015_Instance_1
|
By using the RADEX2 dense cloud models developed by Aalto et al. (2015) to reproduce the HCN(3–2)/(1–0) line luminosity ratios in the outflow of Mrk 231, we can attempt to find a combination of XHCN, XCN, Tkin, and nH2 solutions that can also fit the CN/HCN and CN spin doublet line ratios (Table 2). We assume that the HCN and CN line emissions arise from the same dense cloud population, while the low-J CO line emission is due to a different, more diffuse phase of the outflow. We recall that in these models (see also Aalto et al. 2015), the dense clouds can be either self-gravitating virialised clouds, which implies that their internal velocity dispersion (Δvsg) is locked to their mass (Mvir) and size (R) through Δvsg = (GMvir/G)1/2, or unbound clouds, for which Δv ≫ Δvsg. We explored CN and HCN abundances in the range between 10−8 and 10−6. We find that depending on whether the clouds are self-gravitating or unbound, the models produce very different values for the absolute CN and HCN abundances, hence XCN and XHCN remain quantitatively unconstrained for the outflow with current data. However, all possible solutions that fit the observed line ratios consistently require XCN > XHCN, with a CN abundance that is at least a factor of three higher than the HCN abundance. Gas densities for this outflow phase (traced by the CN and HCN emissions) are nH2 ∼ 105 − 106 cm−3, with temperatures not much higher than Tkin ∼ 20 K. Because CN is a well-known PDR tracer (see also Sect. 1), these results strongly suggest that the whole dense cloud population in outflow is affected by UV radiation. We should mention that high CN abundances may also be due to cosmic rays (e.g. see work done on the Galactic centre by Harada et al. 2015), which are known to permeate the outflow of Mrk 231, as inferred by González-Alfonso et al. (2018) based on the high OH+ abundance. However, it is not clear whether a cosmic-ray chemistry would also explain XCN > XHCN.
|
[
"Aalto et al. (2015)"
] |
[
"By using the RADEX2 dense cloud models developed by",
"to reproduce the HCN(3–2)/(1–0) line luminosity ratios in the outflow of Mrk 231, we can attempt to find a combination of XHCN, XCN, Tkin, and nH2 solutions that can also fit the CN/HCN and CN spin doublet line ratios (Table 2)"
] |
[
"Uses",
"Uses"
] |
[
[
52,
71
]
] |
[
[
0,
51
],
[
72,
299
]
] |
2022MNRAS.511.1121M__Reig_&_Nespoli_2013_Instance_4
|
Critical luminosity (Lcrit) is the luminosity above which a state transition from subcritical to supercritical takes place. The subcritical state (LX Lcrit) is known to be the low luminosity state whereas the supercritical state is high luminosity state (LX > Lcrit) (Becker et al. 2012). The critical luminosity is crucial to determine whether the radiation pressure of the emitting plasma is capable of decelerating the accretion flow (Basko & Sunyaev 1976; Becker et al. 2012). The luminosity during the 2020 giant outburst reached a record high, which was significantly higher than the critical luminosity (Reig & Nespoli 2013). The source entered a supercritical regime from a subcritical regime during the outburst. In the supercritical regime, radiation pressure is high enough to stop the accreting matter at a distance above the neutron star, forming a radiation-dominated shock (Basko & Sunyaev 1976; Becker et al. 2012). For the subcritical regime, accreted material reaches the neutron star surface through nuclear collisions with atmospheric protons or through Coulomb collision with thermal electrons (Harding 1994). These accretion regimes can also be probed by noting changes in the cyclotron line energies, pulse profiles, and changes in the spectral shape (Parmar, White, & Stella 1989; Reig & Nespoli 2013). During the transition from the subcritical to the supercritical regime, sources show two different branches in their hardness–intensity diagram (HID) which are known as horizontal branch (HB) and diagonal branch (DB) (Reig & Nespoli 2013). The HB implies the low-luminosity state of the source, which is represented by spectral changes and high X-ray variability. The DB corresponds to the high-luminosity state that appears when the X-ray luminosity is above the critical limit. The classification HB and DB depends on HID patterns that the source follows. The HB pattern is generally observed in the subcritical regime and the DB pattern is observed in the supercritical regime (Reig & Nespoli 2013).
|
[
"Reig & Nespoli 2013"
] |
[
"The HB pattern is generally observed in the subcritical regime and the DB pattern is observed in the supercritical regime"
] |
[
"Background"
] |
[
[
2008,
2027
]
] |
[
[
1885,
2006
]
] |
2018ApJ...864...76Z__Plunkett_et_al._2013_Instance_1
|
C18O (J = 2–1) observations revealed a rotational structure in the north–south direction at the center of IRAS 4C, but whether the rotation is Keplerian is not known due to the low signal-to-noise ratio (S/N) (Tobin et al. 2015). Spitzer IRAC observations show an outflow cavity structure highlighted by scattered light and shocked emission to the east side of the central source, with the west side being much fainter (Figure 19 of Tobin et al. 2015). The east side is therefore inferred to be the blueshifted side, because the blueshifted outflow cavity tends to be brighter than the redshifted cavity in NIR and MIR as it is less extincted. Despite the outflow cavity structure seen in infrared, there was no clear evidence of a molecular outflow from mm or submm observations. Previous 12CO observations either reported no detection of outflow emission toward this source (Plunkett et al. 2013; Tobin et al. 2015) or only weak compact blueshifted emission towards east of the continuum source (Stephens et al. 2018). One possible explanation for the weak 12CO outflow emission is that the outflow lies close to the plane of sky, so that the low-velocity outflow emission is easily mixed with the emission of the ambient gas, especially for abundant species like 12CO. Indeed, the inclination of the source is estimated to be nearly edge-on (Tobin et al. 2015) to about 25° between the disk plane and the line of sight (Segura-Cox et al. 2016). On the other hand, the 13CO (J = 2–1) emission reveal a compact (≲2″) structure with the blueshifted emission slightly offset to the east of the redshifted emission, which was explained as a slow outflow (Koumpia et al. 2016). Here, we report that the outflow cavity structure is clearly detected in the CCH and CS emissions, with kinematics consistent with rotation with respect to the outflow axis. This allows us to measure the angular momentum distribution in the outflow, and further constrain its launching radii and launching mechanism.
|
[
"Plunkett et al. 2013"
] |
[
"Previous 12CO observations either reported no detection of outflow emission toward this source",
"or only weak compact blueshifted emission towards east of the continuum source"
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
877,
897
]
] |
[
[
781,
875
],
[
918,
996
]
] |
2020AandA...633A..70P__Kocevski_et_al._2011_Instance_1
|
Recent studies have used the spectral indices [OII], Hδ, and Dn4000 to probe the stellar population of galaxies at intermediate redshifts (0.5 ≲ z ≲ 1.2) because they are available in the observed optical frame. All these indicators, when combined, can be used to distinguish actively star-forming, (post-) starburst, and old or passive galaxies because they are expected to occupy different regions of the possible parameter space (e.g. Couch & Sharples 1987; Balogh et al. 1999; Poggianti et al. 1999, 2009; Fritz et al. 2014). The [OII]λ3737 emission traces on-going star formation (timescales of ∼10 Myr, e.g. Couch & Sharples 1987; Poggianti et al. 1999, 2006; Kennicutt 1998; Kewley et al. 2004), but it also depends on the metallicity and can be a poor tracer for dusty galaxies (e.g. Kewley et al. 2004; Yan et al. 2006; Kocevski et al. 2011). By measuring the [OII] equivalent width (EW), we can also crudely trace the specific SFR (sSFR), which is found to be anti-correlated with stellar mass (e.g. Bridge et al. 2015; Cava et al. 2015; Darvish et al. 2015a), with more massive star-forming galaxies having lower [OII] EWs. Additionally, higher density environments are found to depress [OII] emission (e.g. Balogh et al. 1999; Darvish et al. 2015a). The Hδ line (and other strong Balmer absorption lines) can be indicative of a post-starburst phase (≈100 − 1000 Myr after the burst, e.g. Couch & Sharples 1987; Balogh et al. 1999; Poggianti et al. 1999, 2009; Dressler et al. 2004; Vergani et al. 2010; Mansheim et al. 2017a) if a strong absorption (typical of A stars, where hydrogen absorption is strongest) is observed and no tracers of on-going star formation are found (Couch & Sharples 1987). Recently, Wu et al. (2018) found that the Hδ EW correlates with stellar mass, with more massive galaxies having weaker Hδ absorption lines, but they did not study the effect of the environment (see also e.g. Siudek et al. 2017, for a similar result on passive galaxies). Finally, a measure of the flux break at 4000 Å (D4000 and Dn4000, as defined by Bruzual 1983; Balogh et al. 1999, respectively) traces the age of the galaxy and also the stellar metallicity (especially for older systems) to a lesser degree. This break is produced by a combination of metal absorption on the atmosphere of old and cool stars and the lack of flux from young and hot OB stars (e.g. Poggianti & Barbaro 1997; Kauffmann et al. 2003), and so it is sensitive to the average age of the stellar population. The 4000 Å break is also found to be stronger for galaxies with higher stellar mass (e.g. Muzzin et al. 2012; Vergani et al. 2008; Hernán-Caballero et al. 2013; Siudek et al. 2017; Wu et al. 2018), which indicates that their stellar populations might be older, in an average sense. In terms of local density, Muzzin et al. (2012) found that galaxies in cluster environments have stronger breaks on average than their field counterparts at similar stellar masses, which the authors argued can be explained by the different fractions of star-forming and quiescent galaxies in different environments.
|
[
"Kocevski et al. 2011"
] |
[
"The [OII]λ3737 emission traces on-going star formation",
"but it also depends on the metallicity and can be a poor tracer for dusty galaxies (e.g."
] |
[
"Background",
"Background"
] |
[
[
829,
849
]
] |
[
[
530,
584
],
[
703,
791
]
] |
2020AandA...638A.113C__Juneau_et_al._2013_Instance_1
|
To date, large surveys focused on detecting AGN systems have been conducted at a range of different wavelengths and, in particular, for the IR and X-ray domains (Alexander et al. 2005, 2008; Ivison et al. 2004; Lutz et al. 2005; Menéndez-Delmestre et al. 2007, 2009; Valiante et al. 2007; Pope et al. 2008; Bonzini et al. 2013; Smolčić et al. 2017; Wang et al. 2013; Stach et al. 2019). Due to the dust extinction in the near-IR, optical, and UV, as well as gas absorption in X-ray bands, these surveys are often incomplete. In the mid- and far-IR, incompleteness of AGN surveys may arise from the fact that not all AGN have significant IR emission from a dusty torus and therefore may not be detected. It is postulated that up to a third of AGN are undetected in these surveys (Mateos et al. 2017). Moreover, other studies have compared AGN selected from various wavebands and find that their host galaxies tend to have different properties in terms of colour (Hickox et al. 2009) and star-formation rates (SFR; Juneau et al. 2013; Ellison et al. 2016). In particular, Hickox et al. (2009) illustrated that there is only very little overlap between their 122 radio-selected AGN and those selected by X-ray or IR. Therefore, dust-free radio surveys are needed to provided a more complete census of the AGN population. Traditional radio surveys are only sensitive to radio-loud (RL) AGN, which only represent a tiny fraction (10 ∼ 20%) of the whole AGN population; however, modern radio surveys can achieve a flux depth where radio-quiet AGN can be detected (see Prandoni et al. 2018 for a review). Recent work has focused on the radio as it is sensitive to AGN and star formation concordantly, thus providing a method of surveying AGN and star-formation activity across cosmic time (e.g. Smolčić et al. 2017; Padovani et al. 2015). A lot of work has also been done to look for AGN-driven radio emission, which has been identified by an excess of radio emission compared to what is expected based on the radio-FIR correlation, holding for star-forming galaxies (e.g. Ivison et al. 2010; Condon et al. 2002; Thomson et al. 2014; Magnelli et al. 2015).
|
[
"Juneau et al. 2013"
] |
[
"Moreover, other studies have compared AGN selected from various wavebands and find that their host galaxies tend to have different properties",
"and star-formation rates (SFR;"
] |
[
"Motivation",
"Motivation"
] |
[
[
1013,
1031
]
] |
[
[
800,
941
],
[
982,
1012
]
] |
2016ApJ...819...59S__Castelli_&_Kurucz_2004_Instance_1
|
HD 47366 (HIP 31674, BD−12 1566, HR 2437, TYC 5373-2001-1) is listed in the Hipparcos Catalogue (ESA 1997) as a K1 III star, with a visual magnitude of V = 6.11 and a color index of B − V = 0.994. The Hipparcos parallax π = 12.5 ± 0.42 mas (van Leeuwen 2007) corresponds to a distance of 80.0 ± 2.7 pc. The reddening E(B − V) = 0.028 was obtained from the Galactic dust map of Schlegel et al. (1998), with the correction given by Bonifacio et al. (2000) and a scaling factor of
, where d is the distance, b is the Galactic latitude, and h = 125 pc is the scale height of the reddening layer. The absolute magnitude MV = 1.51 was derived from the distance and the interstellar extinction AV = 3.1E(B − V). By adopting the broadband photometric color B − V and the estimated metallicity with the empirical calibration relation of Alonso et al. (1999, 2001), we derived the bolometric correction B.C. = −0.309 and the effective temperature Teff = 4866 ± 100 K. We used a high signal-to-noise ratio (S/N ∼ 200), iodine-free spectrum taken with HRS to measure the equivalent widths (EWs) of ∼30 Fe i lines, to derive the iron abundance [Fe/H]. The line lists as well as their oscillation strengths (
) were mainly taken from Hekker & Meléndez (2007), in which iron lines were carefully selected to avoid any blend by atomic or CN lines. The model atmosphere used in this work was interpolated from the line-blanketed, local thermodynamic equilibrium (LTE) ATLAS9-ODFNEW grid (Castelli & Kurucz 2004). The microturbulent velocity vt was obtained by minimizing the trend between the abundances of different Fe i lines and their reduced equivalent widths (
). The macroturbulent velocity was estimated with the empirical relations of vmacro versus Teff given by Hekker & Meléndez (2007), and the projected rotational velocity (
) was determined with the method of Fekel (1997). The stellar mass, surface gravity (
), radius, and age were derived using a Bayesian approach with the Geneva database (Lejeune & Schaerer 2001), which includes the post-helium flash phases for stars with M ≥ 1.7 M⊙. First, we interpolated an extensive grid of evolutionary tracks, with ΔM = 0.05 within 1.2 ≤ M/M⊙ ≤ 3.6, Δ[Fe/H] = 0.02 within −0.4 ≤ [Fe/H] ≤ +0.3, and 500 points in each track. Then, for each data point, we calculated the likelihood functions of
, Teff and [Fe/H] to match the observed values by assuming Gaussian errors. We adopted uniform prior probabilities of mass and [Fe/H]. It is noted that stars evolving more slowly have a higher probability of being observed. Without correcting this evolution effect, the resulting parameters would bias toward the rapid evolution phases. We therefore weighted the probability of each point along its evolutionary track by the normalized time-step (ai+1,j − ai,j)/(an,j − aa,j), where ai,j is the age of the ith interpolated point in the jth track, and n = 500 is the number of interpolated points in each track. Eventually, the probability distribution functions (PDFs) of the parameters yield M = 1.81 ± 0.13 M⊙, R = 7.30 ± 0.33 R⊙,
, and age = 1.61 ± 0.53 Gyr. The stellar mass is particularly important to derive the minimum masses of the orbiting planets detected with the Doppler technique. However, previous spectroscopic analyses gave discrepant results (1.87 M⊙ by Liu et al. 2010; 1.2 M⊙ by Mishenina et al. 2006) for HD 47366, which may be due to the different methods on finding Teff and
. Our determinations were based on a similar method to that of Liu et al. (2010), but used the Geneva evolutionary tracks, instead of the Y2 model (Yi et al. 2003), which does not account for the evolutionary phases after the helium-core flash. We found that the probability that the star has passed through the RGB tip and in-core helium burning phase is ∼88%. The stellar parameters of HD 47366 are listed in Table 1. In Figure 2, we plotted HD 47366 on the H-R diagram, together with the evolutionary tracks for stars with different masses and metal contents.
