Identifier
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9.23k
| Citation Text
list | Functions Text
list | Functions Label
list | Citation Start End
list | Functions Start End
list |
|---|---|---|---|---|---|---|
2017AandA...599A..13Y__Bučík_et_al._(2009)_Instance_1
|
Figure 2 shows solar wind plasma and magnetic field measurements for a CIR that occurred between July 26 and 27, 2003 (days of year 207–208). Following Chotoo et al. (2000), Richardson et al. (1993), we marked four regions in the plot: the slow wind region (S), the compressed slow wind region (S′), the compressed fast wind region (F