|
[
"Castelli & Kurucz 2004"
] |
[
"The model atmosphere used in this work was interpolated from the line-blanketed, local thermodynamic equilibrium (LTE) ATLAS9-ODFNEW grid"
] |
[
"Uses"
] |
[
[
1482,
1504
]
] |
[
[
1343,
1480
]
] |
2020MNRAS.491.3860S__Schlickeiser_2003_Instance_1
|
The main assumption is that the statistical properties of the interaction of the charged particles with the fields are dominated by Gaussian distributions. This is in correspondence with the Central Limit Theorem (CLT), which requires that all stochastic systems evolve asymptotically towards Gaussian statistics provided that (i) many interactions are involved, (ii) the change in state in individual interactions is always small, and (iii) subsequent changes of state are statistically independent of each other. Besides these necessary conditions for the applicability of the FP equation, also other simplifying assumption for a better tractability of the FP equation are being made, such as (1) the magnetic fluctuations are homogeneous in space, (2) the electromagnetic fields are quasi-static, (3) the interaction has a finite decorrelation time, etc. (see more details in the books Schlickeiser 2003; Zank 2014). Unfortunately, in astrophysical and laboratory plasmas, most of the above assumptions are not valid, yet the FP equation is used extensively, without a proof of its validity. This is especially true when the plasma particles are accelerated to high energies impulsively (e.g. in solar flares, coronal mass ejections, or the Earth’s magnetotail). The acceleration volume is finite and the expected fluctuating electromagnetic fields are strong ($\vert \delta \boldsymbol{B} \vert \ge \vert \boldsymbol{B}_0 \vert$). In solar active regions, the complex magnetic topologies host many null magnetic points which are randomly distributed inside the erupting or flaring volume (Aulanier et al. 2000; Pontin 2011). In these cases, the interaction of the particles with the strong em disturbances is transient and has no time to lead to Gaussian statistics or to become homogeneous in space (Isliker, Archontis & Vlahos 2019). Before analysing the interaction of the particles with the em fluctuations, it is important to understand the evolution of the em waves. With the use of resistive MHD codes one can show that a spectrum of high amplitude electromagnetic fluctuations evolves rapidly and leads to a fragmented current system, where reconnecting current sheets and large amplitude magnetic fluctuations are present (Arzner & Vlahos 2004; Dmitruk, Matthaeus & Seenu 2004; Vlahos, Isliker & Lepreti 2004; Isliker, Vlahos & Constantinescu 2017a), and where the various statistics clearly are non-Gaussian, following largely the paradigm of the stable Levy distributions (Isliker et al. 2017a; Isliker et al. 2019). In strongly turbulent plasmas, the magnetic fluctuations are non-collective modes and cannot be described with a simple dispersion relation $\omega =\omega (\boldsymbol{k}).$ The em environment generated from the evolution of large amplitude em fluctuations is well documented in the current literature and models much better many impulsive astrophysical and laboratory plasmas (see Dmitruk et al. 2004; Zhdankin et al. 2013; Isliker et al. 2017a and the reviews by Cargill et al. 2012; Karimabadi et al. 2014; Vlahos & Isliker 2019).
|
[
"Schlickeiser 2003"
] |
[
"Besides these necessary conditions for the applicability of the FP equation, also other simplifying assumption for a better tractability of the FP equation are being made, such as (1) the magnetic fluctuations are homogeneous in space, (2) the electromagnetic fields are quasi-static, (3) the interaction has a finite decorrelation time, etc. (see more details in the books",
"Unfortunately, in astrophysical and laboratory plasmas, most of the above assumptions are not valid, yet the FP equation is used extensively, without a proof of its validity."
] |
[
"Background",
"Background"
] |
[
[
889,
906
]
] |
[
[
515,
888
],
[
920,
1094
]
] |
2022AandA...667A..15K__Yamashiki_et_al._2019_Instance_1
|
In the search for new exoplanets, M dwarfs are ideal targets due to their high abundance in the Galaxy (Bochanski et al. 2010). However, M-type stars are prone to high levels of stellar activity (Walkowicz et al. 2011; Loyd et al. 2016, 2018b) that can impact the radial velocity and/or transit signal of such systems through phenomena such as flaring (Tofflemire et al. 2012), star spots, plages and faculae (Boisse et al. 2011; Llama & Shkolnik 2015; Cauley et al. 2018; Roettenbacher et al. 2022; Bruno et al. 2022), and other activity-induced variability (Dumusque 2018; Rackham et al. 2019; Bellotti et al. 2022; Collier Cameron et al. 2021). Aside from the observational implications, the planet’s physical and chemical state can be altered by stellar activity as well due to, for example, coronal mass ejections (CMEs) and stellar particle events (SPEs) (Yamashiki et al. 2019; Atri 2017, 2020; Segura et al. 2010), winds (Vidotto et al. 2015; Vidotto & Cleary 2020; Chebly et al. 2022; Colombo et al. 2022), and stellar flares (Segura et al. 2010; Venot et al. 2016; Chadney et al. 2017; Tilley et al. 2019; Chen et al. 2021; Louca et al. 2022), the latter being sudden releases of radiative energy triggered by magnetic reconnection (Benz & Güdel 2010). Stellar flares result in a temporary increase in incident flux on the planet’s atmosphere, which in turn increases the photochemical reaction rates that can change the chemical composition. Photochemistry, and photolysis in particular, is a key driver of chemical disequilibrium in the atmospheres of close-orbiting, gaseous planets. Photolysis does not only deplete the upper layers from species such as CH4 and NH3, but it can enrich the middle regions with haze precursors such as HCN and C2H2 as well, particularly on cooler planets (Moses et al. 2011; Venot et al. 2012; Zahnle & Marley 2014; Agúndez et al. 2014; Moses 2014; Miguel & Kaltenegger 2014; Rimmer & Helling 2016; Drummond et al. 2016; Hobbs et al. 2019; Shulyak et al. 2020; Barth et al. 2021; Baeyens et al. 2022). Stellar flares thus have the potential to alter the chemical composition and, subsequently, alter the atmosphere’s signature in transmission spectra.
|
[
"Yamashiki et al. 2019"
] |
[
"Aside from the observational implications, the planet’s physical and chemical state can be altered by stellar activity as well due to, for example, coronal mass ejections (CMEs) and stellar particle events (SPEs)"
] |
[
"Background"
] |
[
[
862,
883
]
] |
[
[
648,
860
]
] |
2020ApJ...901....8B__Müller-Mellin_et_al._1995_Instance_1
|
Three semiannual galactic hydrogen spectra as a function of energy between 40 and 250 MeV have been obtained in three different consecutive time periods (from 2018 August 6 to 2020 January 5) very much inside the heliosphere (1 au); the energy profiles are shown as black circles in Figure 6. Each measured energy spectrum is compared to the theoretical prediction from the HelMod model (Boschini et al. 2019) in the same period (blue solid curve); the maximum and minimum uncertainties related to this prediction are also reported in the plots, as dashed and dotted lines, respectively. As a further comparison, data from the SOHO/EPHIN spacecraft (red square marker) between 40 MeV and 53 MeV are also presented (Müller-Mellin et al. 1995). The agreement appears to be good in all the three examined periods, considering both statistical and systematic uncertainties. Ratio between HEPD data and models (displayed in the narrower bottom panels of Figure 6) gradually worsens at lower energies, particularly below 65 MeV, where the spectrum calculated by HEPD is systematically higher. Possible explanations for this discrepancy include a contamination from high-energy protons that is not fully removed using the simulation, and a possible contamination derived from nuclei fragmentation or from very inclined sub-cutoff protons that can enter the FoV of the instrument, even after the rigidity cutoff selection. However, although systematic uncertainties are higher than 10% in the lowest portion of the energy spectra, these results could help constrain theoretical models of particle transport from the border of the heliosphere, down to 1 au. From a comparison between the first spectrum (2018 August 6–2019 January 15) and the last one (2019 June 29–2020 January 5) an overall increase of ∼9% is observed, in very good agreement with the variation observed in SOHO/EPHIN (∼8.5%). This behavior is expected, because, as the solar activity continues to wind down (from 2018 to 2020), the effect of the Sun magnetic field diminishes, resulting in higher proton fluxes. On the other hand, HEPD data do not show a clear energy dependence in the modulation over time (typically lower energies should be more modulated with respect to higher ones); unfortunately, for HEPD the overall errors (statistical and systematic) in the first and last energy bins do not allow such a precise evaluation. Table 1 contains explicit values for the galactic hydrogen spectra in the three time periods and for each of the 16 energy bins allowed by the instrument resolution; statistical and systematic uncertainties are also reported.
|
[
"Müller-Mellin et al. 1995"
] |
[
"As a further comparison, data from the SOHO/EPHIN spacecraft (red square marker) between 40 MeV and 53 MeV are also presented",
"The agreement appears to be good in all the three examined periods, considering both statistical and systematic uncertainties."
] |
[
"Uses",
"Similarities"
] |
[
[
715,
740
]
] |
[
[
588,
713
],
[
743,
869
]
] |
2022ApJ...937...58I__Tabatabaei_et_al._2013_Instance_1
|
For JW39 and JW100 the disks’ slopes are consistent with linearity, that is the trend expected from Equation (4). This may suggests that the smoothing scale (∼10 kpc) is similar to the CRe transport scale, and, hence, the slopes resemble the standard SFR calibrators. Therefore, the spatial correlation slopes observed in the disks of JW39 and JW100 are consistent with the idea that the CRe transport scale in these galaxies is larger than observed in spiral galaxy disks (∼1–5 kpc). We investigate this result by comparing the spatial scales, L, of the two principal CRe transport processes in these systems: diffusion and advection. In the case of CRe diffusion, the spatial scale is:
5
L=4Dτ≃1.1×D1028cm2s−1τ10Myrkpc,
where τ is the timescale and D is the diffusion coefficient, which depends on the CRe energy and the local magnetic field power spectrum, and can vary between 1027 and 1029 cm2 s−1 (Strong et al. 2007). For advection, the typical spatial scale is:
6
L=V·τ≃V100kms−1τ10Myrkpc,
where V is the CRe velocity. Equation (5) shows that to reach L = 10 kpc, for a typical D = 1 × 1028 cm2 s−1, the timescale is of the order of ∼900 Myr, which is longer than the typical CRe radiative time in galactic disks. Covering these spatial scales in ≤108 yr requires a diffusion coefficient of D ≥ 9 × 1028 cm2 s−1, which is slightly higher than observed in spiral galaxies (e.g., Strong et al. 2007; Tabatabaei et al. 2013; Heesen et al. 2019). On the other hand, these scales could be consistent with the CRe advection (Equation (6)), as a typical velocity of 100 km s−1 would be able to cover 10 kpc in ∼10 Myr, which is more reasonable for low-energy electrons emitting at 144 MHz. Therefore, our results might hint that CRe transport in the disks of these jellyfish galaxies, JW39 and JW100, is either dominated by advection, due to ram pressure which is stripping the nonthermal ISM (see Section 3.1), or that the CRe diffusion might be more efficient than in normal galaxies. The diffusion coefficient D depends on the local magnetic field configuration and turbulence spectrum (Strong et al. 2007), thus it may be possible that the RPS, by affecting the ISM’s microphysics, may induce higher values of D and, hence, more efficient CRe diffusion (Equation (5)) than observed in normal spiral galaxies. However, we note that, for the rest of the sample, the spatial correlations in their stellar disks are not consistent with linearity (Figure 5). We argue that this might be due to projection effects that mix the disk and the extraplanar emissions, which, on the basis of what we observe for JW39, JO60, JW100, and JO206, follows a flat, almost uniform, distribution. Thus this blend may result in a flattening of the disks’ slopes. Another possible explanation could be a discrepancy between the sampling resolution and the transport scale that was not solved by the smoothing.
|
[
"Tabatabaei et al. 2013"
] |
[
"Covering these spatial scales in ≤108 yr requires a diffusion coefficient of D ≥ 9 × 1028 cm2 s−1, which is slightly higher than observed in spiral galaxies"
] |
[
"Differences"
] |
[
[
1414,
1436
]
] |
[
[
1230,
1386
]
] |
2019MNRAS.482.5651M__Ruiz-Lapuente_et_al._2004_Instance_1
|
To judge the origin of a star in an SNR, its kinetics characteristics may provide very important informations. Generally, except for being stripped-off a part of its envelope, the companion may receive a velocity kick from the supernova ejecta, but the kick velocity is usually much smaller than the orbital velocity (Marietta et al. 2000; Meng et al. 2007; Pakmor et al. 2008; Liu et al. 2012; Pan et al. 2012). Then, the orbital velocity of the companion at the moment of supernova explosion may represent its final space velocity to a great extant, especially for the sdB companions here, which almost do not receive any kick velocity for their large value of $A/R_{\rm 2}^{\rm SN}$. If the spatial velocity of a star in an SNR is very different from the others in the SNR, the star is very possible to be the surviving companion in the remnant (Ruiz-Lapuente et al. 2004). In Fig. 7, we present the companion orbital velocity relative to binary centroid versus the companion radius at the moment of supernova explosion. From the figure, we can see that different companions may have very different orbital velocity. For MS companions, the orbital velocity is from 150 to 200 ${\rm km\, s^{\rm -1}}$, while the RG companions have an orbital velocity of 50–110 ${\rm km\, s^{\rm -1}}$. For the sdB companions, the orbital velocity covers a large range, from 50 to 190 ${\rm km\, s^{\rm -1}}$. Such a large range is mainly derived from the large initial orbital period range for the systems producing sdB companions, i.e. log (Pi/d) is from ∼0.4 to 1.2. In other words, although the mass transfer between a binary system must begin in HG for producing a sdB companion, it may occur at the stage very close to the MS end or at the end of the HG. The upper limit of the orbital velocity here is lower than that in Meng & Podsiadlowski (2017) by 60 ${\rm km\,s^{\rm -1}}$, which originates from the fact that after MWD = 1.378 M⊙, the binary orbital period increases with mass transfer for a reversed mass ratio (see fig. 2 in Meng & Podsiadlowski 2017). For the same reason, the lower limit of the orbital velocity is also lower than that in Meng & Podsiadlowski (2017).
|
[
"Ruiz-Lapuente et al. 2004"
] |
[
"If the spatial velocity of a star in an SNR is very different from the others in the SNR, the star is very possible to be the surviving companion in the remnant"
] |
[
"Uses"
] |
[
[
849,
874
]
] |
[
[
687,
847
]
] |
2022ApJ...925..123N__Frenklach_&_Feigelson_1989_Instance_1
|
Benzene (C6H6), the simplest aromatic hydrocarbon, is a molecule that has raised great interest in the astrophysical community for almost four decades. This is mainly because C6H6 is one of the main precursors of polycyclic aromatic hydrocarbons (PAHs) reported to be present in interstellar dust particles (Leger & Puget 1984; Allamandola et al. 1989; Tielens 2013 and references therein), carbonaceous chondrites (Pering & Ponnamperuma 1971; Hayatsu et al. 1977; Hahn et al. 1988), and other astrophysical environments, such as carbon-rich, high-temperature environments (circumstellar and carbon-rich protoplanetary nebulae; Buss et al. 1993; Clemett et al. 1994). Benzene rings easily produce more complex, polycyclic structures by the one-ring build-up mechanism (Simoneit & Fetzer 1996). In space, an analogous process to carbon soot formation occurring on Earth can be initiated through the completion of that first aromatic ring and may also lead to the synthesis of PAHs (Tielens & Charnley 1997). Mechanisms involving the addition of hydrocarbons, such as acetylene onto aromatic rings as well as the attachment of other aromatic rings, or hydrocarbon pyrolysis, have been proposed to characterize the growth process of PAHs (Bittner & Howard 1981; Frenklach & Feigelson 1989; Wang & Frenklach 1997; Cherchneff 2011 and references therein). PAH synthesis from shocked benzene has also been reported (Mimura 1995). PAHs are crucial materials involved in a variety of cosmochemical processes. For example, amino acids could be synthesized by aqueous alteration of precursor PAHs in carbonaceous chondrites (Shock & Schulte 1990). PAHs are also produced in laboratory-simulated planetary atmospheres of Titan and Jupiter (Sagan et al. 1993; Khare et al. 2002; Trainer et al. 2004), and results from these studies indicate that the formation of aromatic rings and polyaromatics may be, among other sources, a possible chemical pathway for the production of the atmospheric solid particles (Lebonnois et al. 2002; Wilson et al. 2003; Trainer et al. 2004). The formation and evolution of benzene in planetary environments or other solar system objects thus represents a fundamental primary stage of the PAH production and other subsequent relevant chemical and prebiotic processes (like soot formation). In this context, several works related to benzene have been devoted to better understand the physico-chemical processes of irradiated C6H6, in its gaseous and solid phases, and the derived products, by acquiring high-resolution astronomical spectra, carrying out detailed laboratory studies or developing theoretical modeling (Allamandola et al. 1989 and references therein; Callahan et al. 2013; Materese et al. 2015; Mouzay et al. 2021). Laboratory astrophysical investigations have mostly focused on performing vibrational spectroscopy of ion, electron, or UV irradiated C6H6 gas and C6H6 ice. Such investigations aim to provide data on the spectral properties of the irradiated C6H6 materials, compare them with spectra obtained from astronomical observations (e.g., observations of the interstellar medium), or to study photoprocessed benzene ices to understand the fate of benzene ices in Titan’s stratosphere and help understanding the formation of aerosol analogs observed in Saturn’s moon’s stratosphere (Mouzay et al. 2021).
|
[
"Frenklach & Feigelson 1989"
] |
[
"Mechanisms involving the addition of hydrocarbons, such as acetylene onto aromatic rings as well as the attachment of other aromatic rings, or hydrocarbon pyrolysis, have been proposed to characterize the growth process of PAHs"
] |
[
"Background"
] |
[
[
1259,
1285
]
] |
[
[
1007,
1234
]
] |
2022MNRAS.509.6091H___2020a_Instance_1
|
Galactic winds have been ubiquitously observed in galaxies at both low and high redshifts, and they are critical to galaxy formation and evolution. Simulations calibrated to match these observations predict that a large amount of galactic material is ejected as a wind before reaccreting to either form stars or be ejected once again (Oppenheimer et al. 2010; Anglés-Alcázar et al. 2017). Current cosmological hydrodynamic simulations of galaxy formation employ a variety of subgrid models (e.g. Springel & Hernquist 2003; Oppenheimer & Davé 2006; Stinson et al. 2006; Dalla Vecchia & Schaye 2008; Agertz et al. 2013; Schaye et al. 2015; Davé, Thompson & Hopkins 2016; Tremmel et al. 2017; Pillepich et al. 2018; Davé et al. 2019; Huang et al. 2020a) that artificially launch galactic winds, but the results are sensitive to numerical resolution and the exact subgrid model employed (Huang et al. 2019, 2020a). Simulations without these subgrid wind models (e.g. Hopkins et al. 2018; Kim & Ostriker 2015; Martizzi et al. 2016) allow winds to occur ‘naturally’, but these simulations may not resolve the scales necessary to resolve the important known physical processes (Scannapieco & Brüggen 2015; Brüggen & Scannapieco 2016; Schneider & Robertson 2017; McCourt et al. 2018; Huang et al. 2020b). Hence, modelling galactic winds accurately remains a theoretical challenge for even the most refined high-resolution simulations of galaxies (see Naab & Ostriker 2017, for a review). Even if one were able to accurately model the formation of galactic winds, the subsequent propagation in galactic haloes depends on a complicated interplay of many physical processes that occur on a wide range of physical scales that cannot be simultaneously resolved in a single simulation. For example, to robustly model the propagation and disintegration of moving clouds in various situations requires cloud-crushing simulations with at least sub-parsec scale resolution (Schneider & Robertson 2017; McCourt et al. 2018), which is orders of magnitudes below the resolution limits of cosmological simulations. Furthermore, most cosmological hydrodynamic simulations concentrate their resolution in the dense, star-forming regions of galaxies and thus have lower resolution in the circumgalactic medium (CGM, but see Hummels et al. 2019; Mandelker et al. 2019; Peeples et al. 2019; Suresh et al. 2019; van de Voort et al. 2019). To date, cosmological simulations do not include physically motivated subgrid models for galactic wind evolution, which are required to capture these small-scale physical processes.
|
[
"Huang et al. 2020a"
] |
[
"Current cosmological hydrodynamic simulations of galaxy formation employ a variety of subgrid models (e.g.",
"that artificially launch galactic winds,"
] |
[
"Background",
"Background"
] |
[
[
731,
749
]
] |
[
[
389,
495
],
[
751,
791
]
] |
2021MNRAS.503..354G__Cantat-Gaudin_et_al._2020_Instance_2
|
The spatial distribution of OB stars and associations, young long-period Cepheids and open clusters, star-forming regions, H ii regions, interstellar dust, and giant molecular and neutral gas clouds in the solar vicinity that have been in existence generally τ ≲ 108 yr is known to correlate with the location of the inner Sagittarius, the closest Orion, and outer Perseus spiral arm segments. (The distances for the vast majority of these spiral tracers have been determined in the literature with trigonometric or photometric methods.) The Sun is situated at the inner edge of the Orion arm (Levine et al. 2006; Hou & Han 2014; Nakanishi & Sofue 2016; Xu et al. 2018, 2021; Lallement et al. 2019; Reid et al. 2019; Skowron et al. 2019; Cantat-Gaudin et al. 2020; Fig. 2 above).3 These three spatial features nearby to the Sun appear to form part of the global spiral structure in the Galaxy. Contrary, the objects of older population with larger random velocities, for instance, main-sequence A–K stars or the oldest Cepheids and open clusters, do not currently follow the exact location of those arms (e.g. Cantat-Gaudin et al. 2020, fig. 8 therein; Griv et al. 2020, fig. 7 therein). The latter can be explained by the difference in rotation velocity between the spiral density waves and the objects. Investigating the velocity field of Xu et al.’s (2018) O and early B-type stars in the framework of the Lin–Shu density-wave proposal, we also found that the Sun lies within the Orion arm, at the inner edge of this spiral. The radial distance from the Sun to the centre of the Orion arm is ≈0.2 kpc in the direction of the Galactic anticentre, the centre of the Sagittarius arm is ≈1.8 kpc from the Sun in the direction of the GC, and the width of the arms is ≈0.5 kpc. The radial distance between the centres of the Orion and Sagittarius arms near the Sun is λrad ≈ 2 kpc (cf. Hou & Han 2014; Wu et al. 2014; Bovy et al. 2015). As for us, the nearest Orion spiral arm forms part of the dominant density-wave structure of the system.
|
[
"Cantat-Gaudin et al. 2020"
] |
[
"These three spatial features nearby to the Sun appear to form part of the global spiral structure in the Galaxy. Contrary, the objects of older population with larger random velocities, for instance, main-sequence A–K stars or the oldest Cepheids and open clusters, do not currently follow the exact location of those arms (e.g.",
"fig. 8 therein"
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
1110,
1135
]
] |
[
[
781,
1109
],
[
1137,
1151
]
] |
2017AandA...601A..87C__Falcke_(1996)_Instance_1
|
In a quasi-isothermal jet, Uj is (17)\begin{equation} \label{eq:U_j_quasi} U_{\rm j} = \zeta n_0 m_{\rm p} c^2\left(\frac{\gamma_{\rm j}\beta_{\rm j}}{\gamma_0\beta_0}\right)^{-\Gamma}\left(\frac{z}{z_0}\right)^{-2} \cdot \end{equation}Uj=ζn0mpc2γjβjγ0β0−Γzz0-2·Substituting Eqs. (17) and (13) into Eq. (10), and assuming the jet is launched with an initial γ0β0 equal to the sound speed (Eq. (16)), the 1D Euler equation that results is \begin{eqnarray} \label{eq:AGNJET_Corrected} &&\left\{\gamma_{\rm j}\beta_{\rm j}\frac{\Gamma+\xi}{\Gamma-1}-\Gamma\gamma_{\rm j}\beta_{\rm j}-\frac{\Gamma}{\gamma_{\rm j}\beta_{\rm j}}\right\}\frac{\partial \gamma_{\rm j}\beta_{\rm j}}{\partial z} = \frac{2}{z}; \\ &&\xi = \frac{1}{\zeta}\left(\frac{\gamma_{\rm j}\beta_{\rm j}}{\gamma_0\beta_0}\right)^{\Gamma-1}; \qquad \gamma_0\beta_0=\sqrt{\frac{\zeta\Gamma(\Gamma-1)}{1+2\zeta\Gamma-\zeta\Gamma^2}} \cdot \end{eqnarray}γjβjΓ+ξΓ−1−Γγjβj−Γγjβj∂γjβj∂z=2z;ξ=1ζγjβjγ0β0Γ−1; γ0β0=ζΓ(Γ−1)1+2ζΓ−ζΓ2·The above equation should reduce to the jet Lorentz factor profile used in Falcke (1996), Markoff et al. (2005) when ζ = 1. However, it differs from Eq. (2) in Falcke (1996): (20)\begin{equation} \label{eq:Heino96} \left\{\gamma_{\rm j}\beta_{\rm j}\frac{\Gamma+\xi}{\Gamma-1}-\frac{\Gamma}{\gamma_{\rm j}\beta_{\rm j}}\right\}\frac{\partial \gamma_{\rm j}\beta_{\rm j}}{\partial z} = \frac{2}{z}; \end{equation}γjβjΓ+ξΓ−1−Γγjβj∂γjβj∂z=2z;(21)\begin{equation} \xi = \left(\gamma_{\rm j}\beta_{\rm j}\frac{\Gamma+1}{\Gamma(\Gamma-1)}\right)^{1-\Gamma} \cdot \end{equation}ξ=γjβjΓ+1Γ(Γ−1)1−Γ·The difference between our equation and the equation in Falcke (1996) can be accounted for as follows: the − Γγjβj term in Eq. (18) results from a neglected \hbox{$\frac{\partial}{\partial z}(U_{\rm j}/n)$}∂∂z(Uj/n) term, the difference in the exponent in ξ results from an arithmetic error, and finally the difference in the inside of the parenthesis of ξ terms is from setting \hbox{$\gamma_0\beta_0 = \beta_{\rm s0}^{2}$}γ0β0=βs02 instead of using the proper value given in Eq. (16). The difference between the solutions of Eqs. (18) and (20) are small and shown in Fig. 1. In Fig. 1, we also include solutions to the 1D Euler equations when the jet is isothermal (Tj = const., i.e., Eq. 20 with ξ = 1) and adiabatic (Tj ∝ (γjβj)1 − Γz2 − 2Γ, see Eq. (25)).
|
[
"Falcke (1996)"
] |
[
"The above equation should reduce to the jet Lorentz factor profile used in",
"when ζ = 1."
] |
[
"Similarities",
"Similarities"
] |
[
[
1077,
1090
]
] |
[
[
1002,
1076
],
[
1114,
1125
]
] |
2019ApJ...884..132K__Tanihata_et_al._2003_Instance_2
|
First, we discuss the discrepancy of the distribution scale of the radio core positions based on the discussions of the internal shock model (Koyama et al. 2015; Niinuma et al. 2015). As is discussed there, the radio cores in Mrk 501 and Mrk 421 observed at 43 or 22 GHz can usually be considered as the internal shocked regions owing to the convex shape of the radio spectrum peaking around 10 GHz (Giroletti et al. 2008; Sokolovsky et al. 2010; Lico et al. 2012; Blasi et al. 2013). The standard internal shock model of blazars considers that the discrete ejecta with higher speeds (with Lorentz factor Γf) catch up with the preceding slower ejecta (with Lorentz factor Γs), and the collision leads to the nonthermal emission (e.g., Spada et al. 2001; Tanihata et al. 2003; Guetta et al. 2004; Kino et al. 2004; Ghisellini et al. 2005). Based on the model, the distribution scale of the internal shocks (ΔDIS in Figure 7, defined as the difference between the largest distance between the internal shock and the central engine DIS,max and the closest one DIS,min) can be explained as the variation of the Lorentz factors of the ejecta (Koyama et al. 2015; Niinuma et al. 2015), by assuming the Lorentz factor ratio (Γf/Γs) and the initial separation of the ejecta (IIS). The core stable within 200 μas constrained by the VERA can be explained by Lorentz factors within a factor of two variation for the slower ejecta, i.e., 8 ≤ Γs ≤ 17, by assuming a minimum value of 8 (e.g., Kino et al. 2002), Γf/Γs ≤ 1.01 (Tanihata et al. 2003), and IIS ∼ 1 Rs (Koyama et al. 2015). This time we refined the distribution scale of the radio core within 42 μas along its main jet axis, or 4.6 × 103 Rs deprojected (see Figure 7). Based on the same assumptions as in Koyama et al. (2015), to explain the further stable distribution scale of the internal shocks, the variation of Lorentz factors of the slower ejecta is constrained to be much smaller, within 30% or 8 ≤ Γs ≤ 10. On the other hand, the radio core wandering of ΔDIS ∼ 2.6 × 105 Rs in Mrk 421 can be explained by the maximum value as Γs ∼ 60 (with different assumptions; Niinuma et al. 2015). Even by applying the same assumptions to Mrk 421 as those for Mrk 501, the maximum of the slower Lorentz factor is estimated to be Γs ∼ 50, which is still a few times as large as that of Mrk 501. Therefore, even during the X-ray and VHE γ-ray active states in 2012, the maximum Lorentz factors that explain the stability of Mrk 501's core are roughly a few times smaller than those for Mrk 421's wandering core, based on the internal shock model.
|
[
"Tanihata et al. 2003"
] |
[
"The core stable within 200 μas constrained by the VERA can be explained by Lorentz factors within a factor of two variation for the slower ejecta, i.e., 8 ≤ Γs ≤ 17, by assuming",
"Γf/Γs ≤ 1.01"
] |
[
"Background",
"Background"
] |
[
[
1512,
1532
]
] |
[
[
1273,
1450
],
[
1498,
1510
]
] |
2019ApJ...886...34F__Sahijpal_&_Goswami_1998_Instance_2
|
If the variation in 10Be/9Be ratios of CAIs reflects those episodic accretion events, 10Be/9Be ratios of CH–CB CAIs observed in this study would give important constraints on the evolution of the solar protoplanetary disk. Astronomical observations suggest that FUori-type outbursts are confined to the first few hundreds of thousands of years, which correspond to the class I stage of the protoplanetary disk evolution (e.g., Schulz 2012). We propose that the high and variable 10Be/9Be ratios recorded in CH–CB CAIs reflect episodic cosmic-ray fluxes caused by FUori-type outbursts. On the other hand, relatively low and less variable 10Be/9Be ratios recorded in CV CAIs may reflect less intensive episodic accretion events, possibly the EXori-type outbursts, which are confined to the evolutional stage of a few million years after the formation of the protoplanetary disk (=class II). Note that CH–CB CAIs studied here show no (or very low) signs of 26Al-derived 26Mg excesses, while most CV CAIs show clear evidence for the past presence of 26Al. If 26Al was introduced into the solar system at the earliest stage of the disk evolution (e.g., Sahijpal & Goswami 1998), differences in Be–B and Al–Mg systematics between CH–CB and CV CAIs imply that the injection of 26Al have occurred between the evolutionary stages class I and class II of the solar protoplanetary disk. This scenario is in agreement with arguments by other authors that the 26Al-free CAIs formed prior to injection and homogenization of 26Al in the early solar system (Sahijpal & Goswami 1998; Sahijpal et al. 2000; Krot et al. 2008a see more discussion in Krot et al. 2012a). Importantly, as mentioned in the introduction, CH–CB chondrites may have accreted a significant amount of outer solar system materials (Murty et al. 2007; Ivanova et al. 2008; Briani et al. 2009; Bonal et al. 2010; Olsen et al. 2016; Van Kooten et al. 2016), suggesting that CH–CB chondrites formed at outer parts of the solar protoplanetary disk relative to CV chondrites. In this case, our new Be–B and Al–Mg data set implies that the earliest formed CAIs tend to be transported into the outer part of the solar protoplanetary disk, where the parent bodies of CH–CB chondrites likely accreted. Yang & Ciesla (2012) modeled the evolution of the protoplanetary disk and material transport in the protoplanetary disk. Interestingly, Yang & Ciesla (2012) showed that outward radial transport in class I would have been greater than that of later stages of YSO evolution, suggesting that the earliest formed CAIs could be preserved in primitive bodies that accreted in the outer part of the disk. This model is consistent with our interpretation for the Be–B and Al–Mg systematics on CH–CB CAIs. It should be noted, however, that it is possible that 26Al were heterogeneously distributed in the CAI-forming regions at the earliest stage of the solar system evolution (e.g., Krot et al. 2008a; Holst et al. 2013; Park et al. 2017 and reference therein). Because no Pb–Pb ages of CH–CB CAIs are available at present, we cannot discard that possibility. Very recently, Kööp et al. (2018) found helium and neon excesses in the 26Al-free hibonite-rich CAIs, which can be attributed to in situ irradiation by energetic particles. Because 26Al-rich CAIs in CV chondrites lack comparable noble gas irradiation records (Vogel et al. 2004), Kööp et al. (2018) concluded that 26Al-free hibonite-rich CAIs experienced intense energetic particle irradiation at the earliest stage of solar protoplanetary disk evolution. This conclusion seems to be consistent with our above scenario for 26Al-free CH–CB CAIs. Note, however, that 10Be/9Be ratios of 26Al-free hibonite-rich CAIs in CM chondrites tend to be in the range of those for 26Al-rich CV CAIs (Liu et al. 2009, 2010), which is inconsistent with the above scenario. Therefore, the relationship between 10Be and 26Al in the early solar system would be more complicated than we thought.
|
[
"Sahijpal & Goswami 1998"
] |
[
"This scenario is in agreement with arguments by other authors that the 26Al-free CAIs formed prior to injection and homogenization of 26Al in the early solar system"
] |
[
"Similarities"
] |
[
[
1542,
1565
]
] |
[
[
1376,
1540
]
] |
2021MNRAS.501.2934C__Pinte_et_al._2019_Instance_1
|
Understanding how the diverse populations of protoplanetary discs in young stellar regions results in the range of exoplanet types and architectures found in the Galaxy is one of the major goals of planet-formation theory. This is an extremely challenging task due in part to the limited observational constraints available. The Atacama Large Millimetre/submillimetre Array (ALMA) is providing truly transformational images of protoplanetary discs with unprecedented sensitivity and resolution (Andrews 2020). However, millimetre wavelength images reveal the locations of small dust grains but provide little information on the presence of larger particles, beyond centimetre scales. Gas giant planets are mostly made of hydrogen and helium, which ALMA cannot directly observe; therefore, the information on the gas content relies on the observations of less abundant molecules, such as CO and its isotopologues, that are subjected to uncertain depletion processes in the gas-phase (e.g. Miotello et al. 2016). Planets might be detectable by ALMA, although indirectly, by the effects they have on the gas and/or dust in the disc. When planets become massive enough, they can carve gaps (e.g. Rice et al. 2006; Pinilla, Benisty & Birnstiel 2012; Zhu et al. 2012) and disturb the dynamics of the gas (Teague et al. 2018; Casassus & Pérez 2019; Pinte et al. 2019). The minimum gap-opening mass depends on the viscosity and scale-height of the disc (Crida, Morbidelli & Masset 2006; Duffell & MacFadyen 2013), but mini-Neptune-mass (Pérez et al. 2019) or even Earth-mass planets (Rosotti et al. 2016; Dong & Fung 2017) could produce detectable gaps. Gaps consistent with fully formed planets have been imaged by ALMA in discs with estimated ages ranging from 1 Myr (HL Tau and Elias 2–24; ALMA Partnership et al. 2015; Cieza et al. 2017) to ∼10 Myr (TW Hydra; Andrews et al. 2016). However, the origin of these gaps still remains to be established and several alternative explanations have been proposed, including the effect of snow-lines on the dust/gas evolution of different volatiles (Zhang, Blake & Bergin 2015), magneto-hydrodynamic effects (Flock et al. 2015), secular gravitational instability (e.g. Youdin 2011; Takahashi & Inutsuka 2014), and viscous ring-instabilities (Dullemond & Penzlin 2018). Each one of the proposed mechanisms has their merits and shortcomings, and it is possible that different mechanisms operate together or dominate in different objects or in the same object at at different times. For a recent review on disc (sub)structures, see Andrews (2020). Substructures are also expected to be ubiquitous in protoplanetary discs from a theoretical point of view. Without substructures to halt the migration of mm-size grains at large radii, dust particles should migrate towards the innermost part of the disc in time-scales shorter than 0.1 Myr (e.g. Brauer et al. 2007), which is inconsistent with the observations showing significant mm emission at large radii (≳10 au) at much older ages. Understanding the origin and evolution of substructures in protoplanetary discs and their implications for planet formation is currently one the main challenges in the field. To better understand the incidence and properties of disc substructures in any given molecular cloud, here we present 1.3 mm/230 GHz continuum ALMA long-baseline observations at 3–5 au resolution of the 10 brightest targets of the ‘Ophiuchus DIsc Survey Employing ALMA’ (ODISEA) project (Cieza et al. 2019) that were not included in ‘The disc Substructures at High Angular Resolution Project’ (DSHARP) ALMA Cycle-4 Large Program (Andrews et al. 2018). Our new observations result in the largest sample of disc images at ∼3–5 au resolution in any star-forming region observed so far at mm wavelengths (15 objects when combined with the brightest Ophiuchus objects in DSHARP). In Section 2, we discuss the sample selection, the long-baseline observations, and the data reduction. In Section 3, we characterize the observed substructures, including gaps, rings, inner discs, and cavities. In Section 4, we discuss individual objects and use the full sample of 15 bright Ophiuchus discs observed at high-resolution to construct a tentative evolutionary sequence in which the observed substructures are mostly driven by dust evolution and the formation of giant planets. We also discuss possible connections between the substructures observed in primordial discs and those seen in more evolved debris disc systems. A summary of our results and conclusions is presented in Section 5.
|
[
"Pinte et al. 2019"
] |
[
"When planets become massive enough, they can carve gap",
"and disturb the dynamics of the gas"
] |
[
"Background",
"Background"
] |
[
[
1342,
1359
]
] |
[
[
1130,
1184
],
[
1262,
1297
]
] |
2016ApJ...829...29T__Haiman_et_al._2000_Instance_1
|
Among various DM candidates, the most popular candidate is the weakly interacting massive particles (WIMPs; like the neutralino), which have mass in the GeV range (Jungman et al. 1996; Bertone et al. 2005; Hooper & Profumo 2007; Feng 2010). The WIMPs are non-relativistic at the epoch of decoupling from the interacting particles and have negligible free-streaming velocities. Therefore, they are “cold,” called cold dark matter (CDM). In the CDM scenario, “halos” formed in small clumps, and then merged together into larger and massive objects. Galaxies formed in these halos because of the cooling of atomic hydrogen (H; Tegmark et al. 1997) or molecular hydrogen (H2, Ciardi et al. 2000; Haiman et al. 2000). On large cosmological scales (from the range
down to
), the CDM paradigm has had great success in explaining the observed universe and reproducing the luminous structures (Fixsen et al. 1996; Borgani & Guzzo 2001; Lange et al. 2001; Cole et al. 2005; Tegmark et al. 2006; Benson 2010; Hinshaw et al. 2013; Slosar et al. 2013; Wang 2013; Planck Collaboration et al. 2014; Wei et al. 2016). However, on small scales (
), there are still some discrepancies between the CDM paradigm and observations. (a) The core–cusp problem (Navarro et al. 1997; Subramanian et al. 2000). CDM simulations predict a cusp–core DM halo, whereas the observations find them cored (Salucci et al. 2012). (b) The too big to fail problem (Boylan-Kolchin et al. 2012). CDM simulations predict a central DM density significantly higher than observation allows. (c) The “missing satellite problem.” N-body simulations based on the CDM paradigm predict a number of subhalos larger than that of satellites found in our Galaxy (Klypin et al. 1999; Moore et al. 1999; Papastergis et al. 2011). Many methods have been proposed to solve these small-scale problems, such as modifying the nature of the DM from the CDM paradigm (Hu et al. 2000; Spergel & Steinhardt 2000; Su & Chen 2011; Menci et al. 2012), adding supernova feedback effect in simulation (Weinberg & Katz 2002; Mashchenko et al. 2006; Governato et al. 2010; Pontzen & Governato 2014), and considering the interplay between DM and baryons during the formation of the galaxy (El-Zant et al. 2001; Tonini et al. 2006; Pontzen & Governato 2014). However, these methods are insufficient to solve all of the above problems.
|
[
"Haiman et al. 2000"
] |
[
"Galaxies formed in these halos because of the cooling of",
"or molecular hydrogen"
] |
[
"Background",
"Background"
] |
[
[
693,
711
]
] |
[
[
548,
604
],
[
646,
667
]
] |
2015ApJ...798...95B___2010_Instance_1
|
The gravitational microlensing of lensed quasars has proven to be an effective tool for measuring the properties of quasar accretion disks, and is starting to become useful for studying the X-ray corona as well. The time-dependent microlensing magnification (or demagnification) of one or more images of a lensed quasar is moderated by the finite size of the source, which smooths the complicated caustic pattern of microlensing magnifications as the quasar passes over it. This allows us to use the microlensing magnifications to estimate the source size, and such work has shown that in general the accretion disks are larger than would be expected from either thin disk modeling or total flux arguments (Pooley et al. 2007; Anguita et al. 2008; Morgan et al. 2010; Hainline et al. 2012; Jiménez-Vicente et al. 2012). Since the effective temperature of the disk depends on radius, its apparent size depends on wavelength, leading to a chromatic dependence of the microlensing magnification, with the same quasar image experiencing larger variability at blue wavelengths than at red wavelengths. Several studies have used this to constrain the power-law slope of the size-wavelength relation, and the results have been consistent with each other and with the thin disk prediction that the size goes like the four-thirds power of the wavelength, mostly because of their large uncertainties (Poindexter et al. 2008; Bate et al. 2008; Eigenbrod et al. 2008; Floyd et al. 2009; Blackburne et al. 2011; Mosquera et al. 2011). Finally, efforts to put upper limits on the size of the X-ray regions have also been successful (Pooley et al. 2006, 2007; Chartas et al. 2009; Dai et al. 2010), and recently there have been attempts to constrain the direct and reflected components' sizes separately using color cuts or spectral decomposition (Chen et al. 2011; Blackburne et al. 2014; Morgan et al. 2012; Chen et al. 2012; Chartas et al. 2012; Mosquera et al. 2013).
|
[
"Morgan et al. 2010"
] |
[
"This allows us to use the microlensing magnifications to estimate the source size, and such work has shown that in general the accretion disks are larger than would be expected from either thin disk modeling or total flux arguments"
] |
[
"Background"
] |
[
[
750,
769
]
] |
[
[
474,
705
]
] |
2018MNRAS.473.2020L__Preibisch_et_al._1998_Instance_1
|
Upper Scorpius (hereafter UpSco) is part of the nearest OB association to the Sun, Scorpius Centaurus. The region is nearby, with a distance of ∼145 pc from Hipparcos (de Bruijne et al. 1997) and a recent update from Gaia (144.2 ± 17.6 pc: Fang, Herczeg & Rizzuto 2017). UpSco is young, with different age determinations and a possible spread among its members (Preibisch & Zinnecker 1999; Preibisch, Guenther & Zinnecker 2001; Pecaut, Mamajek & Bubar 2012; Song, Zuckerman & Bessell 2012; Pecaut 2016; Rizzuto et al. 2016). The members of UpSco exhibit a significant mean proper motion (μαcos δ = −10.5 and μδ = −23.2 mas yr−1, with a standard dispersion of about 6.4 mas yr−1: de Bruijne et al. 1997; de Zeeuw et al. 1999; Fang et al. 2017). The bright end of the UpSco population has been examined in X-rays (Walter et al. 1994; Kunkel 1999; Preibisch et al. 1998), astrometrically (de Bruijne et al. 1997; de Zeeuw et al. 1999) and spectroscopically (Preibisch & Zinnecker 2002). The low-mass and substellar population has been investigated over the past decade in great detail with the advent of modern detectors, permitting wide and/or deep surveys (Ardila, Martín & Basri 2000; Martín, Delfosse & Guieu 2004; Slesnick, Carpenter & Hillenbrand 2006; Lodieu, Hambly & Jameson 2006; Lodieu et al. 2007, 2011b; Kraus et al. 2008; Béjar et al. 2008; Slesnick, Hillenbrand & Carpenter 2008; Lafrenière, Jayawardhana & van Kerkwijk 2010; Dawson, Scholz & Ray 2011; Lafrenière et al. 2011, 2014; Dawson et al. 2013, 2014; Lodieu 2013; Peña Ramírez, Béjar & Zapatero Osorio 2016; Best et al. 2017). The first transiting systems in the association have been announced in the last years, thanks to the Kepler K2 mission (Borucki et al. 2010; Lissauer, Dawson & Tremaine 2014; Batalha 2014), including a triple system composed of an F star and two solar-type stars (Alonso et al. 2015) and several M dwarf eclipsing binaries (Kraus et al. 2015; Lodieu et al. 2015; David et al. 2016). These systems are of prime importance, because they provide the first mass and radius measurements independent of models in this region. Finally, we should note the existence of a transiting Neptune-size planet candidate around an M3 member of UpSco (Mann et al. 2016; David et al. 2016).
|
[
"Preibisch et al. 1998"
] |
[
"The bright end of the UpSco population has been examined in X-rays"
] |
[
"Background"
] |
[
[
845,
866
]
] |
[
[
744,
810
]
] |
2021AandA...651A..87O__Brunthaler_et_al._2021_Instance_1
|
To complement our study, we also analyzed GLOSTAR continuum images toward sites with maser emission. A full description of the GLOSTAR continuum data calibration and imaging is given in Brunthaler et al. (2021), while the full analysis of continuum images of Cygnus X will be presented in a forthcoming paper. Here, we briefly discuss the imaging strategy. The calibration and imaging of the continuum data was performed with the Obit package (Cotton 2008). The 2 GHz bandwidth was first rearranged into nine frequency subbands, which were used to image each pointing individually. Then, for each frequency subband the pointings were combined into large individual mosaics to cover the entire observed area. Finally, we combined the different frequencies to obtain the image at the reference frequency, which has circular beams of 19″ and 1.5″ in the D and Bconfiguration, respectively. Continuum and methanol line maps from Effelsberg observations have also been obtained as part of the GLOSTAR survey (Brunthaler et al. 2021, Rugel et al., in prep.) We note that continuum images were constructed for Effelsberg data, the VLA D configuration, the VLA B configuration, a combination of the VLA D and B (D+B) configurations, and a combination of the VLA D configuration and Effelsberg observations. The central frequency of these images is 5.8 GHz. Here, we only use B-configuration continuum maps to study the region of the investigated methanol maser positions and D+B maps of the region around DR21 (see Sect. 4.5). Methanol line data from Effelsberg were also inspected to look for flux variations in the VLA-detected masers (Sect. 4.4). The noise in the continuum images is not uniform, but rather varies across the mapped region, and can be high around strong sources with complex or extended emission. We locally measured the noise in regions close to the maser locations, resulting in 1σ values in the range from 0.056 to 0.43 mJy beam−1 for B-configuration images. For the D configuration, the 1σ rms noise ranges from 0.10 to 2.6 mJy beam−1. The higher values measured in D-configuration data are due to bright extended emission, which is present across the Cygnus X region, and are resolved out by the array in the B configuration. The highest local rms noise occurs around the strong radio source, DR21, a compact HII region.
|
[
"Brunthaler et al. (2021)"
] |
[
"A full description of the GLOSTAR continuum data calibration and imaging is given in"
] |
[
"Uses"
] |
[
[
186,
210
]
] |
[
[
101,
185
]
] |
2015ApJ...804..101Y__Hawley_et_al._1995_Instance_1
|
It is interesting to note that for wind originating within
, their value of Be becomes almost constant when
. This corresponds to the slight decrease of the poloidal velocity beyond
shown in the left panel of Figure 8. The reason why Be does not change beyond
is because in this region turbulence has not well developed within our simulation time. Note that this radius is different from the inflow equilibrium radius, which is
. Within
, everything, especially the radial density profile, is fully reliable. Beyond this radius, the density profile is not reliable, but other properties, such as the level of turbulence and subsequently outflow properties, are still reliable up to a much larger “turbulence radius,” the limiting radius of turbulence steady state. This radius can be estimated as follows. Turbulence in accretion flow is because of MRI. The fastest growth rate of MRI at radius r is
. More precisely, it takes 3–4 orbits for MRI to develop and ∼10 orbits to saturate (Hawley et al. 1995). For our simulation time of
, taking a timescale of 3 orbits, we can obtain that the “turbulence radius” is
. This is close to the value of 800 rg mentioned above. Another way to understand the “turbulence radius” is as follows. For a geometrically thick disk, the largest turbulence eddies have size of order r. The corresponding eddy turnover time is
, where
is the rms turbulent velocity. Our simulation data show
. If at a certain radius the eddy turnover time is substantially smaller than the duration of the simulation, then the local turbulence is likely to have reached quasi-steady state. Therefore, the “turbulence radius” should be some fraction of
, which gives a similar result to the above estimation. Therefore, we think that the results beyond
are not reliable. It is very likely that Be will keep changing and the poloidal velocity still remains constant beyond
. This implies that wind can at least escape beyond the outer boundary of accretion flows. Simulations with longer run times can check this point.
|
[
"Hawley et al. 1995"
] |
[
"More precisely, it takes 3–4 orbits for MRI to develop and ∼10 orbits to saturate"
] |
[
"Uses"
] |
[
[
1029,
1047
]
] |
[
[
946,
1027
]
] |
2017AandA...601A.134M__Fung_&_Dong_(2015)_Instance_2
|
Several predictions for planet(s) shaping the disk of SAO 206462 have been proposed (Muto et al. 2012; Garufi et al. 2013; Fung & Dong 2015; Bae et al. 2016; van der Marel et al. 2016a. Using linear equations from the spiral density wave theory, Muto et al. (2012) suggested two planets with separations beyond ~50 au by fitting independently the two spiral arms seen in Subaru/HiCIAO data and with masses of ~0.5 MJ by using the amplitude of the spiral wave. Garufi et al. (2013) proposed that one planet of mass 5–13 MJ located inside the cavity at a separation of 17.5–20 au could be responsible for the different cavity sizes measured for the small and large dust grains. Fung & Dong (2015) presented scaling relations between the azimuthal separation of the primary and secondary arms and the planet-to-star mass ratio for a single companion on a circular orbit with a mass between Neptune mass and 16 MJ around a 1 M⊙ star. They predicted with 30% accuracy that a single putative planet responsible for both spiral features of SAO 206462 would have a mass of ~6 MJ. Bae et al. (2016) presented dedicated hydrodynamical simulations of the SAO 206462 disk and proposed that both the bright scattered-light feature (Garufi et al. 2013) and the dust emission peak (Pérez et al. 2014) seen for the southwestern spiral arm result from the interaction of the spiral arm with a vortex, although a vortex alone can account for the S1 brightness peak. They suggested that a 10–15 MJ planet may orbit at 100–120 au from the star. However, ALMA observations at two different frequencies seem to contradict a dust particle trapping scenario by a vortex (Pinilla et al. 2015). Stolker et al. (2016) performed new fitting of the spiral arms observed in SPHERE data and found a best-fit solution with two protoplanets located exterior to the spirals: r1 ~ 168 au, θ1 ~ 52° and r2 ~ 99 au, θ2 ~ 355°. van der Marel et al. (2016a) proposed that the features seen in thermal emission in ALMA data and the scattered-light spiral arms are produced by a single massive giant planet located inside the cavity at a separation of ~30 au. Recently, Dong & Fung (2017) used the contrast of the spiral arms to predict a giant planet of ~5–10 MJ at ~100 au.
|
[
"Fung & Dong (2015)"
] |
[
"presented scaling relations between the azimuthal separation of the primary and secondary arms and the planet-to-star mass ratio for a single companion on a circular orbit with a mass between Neptune mass and 16 MJ around a 1 M⊙ star. They predicted with 30% accuracy that a single putative planet responsible for both spiral features of SAO 206462 would have a mass of ~6 MJ."
] |
[
"Background"
] |
[
[
676,
694
]
] |
[
[
695,
1071
]
] |
2019ApJ...883...73C___2010_Instance_1
|
Assuming that the force-field approach to the solution of the Parker (1965) cosmic-ray transport equation is valid, the connection between historic cosmic-ray intensities and the solar properties they encountered lies in the effective diffusion coefficient that is assumed in this approximation. Establishing such a connection, however, is no simple task. Many theories have been proposed to describe the scattering of cosmic rays in the heliosphere. The most likely candidates for this task, given their reasonable agreement with observations and numerical simulations of cosmic-ray diffusion coefficients, are the quasilinear theory (QLT) of Jokipii (1966), the weakly nonlinear theory (WNLT) of Shalchi et al. (2004b), and the nonlinear guiding center theory (NLGC; or one of its variants; see, e.g., Matthaeus et al. 2003; Shalchi 2006, 2009, 2010; Ruffolo et al. 2012). Shalchi (2009) provides in depth theoretical treatments of most the abovementioned theories. These scattering theories all require as a key input an expression for the power spectrum of the turbulent fluctuations of the HMF. These spectra depend upon basic turbulence quantities, such as the magnetic variance, and various correlation scales. Turbulence power spectra are discussed in detail by, e.g., Batchelor (1970) and Matthaeus et al. (2007), whereas more background on the abovementioned turbulence quantities can be found in, e.g., Matthaeus & Goldstein (1982), Petrosyan et al. (2010), Matthaeus & Velli (2011), and Bruno & Carbone (2013). These basic turbulence quantities have been observed to show a marked dependence on the solar cycle at Earth (see, e.g., Smith et al. 2006b; Burger et al. 2014; Zhao et al. 2018). It follows then that mean free paths derived from these scattering theories would be expected to depend on the solar cycle as well, and several studies have reported such a dependence. Chen & Bieber (1993) find from an analysis of cosmic-ray anisotropies and gradients as observed by means of NMs, that larger mean free paths are associated with solar minima, and smaller mean free paths with solar maxima. The authors also report a mean free path dependence on solar magnetic polarity. Nel (2016) and Zhao et al. (2018) both extensively analyze spacecraft observations, using the turbulence quantities so calculated as inputs for expressions for diffusion coefficients derived from the QLT and NLGC theories. Both authors report that the resulting mean free paths display solar cycle dependences.
|
[
"Matthaeus et al.",
"2010"
] |
[
"The most likely candidates for this task, given their reasonable agreement with observations and numerical simulations of cosmic-ray diffusion coefficients, are the quasilinear theory (QLT) of Jokipii (1966), the weakly nonlinear theory (WNLT) of Shalchi et al. (2004b), and the nonlinear guiding center theory (NLGC; or one of its variants; see, e.g.,"
] |
[
"Background"
] |
[
[
804,
820
],
[
847,
851
]
] |
[
[
451,
803
]
] |
2018AandA...610A..44M__Krüger_&_Dreizler_(1992)_Instance_2
|
The first investigations of the rotational spectra of ethyl isocyanide were carried out in 1966 by Bolton et al. (1966). The spectra of the first vibrational and torsional excited states were measured in the centimeter wave domain (Anderson & Gwinn 1968). In this initial study, the dipole moment was determined to be μa = 3.79 D and μb = 1.31 D; this value is usually large for a molecule that includes a CN group. This causes dense and intense rotational spectra in the millimeter wave range and also in the submillimeter wave range up to 900 GHz (bQ lines). Anderson & Gwinn (1968) also observed some A–E splittings due to the internal rotation motion of the methyl group. The most recent spectroscopic study is from Krüger & Dreizler (1992) who reinvestigated the internal rotation measurements and also determined hyperfine coupling parameters due to the nitrogen quadrupole. As in our previous studies of ethyl cyanide isotopologs, it was not possible to observe internal rotation and hyperfine splittings due to our Doppler limited resolution. Our analysis was rather easy, starting from a prediction based on Krüger & Dreizler (1992) parameters. First, we analyzed and fit the most intense transitions, the aRh transitions, up to 330 GHz. These transitions were shifted only a few MHz from the initial predictions. Then bR and bQ lines were searched and included in the fit up to 330 GHz. Next, all the spectra were analyzed up to 990 GHz without difficulty. For the fitting, we employed ASFIT (Kisiel 2001) and predictions were made with SPCAT (Pickett 1991). The global fits included 6 transitions from Anderson & Gwinn (1968), 29 lines from Krüger & Dreizler (1992), and 2906 from this work. The maximal quantum numbers are J = 103 and Ka = 30. Both reductions A and S were tested. A reduction permits us to check theagreement of our new parameters set with those from Krüger & Dreizler (1992) (Table 1). Using S reduction slightly decreases root mean square from 30.3 to28.7 kHz. The condition numbers are nearly the same: 295 and 310 for the A and S reductions, respectively.The A reduction requires 23 parameters, but 5 additional parameters are required for the S reduction (Table 2). For this reason we used the A reduction even if this molecule is close to the prolate limit with kappa = −0.9521. Part of the new measurements are in Table 3. Owing to its large size, the complete version of the global fit Table S1 is supplied at the CDS. The fitting files .lin (S2), .par (S3), and the prediction .cat (S4) are also available at CDS.
|
[
"Krüger & Dreizler (1992)"
] |
[
"Our analysis was rather easy, starting from a prediction based on",
"parameters."
] |
[
"Uses",
"Uses"
] |
[
[
1117,
1141
]
] |
[
[
1051,
1116
],
[
1142,
1153
]
] |
2021MNRAS.500..291B__Serafinelli_et_al._2019_Instance_1
|
We have presented the analysis of the current X-ray observations of the disc wind in MCG-03-58-007. Here, multiple and variable wind components with velocities ranging from $\sim \! -0.08\, c$ to $\sim \! - 0.2\, c$ (and potentially up to $0.35\, c$) are seen at different times. Multi-epoch observations of disc winds, like the one presented here, are crucial for revealing all the possible phases of the disc wind. For example, over a decade worth of observations of PDS 456 revealed that the wind is most likely clumpy and/or stratified with the ionization ranging from log ($\xi /\rm {erg\, cm \, s^{-1})}\sim 2$ erg cm s−1 up to log ($\xi /\rm {erg\, cm \, s^{-1})}=6$ erg cm s−1 and velocities ranging from $\sim \! -0.2\, c$ up to $\sim \! -0.46\, c$ (Reeves et al. 2016, 2018a, 2020). It is possible, as suggested in other examples of ultra fast disc winds, that we are looking at a stratified wind, where multiple components are launched at different disc radii, but not all of them are always detected. This adds MCG-03-58-007 to the small but growing list of multiphase fast X-ray winds. Other examples of AGN with at least two variable phases of the X-ray winds are PG 1211+143 (Pounds et al. 2016; Reeves et al. 2018b), IRAS 13224-3809 (Parker et al. 2017; Chartas & Canas 2018; Pinto et al. 2018), 1H 0707-495 (Kosec et al. 2018), IRAS 17020+4544 (Longinotti et al. 2015), and PG 1114+445 (Serafinelli et al. 2019). In those cases, multiple phases with a common or different outflowing velocities are detected in the X-ray band. In contrast to most of the cases reported so far, neither of the two phases seen in MCG-03-58-007 requires a different ionization (aside from slice B) suggesting that we are seeing different streamlines of the same highly ionized flow. The only exception could be the eclipsing event seen in 2015, where a solution is found with a lower ionization for the Fe K intervening absorber. However, what we most likely see during this occultation event is a higher density and lower ionization clump of the wind, which could be faster because its higher opacity makes it easier to accelerate (Waters et al. 2017). Note that this does not imply that the soft X-ray wind components, like the ones seen for example in PDS 456 or PG 1211+143, are not present; in contrast to the other examples, MCG-03-58-007 is seen through a relatively high column density (NH ∼ 2 × 1023 cm−2, see Table 2) neutral absorber, therefore these phases may be hidden behind it. MCG-03-58-007 is not the only example where multiple Fe-K zones with the same ionization and outflowing with different velocities had been detected in a single observation. For instance, two simultaneous Fe-K phases were detected at least twice in PDS456 (Reeves et al. 2018a, 2020) and possibly in PG 1211+143 (Pounds et al. 2016) and IRAS 13349+2438 (Parker et al. 2020).
|
[
"Serafinelli et al. 2019"
] |
[
"Other examples of AGN with at least two variable phases of the X-ray winds are",
"and PG 1114+445"
] |
[
"Background",
"Background"
] |
[
[
1404,
1427
]
] |
[
[
1099,
1177
],
[
1387,
1402
]
] |
2022AandA...662A..42M__Nóbrega-Siverio_et_al._2020b_Instance_1
|
In the second part of the paper, we propose the set of self-similar solutions as tests for MHD numerical codes with ambipolar diffusion capabilities. To show their usefulness and validity, a battery of tests was carried out for the Bifrost code in two spatial dimensions starting from initial conditions with cylindrical symmetry (Sect. 4). We showed that the ambipolar diffusion module in Bifrost can cope with the passage of the solutions through the current sheets, with the level of accuracy increasing the higher the spatial resolution and in spite of the intrinsic singularity in them. Vice versa, the tests show that these functions can probe the capabilities of ambipolar diffusion modules to a larger extent than the simple ZKBP solution that has been used thus far (e.g. Masson et al. 2012; Viganò et al. 2019; Nóbrega-Siverio et al. 2020b). As test functions, the various harmonics proposed in our paper have the comparative advantage that they combine the B ∝ | δ |1/3 singularity at the internal nulls (with δ the distance to the null) with the finiteness of the nonlinear diffusive flux, and this combination must be sufficiently well reproduced by the code if it is to pass the test. The ZKBP solution, instead, has a null just at the outer front, and the singularity there is of a lower order (∝|δ|1/2), with zero diffusive flux across it. On the other hand, the scarcity of tests for the ambipolar diffusion term until now is in contrast with, for instance, the case of HD shocks, for which a whole category of exact solutions is available (the solutions of the Riemann problem) that have been used to develop sophisticated numerical schemes and tests (see Laney 1998; Toro 2009). The contrast to shocks, in fact, is interesting because of the differences in their mathematical and physical nature: in shocks, it is the (magneto)hydrodynamic evolution of the hyperbolic components of the PDE that leads to the formation of the singularities, which is then smoothed through simple diffusive phenomena (typically viscosity). In the ambipolar diffusion problem, wherever there is a null, it is the diffusive phenomenon itself that creates and maintains the singularity.
|
[
"Nóbrega-Siverio et al. 2020b"
] |
[
"We showed that the ambipolar diffusion module in Bifrost can cope with the passage of the solutions through the current sheets, with the level of accuracy increasing the higher the spatial resolution and in spite of the intrinsic singularity in them. Vice versa, the tests show that these functions can probe the capabilities of ambipolar diffusion modules to a larger extent than the simple ZKBP solution that has been used thus far (e.g."
] |
[
"Compare/Contrast"
] |
[
[
821,
849
]
] |
[
[
341,
780
]
] |
2015ApJ...799..170C__Roxburgh_&_Vorontsov_2003_Instance_1
|
Stellar properties were determined using three different techniques to model the oscillation frequencies extracted from the data. The first method relies on a dense grid of stellar models computed with the GARching STellar Evolution Code (GARSTEC; Weiss & Schlattl 2008) including the effects of microscopic diffusion, and on theoretical frequencies calculated using the Aarhus aDIabatic PuLSation code (Christensen-Dalsgaard 2008a). The results were obtained implementing a Bayesian scheme that uses the spectroscopic constraints and frequency ratios as the parameters in the fit (V. Silva Aguirre et al., submitted), the latter being almost insensitive to the surface effects in solar-like oscillators (Roxburgh & Vorontsov 2003; Silva Aguirre et al. 2011). Central values are given as the estimates of the stellar properties obtained in this manner. We also computed models using the ASTEC and YREC codes. In these cases, the fit was made to the individual frequencies after correcting for the surface effect with, respectively, an appropriately scaled version of the observed solar surface correction (Christensen-Dalsgaard 2012) and a solar-type correction as described in Carter et al. (2012). The stellar properties derived using the three techniques described above are consistent within the returned formal errors. Therefore, we added in quadrature the difference in central values of each property to the formal uncertainties determined from the GARSTEC Bayesian scheme as a measurement of the systematic spread arising from different codes and fitting techniques. In Table 2, we provide a precise estimate of the stellar age, t, from detailed frequency modeling. Values for the remaining fundamental stellar properties are consistent, within errors, with those obtained from grid-based modeling. In particular, no gain in precision was obtained for the stellar radius. On the other hand, the precision on the stellar mass is improved by nearly a factor of two, from 5.7% to 3.2%.
|
[
"Roxburgh & Vorontsov 2003"
] |
[
"The results were obtained implementing a Bayesian scheme that uses the spectroscopic constraints and frequency ratios as the parameters in the fit",
", the latter being almost insensitive to the surface effects in solar-like oscillators"
] |
[
"Uses",
"Uses"
] |
[
[
705,
730
]
] |
[
[
434,
580
],
[
617,
703
]
] |
2015AandA...577A..43S__Odstrčil_&_Karlický_(1997)_Instance_2
|
The initialization of solar flares remains an unsolved problem. Early ideas on how the initialization might occur were described by Norman & Smith (1978). They argued that flare process cannot start in the entire flare volume at one instant, and proposed that the flare onset was localized in a small part of an active region, from which the energy release extends as dissipation spreading process throughout the flare volume. Two types of agents that may lead to this kind of a dissipation process were addressed: electron beams and shock waves. These agents can trigger flares at large distances from their initial locations, causing sympathetic (simultaneous) flares or leading to a sequential flare energy release in one active region (Liu et al. 2009; Zuccarello et al. 2009). These triggering processes were numerically studied by Karlický & Jungwirth (1989) and Odstrčil & Karlický (1997). Karlický & Jungwirth (1989) assumed that electron beams, penetrating into the current sheet in the magnetic reconnection region, generate Langmuir waves. Then, using the particle-in-cell model, the authors studied the effects of these electrostatic waves on the plasma system. Sufficiently strong Langmuir waves were found to be able to generate ion-sound waves through the three-wave decay process (Bárta & Karlický 2000). These ion-sound waves increase electrical resistivity in the current sheet system, which results in the onset of the energy dissipation. Thus, the electron beams are able to cause magnetic reconnection. Odstrčil & Karlický (1997) studied the mechanism for the flare trigger by shock waves. They used a 2D magnetohydrodynamic model with the MHD shock wave propagating towards the current sheet. A portion of the shock wave passed through the sheet, and the rest was reflected. Nothing occurred at the very beginning of the wave-current sheet interaction. However, after some time, specific plasma flows around the current sheet were formed, which led to the start of magnetic reconnection. This shows that for reconnection to be triggered, the enhanced electrical resistivity as well as the plasma flows are important.
|
[
"Odstrčil & Karlický (1997)"
] |
[
"studied the mechanism for the flare trigger by shock waves. They used a 2D magnetohydrodynamic model with the MHD shock wave propagating towards the current sheet. A portion of the shock wave passed through the sheet, and the rest was reflected. Nothing occurred at the very beginning of the wave-current sheet interaction. However, after some time, specific plasma flows around the current sheet were formed, which led to the start of magnetic reconnection.",
"This shows that for reconnection to be triggered, the enhanced electrical resistivity as well as the plasma flows are important."
] |
[
"Background",
"Motivation"
] |
[
[
1524,
1550
]
] |
[
[
1551,
2009
],
[
2010,
2138
]
] |
2022AandA...659A.124H__Harrison_et_al._2018_Instance_1
|
Active galactic nuclei (AGN) have drawn a lot of attention over the last decades because they have been beacons for the existence and demographics of super-massive black holes (BHs) throughout the history of the Universe (e.g., Soltan 1982; Kollmeier et al. 2006; Greene & Ho 2007; Schulze & Wisotzki 2010; Kelly & Shen 2013; Kormendy & Ho 2013; Schulze et al. 2015; Bañados et al. 2018). The release of gravitational binding energy via accretion of matter onto these BHs is expected to have a profound impact on the evolution of their host galaxies (e.g., Silk & Rees 1998; Granato et al. 2004; Di Matteo et al. 2005; Springel et al. 2005; Hopkins et al. 2008; Somerville et al. 2008; Fabian 2012; Harrison 2017; Harrison et al. 2018; Gaspari et al. 2019; Nelson et al. 2019). Large spectroscopic surveys such as the Sloan Digital Sky Survey (SDSS, York et al. 2000; Abazajian et al. 2009), the 2df QSO redshift survey (2QZ, Croom et al. 2004), the VIMOS VLT Deep Survey (VVDS, Le Fèvre et al. 2013) or the VIMOS Public Extragalactic Redshift Survey (VIPERS, Scodeggio et al. 2018) in combination with several deep X-ray surveys taken with ROSAT (Voges et al. 1999), Chandra (Elvis et al. 2009; Xue et al. 2011), XMM-Newton (Pierre et al. 2016), and eROSITA (Predehl et al. 2021) as well as large radio surveys (e.g., Becker et al. 1995; Condon et al. 1998; Smolčić et al. 2017; Shimwell et al. 2019; Lacy et al. 2020; Gordon et al. 2021) have provided an enormous data set to characterize the AGN population and its evolution with redshift in great detail. While the standard model for the AGN central engine has been successful in unifying the various classes of AGN appearance (Antonucci 1993; Urry & Padovani 1995; Padovani et al. 2017), some aspects such as changing-look AGN (CLAGN, e.g. MacLeod et al. 2016; Ruan et al. 2016; Graham et al. 2017; Noda & Done 2018) and tidal-disruption events (e.g., Gezari et al. 2009; Merloni et al. 2015; Auchettl et al. 2017) are just being explored more extensively in the time domain.
|
[
"Harrison et al. 2018"
] |
[
"The release of gravitational binding energy via accretion of matter onto these BHs is expected to have a profound impact on the evolution of their host galaxies (e.g.,"
] |
[
"Motivation"
] |
[
[
714,
734
]
] |
[
[
389,
556
]
] |
2019AandA...623A.140G__Pohl_et_al._2017_Instance_1
|
HD 169142 is a very young Herbig Ae-Be star with a mass of 1.65–2 M⊙ and an age of 5–11 Myr (Blondel & Djie 2006; Manoj et al. 2007) that is surrounded by a gas-rich disk (i = 13°; Raman et al. 2006; PA = 5°; Fedele et al. 2017) that is seen almost face-on. The parallax is 8.77 ± 0.06 mas (Gaia DR2 2018). Disk structures dominate the inner regions around HD 169142 (see, e.g., Ligi et al. 2018). Figure 1 shows the view obtained from polarimetric observations: the left panel shows the QΦ image in the J band obtained by Pohl et al. (2017) using SPHERE on a linear scale, and the two rings are clearly visible. The right panel shows a pseudo-ADI image of the inner regions obtained by differentiating the QΦ image (see Ligi et al. 2018, for more details). Biller et al. (2014) and Reggiani et al. (2014) discussed the possible presence of a point source candidate at small separation (0.2 arcsec from the star). However, the analysis by Ligi et al. (2018) based on SPHERE data does not support or refute these claims; in particular, they suggested that the candidate identified by Biller et al. (2014) might be a disk feature rather than a planet. Polarimetricimages with the adaptive optics system NACO at the Very Large Telescope (VLT; Quanz et al. 2013b), SPHERE (Pohl et al. 2017; Bertrang et al. 2018) and GPI (Monnier et al. 2017) show a gap at around 36 au, with an outer ring at a separation >40 au from the star. This agrees very well with the position of the rings obtained from ALMA data (Fedele et al. 2017); similar results were obtained from VLA data (Osorio et al. 2014; Macías et al. 2017). We summarize this information about the disk structure in Table 1 and call the ring at 0.17–0.28 arcsec from the star Ring 1 and the ring at 0.48–0.64 arcsec Ring 2. We remark that in addition to these two rings, both the spectral energy distribution (Wagner et al. 2015) and interferometric observations (Lazareff et al. 2017; Chen et al. 2018) show an inner disk at a separation smaller than 3 au. This inner disk isunresolved from the star in high-contrast images and consistent with ongoing accretion from it onto the young central star.While the cavities between the rings seem devoid of small dust, some gas is present there (Osorio et al. 2014; Macías et al. 2017; Fedele et al. 2017). Fedele et al. (2017) and Bertrang et al. (2018) have suggested the possibility that the gap between Rings 1 and 2 is caused by a planet with a mass slightly higher than that of Jupiter. However, this planet has not yet been observed, possibly because it is at the limit of or beyond current capabilities of high-contrast imagers. On the other hand, Bertrang et al. (2018) found a radial gap in Ring 1 at PA ~ 50° that might correspond to a similar radial gap found by Quanz et al. (2013b) at PA ~ 80°. The authors noted that if this correspondence were real, then this gap might be caused by a planet at about 0.14 arcsec from the star. So far, this planet has not been unambiguously detected either.
|
[
"Pohl et al. (2017)"
] |
[
"Figure 1 shows the view obtained from polarimetric observations: the left panel shows the QΦ image in the J band obtained by",
"using SPHERE on a linear scale, and the two rings are clearly visible."
] |
[
"Uses",
"Uses"
] |
[
[
523,
541
]
] |
[
[
398,
522
],
[
542,
612
]
] |
2022AandA...666A.190S___2014b_Instance_1
|
For our dataset of absolute magnitudes, we used data collected at the Institute of Astronomy of V. N. Karazin Kharkiv National University within the long-term observational programme to study asteroid magnitude-phase curves (Shevchenko et al. 2010, 2012, 2014a, 2016; Slyusarev et al. 2012). We also used some observational data obtained within several other programmes (Belskaya et al. 2010; Chiorny et al. 2007, 2011; Dotto et al. 2009; Hahn et al. 1989; Kaasalainen et al. 2004; Lagerkvist et al. 1998; Michalowski et al. 1995; Mohamed et al. 1994, 1995; Oszkiewicz et al. 2021; Shevchenko et al. 1992, 2003, 2009, 2014b, 2021; Velichko et al. 1995; Wilawer et al. 2022). All magnitudes were measured in the Johnson V band and extrapolated to zero phase angle using the HG1G2 system proposed by Muinonen et al. (2010), with some modifications presented by Penttilä et al. (2016). For computations, the online calculator1 for the HG1G2 photometric system was used. Since we derived absolute magnitudes in our data from the light curve maxima, and the definition of H is based on the rotationally averaged brightness, we added a half of the light curve amplitude corrected to zero phase angle to our results. We used the average correction coefficients from Zappala et al. (1990) for low- and moderate-albedo asteroids. This correction is typically very small because our light curve observations covered small phase angles. Absolute magnitudes obtained at different aspects were averaged. In such a manner, we obtained a homogeneous dataset of absolute magnitudes of about 400 asteroids up to H = 16.5 mag. Our database includes the absolute magnitude data, the G1 and G2 parameters, and the albedo and diameter values from different databases (such as Tedesco et al. 2002; Masiero et al. 2011, 2012; Nugent et al. 2015; Usui et al. 2011). The database is available at the CDS. Figure 1 shows the correlations of the absolute magnitudes from the largest datasets (MPC (HMPC), Pan-STARRS (HPS), and ATLAS (HATLAS)) with those of the Kharkiv dataset (HKH). For the ATLAS dataset, we used the absolute magnitudes in a cyan filter, since this filter overlaps the Johnson V band (Mahlke et al. 2021).
|
[
"Shevchenko et al.",
"2014b"
] |
[
"We also used some observational data obtained within several other programmes"
] |
[
"Uses"
] |
[
[
582,
599
],
[
618,
623
]
] |
[
[
292,
369
]
] |
2017AandA...598A..21B__Formicola_et_al._(2004)_Instance_1
|
In this section, we describe a few numerical experiments carried out to analyse the importance of various hypotheses used to compute structural kernels. All models were computed using the Clés stellar evolution code (Scuflaire et al. 2008b) with the following ingredients: the CEFF equation of state (Christensen-Dalsgaard & Daeppen 1992), the OPAL opacities from Iglesias & Rogers (1996), supplemented at low temperature by the opacities of Ferguson et al. (2005) and the effects of conductivity from Potekhin et al. (1999) and Cassisi et al. (2007). The nuclear reaction rates are those from the NACRE project (Angulo et al. 1999), supplemented by the updated reaction rate from Formicola et al. (2004) and convection was implemented using the classical, local mixing-length theory (Böhm-Vitense 1958). We also used the implementation of microscopic diffusion from Thoul et al. (1994), for which three groups of elements are considered and treated separately: hydrogen, helium and the metals (all considered to have diffusion speeds of 56Fe). The oscillation frequencies and eigenfunctions were computed using the Liège adiabatic oscillation code (Scuflaire et al. 2008a). We took much care to analyse the numerical quality of the eigenfunctions and the models before computing structural kernels. Irregularities and poor quality of the computed eigenfunctions can bias the results and lead to wrong structural kernels and thus wrong inferences from inverted results. From our experience in hare-and-hounds exercises and inversions, we have determined that adding seismic constraints to the model is very efficient at bringing the reference model into the linear regime thus validating the inversion process. In other words, fitting the average large and small frequency separations is already a big improvement in terms of linearity, although individual seismic constraints, such as individual frequency ratios and individual small frequency separations are the best way to maximise the chances of being in the linear regime. Individual large frequency separations can also be used, but due to their sensitivity to surface effects, they should not be used in observed cases. As such, since in this study we did not use very elaborate seismic fitting techniques, our tests serve the only purpose of isolating various contributions to the errors and to test various hypotheses usually done when carrying out structural inversions in the context of helio- and asteroseismology.
|
[
"Formicola et al. (2004)"
] |
[
"The nuclear reaction rates are those from the NACRE project",
", supplemented by the updated reaction rate from"
] |
[
"Uses",
"Uses"
] |
[
[
681,
704
]
] |
[
[
552,
611
],
[
632,
680
]
] |
2017ApJ...835...25E__Rutten_1984_Instance_2
|
We compare our results with a new reduction of observations from the Lowell Observatory SSS, which is running a long-term stellar activity survey complementary to the MWO HK Project. The SSS observes solar and stellar light with the same spectrograph, with the solar telescope consisting of an exposed optical fiber that observes the Sun as an unresolved source (Hall & Lockwood 1995; Hall et al. 2007). The basic measurement of SSS is the integrated flux in 1 Å bandpasses centered on the Ca ii H & K cores from continuum-normalized spectra, ϕHK, which can then be transformed to the S-index using a combination of empirical relationships derived from stellar observations:
7
where
is the continuum flux scale for the Ca ii H & K wavelength region, which converts ϕHK to physical flux (erg cm−2 s−1).
is a function of Strömgren
and is taken from Hall (1996).
(simply K in other works) is the conversion factor from the MWO HKP-2 H & K flux (numerator of Equation (1)) to physical flux (Rutten 1984). Ccf is a factor that removes the color term from S and is a function of Johnson
(Rutten 1984). Finally, Teff is the effective temperature. See Hall et al. (2007) and Hall & Lockwood (1995) for details on the extensive work leading to this formulation. What is important to realize about this method of obtaining S is that it requires three measurements of solar properties,
,
, and
, along with the determination of one constant,
. The solar properties are taken from best estimates in the literature, which vary widely depending on the source used, and can dramatically affect the resulting SSSS for the Sun. Hall et al. (2007) used
,
, and
. The constant
was empirically determined to be 0.97 ± 0.11 erg cm−2 s−1 in Hall et al. (2007) as the value that provides the best agreement between SSSS and
from Baliunas et al. (1995) for an ensemble of stars and the Sun. This combination of parameters resulted in a mean SSSS of 0.170 for the Sun using observations covering cycle 23. A slightly different calibration of SSS data in Hall & Lockwood (2004) used a flux scale
based on Johnson
, set to 0.65 for the Sun, and
K. In Table 1 we estimated that this calibration resulted in a mean S = 0.168 for cycle 23. Hall et al. (2009), which included a revised reduction procedure and one year of data with the upgraded camera (see below), found
.
|
[
"Rutten 1984"
] |
[
"Ccf is a factor that removes the color term from S and is a function of Johnson"
] |
[
"Uses"
] |
[
[
1122,
1133
]
] |
[
[
1034,
1113
]
] |
2016AandA...587A.133G__Bruzual_&_Charlot_(2003)_Instance_1
|
Notes. Intrinsic Lν(1400)/Lν(900) ratios obtained from different stellar population models: the BC03, assuming different initial mass functions (Salpeter, Salp, and Chabrier, Chab), star-formation histories (constant, cSFR, exponentially declining with τ = 0.1Gyr, exponentially rising with τ = −0.1Gyr), metallicties (Z⊙ and 0.02 × Z⊙), and stellar-population ages; the BPASS (Binary Population and Spectral Synthesis code Stanway et al. 2016); the Starburst99 models. We consider a range of wavelengths around 900 Å and 1400 Å, comparable to that covered by the narrow-band and the HST/ACS F606W filters in the rest frame, to estimate the average Lν. As described in Bruzual & Charlot (2003), BC03 provides the spectral energy distribution of stars obtained from a comprehensive library of theoretical model atmospheres (KBFA in the table). The library consists of Kurucz (1996) spectra for O-K stars, Bessell et al. (1991) and Fluks et al. (1994) spectra for M giants, and Allard & Hauschildt (1995) spectra for M dwarfs. In BPASS the stellar evolutionary tracks contain a contribution from isolated stars and binary systems. Also, the stellar atmosphere models are selected from the BaSeL v3.1 library (Westera et al. 2002), supplemented by Wolf-Rayet stellar atmosphere models from the Potsdam PoWR group (Hamann & Gräfener 2003). With Starburst99, it is possible to generate SEDs assuming a bunch of stellar evolutionary tracks, including stars with and without rotation (GenevaV40 and GenevaV00 respectively, Leitherer et al. 2014), and stellar atmospheres (the combination of model atmosphere from Pauldrach et al. 1998; Hillier & Miller 1998, is the recommended option). The 70%ROT+30%noROT entry indicates a model that is a combination of GenevaV40 for the 70% and GenevaV00 for the 30% (Levesque et al. 2012). The GenevaV40 and GenevaV00 tracks were released for metallicity equal to Z = 0.014 (~Z⊙, Eldridge 2012). The change in the intrinsic ratio due to the different evolutionary tracks and stellar atmospheres is shown for a Salpeter IMF, Z⊙, cSFR, and 100 Myr old stellar population. In the bottom part of the table, we report the ratios calculated from the best fit templates of AGN (Bongiorno et al. 2012), in which the SEDs are dominated by the radiation coming from an (un-)obscured AGN. (1) The intrinsic Lν(1400)/Lν(900) ratio for an unobscured TypeII AGN can be as low as 2.35 (Richards et al. 2006). In the estimation of the ratios we take into account the HI absorption occurred in star atmospheres, but we neglect that within the interstellar and circum-galactic medium.
|
[
"Bruzual & Charlot (2003)"
] |
[
"Intrinsic Lν(1400)/Lν(900) ratios obtained from different stellar population models: the BC03, assuming different initial mass functions (Salpeter, Salp, and Chabrier, Chab), star-formation histories (constant, cSFR, exponentially declining with τ = 0.1Gyr, exponentially rising with τ = −0.1Gyr), metallicties (Z⊙ and 0.02 × Z⊙), and stellar-population ages;",
"As described in",
"BC03 provides the spectral energy distribution of stars obtained from a comprehensive library of theoretical model atmospheres (KBFA in the table).",
"The library consists of Kurucz (1996) spectra for O-K stars, Bessell et al. (1991) and Fluks et al. (1994) spectra for M giants, and Allard & Hauschildt (1995) spectra for M dwarfs."
] |
[
"Uses",
"Uses",
"Uses",
"Background"
] |
[
[
669,
693
]
] |
[
[
7,
366
],
[
653,
668
],
[
695,
842
],
[
843,
1024
]
] |
2021AandA...656A..64R__Ruiz-Lara_et_al._2020_Instance_1
|
Finally, we confirm our hypothesis on the origin of the bi-modal A(Li) distribution, and of the isthmus, by comparing results from our GCE model with data. In Fig. 8 we show the normalised A(Li) distribution of all the stars in our sample (blue histogram) compared with the result from three models of GCE: (i) imposing a gently declining SFH (SFR_0b model, dashed red lines); (ii) a gently declining SFH plus a single star formation burst at a ∼6 Gyr look-back time (SFR_1b model, dashed blue lines); and (iii) our best model (SFR_2b, black-solid lines), which contains a gently declining SFR plus two star formation bursts at ∼6 Gyr and ∼1.5 Gyr look-back times, the latter being the most impulsive. In our best model the first star formation peak is only 68% as intense as the latter and lasts 25% longer. The SFR_2b model is the one that best reproduces the observed A(Li) distribution while recovering the presently observed SFR and the thin-disk stellar and HI mass (Mor et al. 2019; Isern 2019; Ruiz-Lara et al. 2020). This result is robust as it only depends on the observed A(Li) distribution. Possible biases from the determination of the stellar ages do not have an impact on our conclusions. Figure 8 clearly shows that only a two-peaked SFH can explain the observed bi-modal A(Li) distribution when accounting for most known 7Li production and depletion mechanisms (solid black line). It is important to remark that we cannot quantify the exact intensity and length of the star formation bursts from our data and models, but we can only give a qualitative analysis. This is a consequence of: (i) the large age determination uncertainties still present in current data; (ii) the fact that the data we use is restricted to the solar neigbourhood and thus do not represent the entire Galaxy; and finally (iii) the fact that our model still misses some 7Li production and destruction processes that are currently under study (e.g. the effect of planet engulfment). In the future, new data from large surveys such as Galah (Buder et al. 2021) combined with better determinations of A(Li) in 7Li-poor low-luminosity cold dwarf stars will result in large, unbiased samples of MS stars. These new catalogues will allow us to better understand the many channels of 7Li production and destruction that occur inside dwarf stars. Also, future catalogues that include precise age determinations will allow us to better constrain the SFH of the MW, also by analysing its A(Li) distribution.
|
[
"Ruiz-Lara et al. 2020"
] |
[
"The SFR_2b model is the one that best reproduces the observed A(Li) distribution while recovering the presently observed SFR and the thin-disk stellar and HI mass"
] |
[
"Compare/Contrast"
] |
[
[
1002,
1023
]
] |
[
[
809,
971
]
] |
2016MNRAS.457.2236K__Ouyed_&_Pudritz_1997_Instance_1
|
Strong outflows are commonly associated with the early stages of stellar evolution. They are likely responsible for transporting excess angular momentum away from the star–disc system and regulating the mass-accretion process and spin evolution of newly born stars (e.g. Hartmann & Stauffer 1989; Matt & Pudritz 2005; Bouvier et al. 2014). There are at least three possible types of outflows around young stellar objects: (1) a disc wind launched from the accretion disc (e.g. Blandford & Payne 1982; Ustyugova et al. 1995, 1999; Romanova et al. 1997; Ouyed & Pudritz 1997; Königl & Pudritz 2000; Pudritz et al. 2007), (2) an X-wind or a conical wind launched near the disc–magnetosphere interaction region (e.g. Shu et al. 1994, 1995; Romanova et al. 2009) and (3) a stellar wind launched from open magnetic field lines anchored to the stellar surface (e.g. Decampli 1981; Hartmann & MacGregor 1982; Kwan & Tademaru 1988; Hirose et al. 1997; Strafella et al. 1998; Matt & Pudritz 2005; Romanova et al. 2005; Cranmer 2009). However, despite recent efforts, the exact launching mechanisms of the winds and outflows as well as the mechanism behind the collimation of the ejected gas into jets are still not well understood (e.g. Edwards et al. 2006; Ferreira, Dougados & Cabrit 2006; Kwan, Edwards & Fischer 2007; Tatulli et al. 2007b; Kraus et al. 2008; Eisner, Hillenbrand & Stone 2014). Hence, high-resolution interferometric observations which can resolve the wind-launching regions are crucial for addressing this issue. If a high spectral resolution is combined with a high spatial resolution, the emission-line regions near the base of the wind (e.g. Brγ emission regions in Herbig Ae/Be stars) can be resolved in many spectral channels across the line (e.g. Weigelt et al. 2011; Kraus et al. 2012b,c; Garcia Lopez et al. 2015). This allows us to study the wavelength-dependent extent and the kinematics of the winds, which can be derived from the line visibilities and wavelength-dependent differential and closure phases. Such observations would help us to distinguish the different types of outflow scenarios.
|
[
"Ouyed & Pudritz 1997"
] |
[
"There are at least three possible types of outflows around young stellar objects: (1) a disc wind launched from the accretion disc (e.g."
] |
[
"Background"
] |
[
[
552,
572
]
] |
[
[
340,
476
]
] |
2021AandA...647A.177D__Awad_&_Abu-Shady_2020_Instance_1
|
An estimate of a sputtering cross-section can be inferred from our measurements with σs ≈ V∕d, where V is the volume occupied by
${Y}_{\textrm{s}}^{\infty}$Ys∞
molecules and d the depth of sputtering.
$\sigma_{\textrm{s}}\approx {Y}_{\textrm{s}}^{\infty}/l_{\textrm{d}}/\textrm{ml}$σs≈Ys∞/ld/ml
, where ml is the number of CO or CO2 molecules cm−2 in a monolayer (about 6.7 × 1014 cm−2 and 5.7 × 1014 cm−2, respectively,with the adopted ice densities). As is shown in Table 1, the sputtering radius rs would therefore be about 1.26 to 2.12 times larger than the radiolysis destruction radius rd in the case of the CO2 ice, and 2.03 to 2.36 for CO in the considered energy range (~0.5-1 MeV/u). The net radiolysis is the combined effect of the direct primary excitations and ionisations, the core of the energy deposition by the ion, and the so-called delta rays (energetic secondary electrons) travelling at much larger distances from the core; that is, several hundreds of nanometres at the considered energies in this work (e.g. Mozunder et al. 1968; Magee & Chatterjee 1980; Katz et al. 1990; Moribayashi 2014; Awad & Abu-Shady 2020). The effective radiolysis track radius that we calculate is lower than the sputtering one, which points towards a large fraction of the energy deposited in the core of the track. The scatter on the ratio of these radii is due to the lack of more precise data. It nevertheless allows to put a rough constraint on the estimate of Nd in the absence of additional depth measurements, with
${N}_{\textrm{d}} \lesssim {Y}_{\textrm{s}}^{\infty}/\sigma_{\textrm{d}}$Nd≲Ys∞/σd
. If the rs∕rd ratio is high, a large amount of species come from the thermal sublimation of an ice spot less affected by radiolysis, and the fraction of ejected intact molecules is higher. The aspect ratio corresponding to these experiments evolves between about ten and twenty for CO2 and CO, whereas for water ice at a stopping power of Se ≈ 3.6 × 103eV∕1015 H2 O molecules cm−2, we show that it is closer to one (Dartois et al. 2018). The depth of sputtering is much larger for CO and CO2 than for H2O at the same energy deposition, not only because their sublimation rate is higher, but also because they do not form OH bonds. For complex organic molecules embedded in ice mantles dominated by a CO or CO2 ice matrix, with the lack of OH bonding and the sputtering for trace species being driven by that of the matrix (in the astrophysical context), the co-desorption of complex organic molecules present in low proportions with respect to CO/CO2 cannot only be more efficient, but will thus arise from deeper layers.
|
[
"Awad & Abu-Shady 2020"
] |
[
"The net radiolysis is the combined effect of the direct primary excitations and ionisations, the core of the energy deposition by the ion, and the so-called delta rays (energetic secondary electrons) travelling at much larger distances from the core; that is, several hundreds of nanometres at the considered energies in this work"
] |
[
"Uses"
] |
[
[
1119,
1140
]
] |
[
[
699,
1029
]
] |
2021MNRAS.502.3179T__Sreenivasan,_Ramshankar_&_Meneveau_1989_Instance_1
|
These properties can also be related to the fractal nature of radiative mixing layers. Recently, Fielding et al. (2020) showed that the area of the cooling surface in radiative mixing layer simulations obeys a fractal scaling, with
(31)$$\begin{eqnarray*}
\frac{A_{\rm T}}{A_{\rm L}} = \left(\frac{\lambda }{L} \right)^{2-D},
\end{eqnarray*}$$where λ is the smoothing scale and D = 2.5 was the fractal dimension argued to hold by analogy with well-known fractals, and verified in their simulations. Turbulence combustion fronts are indeed well known to be fractals, due to the dynamical self-similarity of turbulence in the inertial range. Experimental measurements by e.g. instantaneous laser tomography have given values ranging from D = 2.1−2.4 in a variety of flow geometries, with a preferred value of D = 2.35 (Hentschel & Procaccia 1984; Sreenivasan, Ramshankar & Meneveau 1989); it has been argued that this fractal dimension is universal (Catrakis, Aguirre & Ruiz-Plancarte 2002; Aguirre & Catrakis 2005). From equation (19), the fractal dimension can be used to calculate the turbulent flame speed (Gouldin et al. 1986; Peters 1988). The fractal scaling and consequent increase in area AT should extend all the way down to the Gibson scale λG, which is defined to be the scale where the turbulent velocity equals the laminar flame speed, v(λG) = SL. This is often unresolved in simulations. If we use the Kolmogorov scaling v∝λ1/3, then we obtain:
(32)$$\begin{eqnarray*}
\frac{S_{\rm T}}{S_{\rm L}} = \frac{A_{\rm T}}{A_{\rm L}} = \left(\frac{\lambda _{\rm G}}{L} \right)^{2-D} = \left(\frac{u^{\prime } }{S_{\rm L}} \right)^{3(D-2)} ,
\end{eqnarray*}$$where we have used equation (31) and v(λG) = SL. Thus, in equation (20), we have n = 3(D − 2). The experimental value of D = 2.35 gives n = 1.05, in good agreement with Damköhler’s scaling, and fair agreement with the scaling in equation (23). The Fielding et al. (2020) value of D = 2.5 gives n = 1.5, or ST = u′(u′/SL)1/2. If one uses the laminar $S_L \propto t_{\rm cool}^{-1/2}$ from our static simulations, this would imply $S_T \propto t_{\rm cool}^{1/4}$. However, in the Fielding et al. (2020) model, the speed at which a cooling layer advances is $S_L \propto t_{\rm cool}^{1/2}$, so they end up with $S_T \propto t_{\rm cool}^{-1/4}$ as well. The scalings are sensitive to the fractal dimension D and the measurement error on D obtained from the simulations is unclear at this point. In addition, the cutoff scale of turbulence may not be the Gibson scale. We caution that fractal arguments have not proven to be fully robust in the turbulent combustion context. For instance, the measured fractal parameters fluctuate depending on the extraction algorithm, and have not been able to correctly predict the turbulent burning velocity (Cintosun, Smallwood & Gülder 2007).
|
[
"Sreenivasan, Ramshankar & Meneveau 1989"
] |
[
"Experimental measurements by e.g. instantaneous laser tomography have given values ranging from D = 2.1−2.4 in a variety of flow geometries, with a preferred value of D = 2.35"
] |
[
"Uses"
] |
[
[
847,
886
]
] |
[
[
642,
817
]
] |
2020AandA...644A..88K__Shapovalova_et_al._2010b_Instance_1
|
Type 1 AGN (NLSy1, quasars) and aligned CB-SMBH models expected signatures. Two SMBH and their BLRs induce more rich and complex differential phase patterns. There are many configurations for which the aligned CB-SMBH differs between them and single SMBH. The features of synthetic spectra given in Figs. C.2b and C.2c show a similarity with those found in 3C 390.3 (Landt et al. 2014, see their Fig. A1). This object is well-known as a double-peaked line emitter in the optical band. The double-peaked profiles can be also associated with accretion disc emission (Eracleous & Halpern 1994, 2003; Gezari et al. 2007). However, if the binary model is appropriate for this object, then associated differential phases will have a complex double S-shaped structure. This model reflect the possibility of a non-coplanar CB-SMBH, with highly inclined orbits of clouds in both BLRs. If a single SMBH model with a BLR d of 95 light days is true (see Shapovalova et al. 2010b) a differential signal would be 0.9°. Even that spectral lines of objects can share some characteristics, our model predicts that corresponding differential phases are distinct because of different SMBH and the clouds’ orbital parameters. The optically bright quasar PG 1211+143 (Landt et al. 2014) has a convex Paα shape slightly depressed of the centre, as in our synthetic case present in Fig. C.2a. The predicted differential phase would resemble an asymmetric double S shape, caused by non-coplanarity of the CB-SMBH system and high values for the clouds’ orbital elements Ωc and ωc. Moreover, the high rise spectral line in Fig. 5a is also observed in the spectrum of SDSSJ032213.90+005513.4 (Kim et al. 2010, see their Fig. 1). If the CB-SMBH model is appropriate for this object, then the expected differential phase would be similar in morphology with the previous case but distinct in their details because of different values of Ω2. In addition, asymmetric two horn feature (blue line) found in Fig. C.3b, is also highly consistent with observed line in Mrk 79 (Landt et al. 2008, see their Fig. 13). Anticipated differential phases differ from those seen in previous cases because of large values, 230° Ω1, Ω2 330° , and decreasing ωk, k = 1, 2.
|
[
"Shapovalova et al. 2010b"
] |
[
"If a single SMBH model with a BLR d of 95 light days is true (see",
"a differential signal would be 0.9°."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
942,
966
]
] |
[
[
876,
941
],
[
968,
1004
]
] |
2020ApJ...903...46C__Banerjee_et_al._2009_Instance_1
|
The large values of α (significantly larger than unity) observed in clouds of low column densities or masses are often interpreted in terms of the presence of a large external confining pressure (P/
e.g., Keto & Myers 1986; Oka et al. 2001; Field et al. 2011; Leroy et al. 2015; Traficante et al. 2018a), although it is hard to imagine that such high pressures can be thermal in general, since the mean ambient thermal pressure in the interstellar medium (ISM) is rather low,
, and large deviations from it occur very infrequently (e.g., Boulares & Cox 1990; Jenkins 2004; Jenkins & Tripp 2011). Instead, it is most likely that these values correspond to ram pressure, in which case they imply mass, momentum, and energy flux across Eulerian cloud boundaries, or a displacement of Lagrangian boundaries (Ballesteros-Paredes et al. 1999; Banerjee et al. 2009). Indeed, in a previous study (Camacho et al. 2016), we find, through measurement of the mean velocity divergence within the clouds in numerical simulations of cloud formation and evolution, that roughly half the clouds with an excess of kinetic energy are undergoing compression. This can be interpreted as the clouds being subject to a ram pressure (which amounts to an inertial compression) that is making them denser and smaller, so that they eventually will become gravitationally bound. The origin of this ram pressure can be large-scale turbulence, a large-scale potential well, or other instabilities. Furthermore, one important possibility is that clouds may be falling into the potential well of a stellar spiral arm, which is the main source of large-scale compression for the gas in the Galactic disk (Roberts 1969). Thus, this is not really a “confinement,” since the clouds are not at rest. The same goes for the other half of the clouds, which are undergoing expansion. In this case, the excess of kinetic energy corresponds to the expansion motions, and again the cloud is not confined, so there is no need for a high confining pressure. In Camacho et al. (2016) and Ballesteros-Paredes et al. (2016, hereafter BP+18) it has been suggested that, for clouds formed by inertial compressions in the background medium (Ballesteros-Paredes et al. 1999), and which gradually become more strongly gravitationally bound, while the inertial compressive motions decay or dissipate, the kinetic energy transits from being dominated by the inertial motions to being dominated by the gravitationally driven motions (see also Collins et al. 2012). In that case, an initial decay of the Larson ratio and the virial parameter may be expected.
|
[
"Banerjee et al. 2009"
] |
[
"Instead, it is most likely that these values correspond to ram pressure, in which case they imply mass, momentum, and energy flux across Eulerian cloud boundaries, or a displacement of Lagrangian boundaries"
] |
[
"Uses"
] |
[
[
863,
883
]
] |
[
[
622,
828
]
] |
2019MNRAS.488..803M__Long_2017_Instance_1
|
The identification and confirmation of extragalactic SNRs are mainly done using data from the radio (Lacey, Duric & Goss 1997; Hyman et al. 2001; Lacey & Duric 2001), visible (Matonick & Fesen 1997; Matonick et al. 1997; Gordon et al. 1998; Blair & Long 2004; Sonbas, Akyuz & Balman 2009; Sonbaş et al. 2010), and X-Ray (Pence et al. 2001; Ghavamian et al. 2005) wavelength ranges. Only a few extragalactic SNRs have been identified in the infrared using the [Fe ii] 1.64 $\hbox{$\mu $m}$ emission line (Greenhouse et al. 1997). The radiation of SNRs in different wavelength ranges is under the influence of biases as they cover different aspects of the ISM environment and SNR age and evolution (Pannuti et al. 2000; Leonidaki, Zezas & Boumis 2010; Sánchez-Ayaso et al. 2012; Long 2017). In the radio, only SNR candidates associated with H ii regions are identified, which means that radio SNR samples are biased towards star-forming regions. In the X-Ray, candidates are selected if they display a soft spectrum and if they are associated with an H ii region, which means that X-Ray samples have the same bias than the radio samples and are also biased against SNR candidates with hard spectra and no optical counterparts. In the optical, samples are under the influence of biases privileging the identification of SNRs located in low-density environments. A multiwavelength (X-Ray, optical, and radio) study of SNRs in NGC 300 by Pannuti et al. (2000) revealed 16 new SNRs, 2 in the radio, and 14 in the X-Ray, in addition to the 28 SNRs previously identified in the optical. The lack of new optical detection is explained by the fact that optical SNRs can only be detected when they represent relatively low confusion with other H α emission sources. The optical SNRs found here are generally located well away from star-forming regions. Consequently, SNR samples identified optically are often not complete. Another technique, based on the search for Large-Velocity-Width Sources (LVWS), is used to identify SNRs with optical spectroscopy (Chu & Kennicutt 1986; Castaneda, Vilchez & Copetti 1990). While H ii regions show low-velocity dispersion (σv ≤ 30 km s−1; Melnick 1977; Gallagher & Hunter 1983), broad emission line widths observed in the LVWS can be caused by stellar winds or SN. Using this technique, Chu & Kennicutt (1986) discovered four LVWS inside the Giant H ii regions NGC 5471 A, B, C, and NGC 5461 in the nearby galaxy M101.
|
[
"Long 2017"
] |
[
"The radiation of SNRs in different wavelength ranges is under the influence of biases as they cover different aspects of the ISM environment and SNR age and evolution"
] |
[
"Background"
] |
[
[
777,
786
]
] |
[
[
529,
695
]
] |
2018AandA...610A..10C__Montaigne_et_al._2005_Instance_1
|
According to chemical models, the molecules HCS+ and CS are among the most closely related species present in the ISM, as they participate in a direct exchange in both the formation and destruction of CS (Drdla et al. 1989; Lucas & Liszt 2002). In the reaction network considered by Drdla et al. (1989) and Lucas & Liszt (2002), CS is believed to be formed primarily from the dissociative recombination reaction of (6)\begin{equation} \label{eq:HCSdisRecomb} \text{HCS}^+ + {\rm e}^- \rightarrow \text{CS} + \text{H}, \end{equation}HCS++e−→CS+H,where HCS+ is first formed by reactions beginning with S+. However, recent experiments indicate that this product channel occurs in only 19 percent of collisions, with 81 percent of collisions forming CH + S (Montaigne et al. 2005). Three mechanisms dominate the destruction of CS in these networks. In the first two, photoionization and ion-exchange reactions destroy CS to form CS+, by (7)\begin{equation} \label{eq:CSphotoIo} \text{CS} + \gamma \rightarrow \text{CS}^+ + {\rm e}^- , \end{equation}CS+γ→CS++e−,and (8)\begin{equation} \label{eq:CSionExchange} \text{CS} + X^+ \rightarrow \text{CS}^+ + X, \end{equation}CS+X+→CS++X,where X+ is a cationic species and X is the corresponding neutral species. CS+ then quickly reacts to form HCS+: (9)\begin{equation} \label{eq:CSio_recomb} \text{CS}^+ + \text{H}_2 \rightarrow \text{HCS}^+ + \text{H} . \end{equation}CS++H2→HCS++H.In the third destruction route, CS reacts with H\hbox{$_3^+$}+3 to form HCS+ directly: (10)\begin{equation} \label{eq:CSandH3plus} \text{CS} + \text{H}_3^+ \rightarrow \text{HCS}^+ + \text{H}_2 . \end{equation}CS+H3+→HCS++H2.As CS is formed by a reaction of HCS+ and the dominant destruction pathways for CS form HCS+, we would expect the abundances of CS and HCS+ to be delicately balanced, and for the two abundances to track each other. While chemical models of sulfur-bearing chemistry in diffuse clouds underpredict the abundance of HCS+ by multiple orders of magnitude, Lucas & Liszt (2002) determined that if the observed abundance of HCS+ is injected into a diffuse cloud, it is possible to account for the CS abundances observed in their sample, further emphasizing the theoretical importance of NCS/NHCS+. In clouds in the Galactic center, where H\hbox{$_3^+$}+3 is ≳10 times more abundant compared to diffuse clouds in the disk (Oka et al. 2005), we might expect an offset in the ratio, tending toward a higher relative abundance of HCS+.
|
[
"Montaigne et al. 2005"
] |
[
"However, recent experiments indicate that this product channel occurs in only 19 percent of collisions, with 81 percent of collisions forming CH + S"
] |
[
"Compare/Contrast"
] |
[
[
754,
775
]
] |
[
[
604,
752
]
] |
2020ApJ...904..117K___2014b_Instance_1
|
Simultaneously solving the MHD equations and the global angle- and energy-dependent radiation transport equation, in general relativity, is both computationally expensive (typically prohibitively so) and technically challenging. Even so, significant progress has been made in the last decade, though the problem is usually made tractable by introducing at least one of the following simplifying assumptions: abandoning general relativity in favor of a pseudo-Newtonian description of the gravitational potential while performing realistic, multi-angle group radiation transport (Jiang et al. 2014a, 2014b, 2019a, 2019b); limiting the possible angular dependence of the radiation field by invoking either flux-limited diffusion (Zanotti et al. 2011; Roedig et al. 2012) or, more recently, the “M1 closure” relation in either axisymmetric (2D; Sadowski et al. 2014) or 3D simulations (Fragile et al. 2012, 2014; McKinney et al. 2014; Sadowski et al. 2016); or using Monte Carlo (Ryan et al. 2015) / hybrid Monte Carlo techniques (Ryan & Dolence 2020). Most attempts have eschewed energy-dependent transfer in favor of a “gray” atmosphere—the radiation field is treated as monochromatic, coupled to the fluid only through the Rosseland mean opacity (Rybicki & Lightman 1986). The first of these approximations, the pseudo-Newtonian potential, is especially problematic in regions close to the black hole, where general relativistic effects play a critical role in determining both the structure of the accretion flow and the photon trajectories. The others are essentially variants of a diffusion approximation and are best suited to the cooler, denser, and optically thick body of the accretion disk, where the environment is similar to those found in stellar atmospheres—the field from which these methods, and gray transfer, originate (Chandrasekhar 1960). With the exception of Monte Carlo methods (Ryan et al. 2015; Ryan & Dolence 2020), these are all especially poorly suited to the diffuse, hot, optically thin corona, especially at small radii near the black hole.
|
[
"Jiang et al.",
"2014b"
] |
[
"Even so, significant progress has been made in the last decade, though the problem is usually made tractable by introducing at least one of the following simplifying assumptions: abandoning general relativity in favor of a pseudo-Newtonian description of the gravitational potential while performing realistic, multi-angle group radiation transport"
] |
[
"Background"
] |
[
[
579,
591
],
[
599,
604
]
] |
[
[
229,
577
]
] |
2022MNRAS.513.5377F__Blum_et_al._2017_Instance_2
|
At each heliocentric distance rh, the activity model (Fulle et al. 2020b) is defined by five analytical equations fixing (i) the gas pressure P(s) depending on the depth s from the nucleus surface (Fig. 1 for the CO2 case), (ii) the gas flux Q from the nucleus surface, (iii) the temperature gradient ∇T at depths of a few cm, (iv) the heat conductivity λs at depths of a few cm below the nucleus surface, and (v) the temperature Ts of the nucleus surface
(3)$$\begin{eqnarray*}
P(s) = P_0 ~f(s) ~\exp \left[{- {T_0 \over {T_s - s ~\nabla T}}}\right]
\end{eqnarray*}$$(4)$$\begin{eqnarray*}
Q = {{14 ~r ~P(R)} \over {3 ~R}} \sqrt{{2 ~m} \over {\pi k_B ~(T_s - R ~\nabla T)}}
\end{eqnarray*}$$(5)$$\begin{eqnarray*}
\nabla T = {\sqrt{\Lambda ~Q ~/ ~\sigma _B} \over {8 ~(T_s - R ~\nabla T) ~R}}
\end{eqnarray*}$$(6)$$\begin{eqnarray*}
\lambda _s = {32 \over 3} ~(T_s - R ~\nabla T)^3 ~\sigma _B ~R
\end{eqnarray*}$$(7)$$\begin{eqnarray*}
(1 - A) ~I_\odot ~\cos \theta ~r_h^{-2} = \epsilon \sigma _B T_s^4 + \lambda _s ~\nabla T + \Lambda ~Q
,
\end{eqnarray*}$$where P0, T0, and Λ values are listed in Table 1, s is the depth from the nucleus surface, $f(s) = 1 - (1 - {s \over R})^4$ for s ≤ R, f(s) = 1 elsewhere, r ≈ 50 nm and R ≈ 5 mm are the radii of the grains of which cometary dust consists (Levasseur-Regourd et al. 2018; Güttler et al. 2019; Mannel et al. 2019) and of the pebbles of which cometary nuclei consist (Blum et al. 2017; Fulle et al. 2020b), m is the mass of the gas molecule, kB is the Boltzmann constant, σB is the Stefan–Boltzmann constant, A is the nucleus Bond albedo (e.g. A = 1.2 per cent measured at 67P; Fornasier et al. 2015), I⊙ is the solar flux at the heliocentric distance of Earth, θ is the solar zenithal angle, and ϵ ≈ 0.9 is the nucleus emissivity. Since the gas originates from the superficial pebbles and is assumed to share the temperature Ts − s ∇T of refractories and ices, the thermal diffusion due to gas convection is negligible with respect to the sublimation sink Λ Q. A nucleus is active if the gas pressure overcomes the tensile strength S bonding dust particles to the nucleus surface (Skorov & Blum 2012), thus defining the activity onset for each ice (Table 2), occurring (i) at rh = 85 au for carbon monoxide (Fulle et al. 2020a); (ii) at rh = 60 au for molecular oxygen; (iii) at rh = 52 au for methane; (iv) at rh = 18 au for ethane; (v) at rh = 13 au for carbon dioxide (dotted line in Fig. 1); and (vi) at rh = 3.8 au for water (Fulle et al. 2020b; Ciarniello et al. 2021). The value R ≈ 5 mm has been constrained by several data collected at comet 67P, by laboratory experiments of dust accretion in conditions expected to occur in the solar protoplanetary disc and by observations of other protoplanetary discs (Blum et al. 2017). Other R-values would not provide the best fit of the 67P water-loss time-evolution (Ciarniello et al. 2021).
|
[
"Blum et al. 2017"
] |
[
"The value R ≈ 5 mm has been constrained by several data collected at comet 67P, by laboratory experiments of dust accretion in conditions expected to occur in the solar protoplanetary disc and by observations of other protoplanetary discs"
] |
[
"Uses"
] |
[
[
2783,
2799
]
] |
[
[
2543,
2781
]
] |
2015ApJ...805..105C__Li_1995_Instance_1
|
As we said in the previous section, the evolution of the magnetic fields in the two domains (the disk and the magnetosphere) are related. It is important to note here that there is an important complication that enters into the matching between the two domains and this has to do with the abrupt transition between the dense turbulent accreting flow and the almost empty ionized (probably also evaporating and outflowing) magnetosphere. This is very similar to the well known transition region in the Sun where, within a very thin layer of about 100 km, the temperature rises by almost two orders of magnitude to about one million degrees, and the matter density drops by about one order of magnitude. The solar transition region is the subject of ongoing intense theoretical and numerical investigations. The astrophysical disk transition region is at least equally complicated since it is not only the base of the disk corona, but also the origin of purported disk winds and outflows (e.g., Ferreira & Pelletier 1995; Li 1995). It is not the aim of the present study to solve in detail for the structure of the flow and magnetic field in this region. In practice, we define a transition region of two θ-grid zones above
. In that region we set
14
This matching allows for magnetic field loops that reach the disk surface zones to escape to the magnetosphere. We also set the field lines in that region in rotation by introducing a magnetospheric poloidal electric field
15
Finally, we slightly modify the induction equation that we solve for the disk material in that region by introducing an extra phenomenological term that mimics various complex physical effects taking place in the surface layers of astrophysical accretion disks such as surface convection, buoyancy of magnetic field loops, and the fact that the surface layers are highly ionized (thus also fully conducting) due to cosmic ray irradiation, namely
16
where,
is the disk magnetic field in the transition layer. The introduction of this transition layer effectively shields the disk interior: (a) it prevents any generated magnetospheric field from “entering” the disk from above (especially in the case of turbulent high η flow conditions), and (b) it “pushes outward” (in the θ direction) at a very small fraction of the speed of light any magnetic field loop that enters it from below. We implemented different fractions (
,
, etc.) and the results were qualitatively similar. As we said above, we are not claiming that we study in detail the disk-magnetosphere transition region. This is the reason we opted for a minimal two θ-zone layer that only serves the purpose of shielding the disk interior from the disk magnetosphere. Note that Parfrey et al. (2015) ignored the important physical significance of a transition layer by directly coupling their magnetosphere to a prescribed distribution of the magnetic field on the surface of the accretion disk without solving for the magnetic field distribution in the disk interior.
|
[
"Li 1995"
] |
[
"The astrophysical disk transition region is at least equally complicated since it is not only the base of the disk corona, but also the origin of purported disk winds and outflows (e.g.,"
] |
[
"Motivation"
] |
[
[
1020,
1027
]
] |
[
[
806,
992
]
] |
2019ApJ...871..257P__Gabuzda_et_al._2018_Instance_1
|
We note that it is unlikely that poloidal magnetic fields are responsible for the observed RMs of M87 because in that case one expects ρ ∝ r0 from B ∝ r−2, which is impossible to explain with the accretion models currently available (Yuan & Narayan 2014). However, there is indication of non-negligible poloidal fields as well as toroidal fields—resulting in helical magnetic fields—in the jet environment of other distant AGNs, which results in transverse RM gradients with no sign changes (e.g., Asada et al. 2002; Zamaninasab et al. 2013; Gómez et al. 2016; Gabuzda et al. 2018, see also Section 3.5.2). The existence of non-negligible poloidal fields was indicated even for the M87 jet at HST-1 from the observed moving knots with both fast and slow velocities, which could be explained by quad relativistic MHD shocks in a helical magnetic field permeating the jet (Nakamura et al. 2010; Nakamura & Meier 2014). In Sections 3.5.2 and 4.1, we explained that poloidal magnetic fields might be very weak at distances ≳5000 rs probed in this study and we concluded that hot accretion flows and winds are more probable to be the Faraday screen than the jet sheath. However, if the jet experiences recollimation, which may lead to formation of standing shocks (e.g., Daly & Marscher 1988; Gómez et al. 1995; Agudo et al. 2001; Mizuno et al. 2015; Martí et al. 2016; Fuentes et al. 2018), then the strength of poloidal fields could be substantially enhanced. Indeed, the width of HST-1 is significantly smaller than expected from the parabolic (conical) width profile inside (outside) the Bondi radius (Asada & Nakamura 2012), which has been explained with a hydrodynamic recollimation shock (e.g., Stawarz et al. 2006; Bromberg & Levinson 2009; Asada & Nakamura 2012). Also, the core of blazars is often identified with a recollimation shock (e.g., Daly & Marscher 1988; Marscher 2008; Cawthorne et al. 2013). This may explain the presence of non-negligible poloidal fields in the sheath of blazar jets and in HST-1, but not in the M87 jet inside the Bondi radius.
|
[
"Gabuzda et al. 2018"
] |
[
"However, there is indication of non-negligible poloidal fields as well as toroidal fields—resulting in helical magnetic fields—in the jet environment of other distant AGNs, which results in transverse RM gradients with no sign changes (e.g.,"
] |
[
"Background"
] |
[
[
561,
580
]
] |
[
[
256,
497
]
] |
2016ApJ...822...15S__Brogaard_et_al._2012_Instance_1
|
However, a significant outstanding problem of using red giants is that modeling their individual frequencies is too time consuming for the analysis of tens of thousands of stars. We therefore rely on using asteroseismic scaling relations,
ν
max
∝
gT
eff
−
1
/
2
and
Δ
ν
∝
ρ
(Brown et al. 1991; Kjeldsen & Bedding 1995), to estimate their radius and mass (and hence age). Here,
ν
max
is the frequency of maximum amplitude and
Δ
ν
the average large frequency separation. These relations assume that the structure of a red giant star is homologous with respect to the Sun. In reality this assumption is not strictly correct and verification of the relations is ongoing. However, the independent high-precision estimates of mass and radius required for this verification are difficult to obtain. For subgiants and dwarfs, the
ν
max
scaling relation has been shown to work well (Bedding 2014) and recently, Coelho et al. (2015) found the proportionality
ν
max
∝
gT
eff
−
1
/
2
to be accurate to within 1.5%. Using Hipparcos parallaxes and/or interferometry, the asteroseismic radii calculated from scaling relations have been found to be accurate to within 5% (Bruntt et al. 2010; Huber et al. 2012; Silva Aguirre et al. 2012). For giants we generally do not have accurate parallaxes, so such studies are awaiting results from Gaia (Perryman 2002). Open clusters have been used to test the scaling relations for giants (Brogaard et al. 2012; Miglio et al. 2012; Sandquist et al. 2013). Miglio et al. (2012) found agreement to within 5% for scaling relation-based radii. Testing of masses is more challenging. For a few cases where such verification have been performed, the scaling relation-based masses seem to be overestimated for giants (Miglio et al. 2012; Frandsen et al. 2013; Epstein et al. 2014). For two lower red giant branch stars (Epstein et al. 2014) find evidence that the mass estimated by using only
Δ
ν
(but with additional
ν
max
independent quantities) is lower compared to using both
Δ
ν
and
ν
max
. Based on this they suggested that a modification to the
ν
max
scaling relation might be required. Theoretical modeling has suggested corrections to the
Δ
ν
scaling relation (Stello et al. 2009; White et al. 2011; Miglio et al. 2013a), but there has been no comprehensive study to verify the corrections. In relation to
ν
max
, Houdek et al. (1999) and Chaplin et al. (2008) had suggested theoretically that
ν
max
coincides with the plateau of the damping rate with frequency. Belkacem et al. (2011) confirmed this for the Sun using SoHO GOLF observations. Balmforth (1992) suggested that this is caused by a resonance between the thermal adjustment time of the superadiabatic boundary layer and the mode frequency, which was also confirmed by the theoretical study of Belkacem et al. (2011; see also Belkacem 2012; Belkacem et al. 2013). However, there is currently no way to accurately predict
ν
max
from theory.
|
[
"Brogaard et al. 2012"
] |
[
"Open clusters have been used to test the scaling relations for giants"
] |
[
"Background"
] |
[
[
1507,
1527
]
] |
[
[
1436,
1505
]
] |
Subsets and Splits
No community queries yet
The top public SQL queries from the community will appear here once available.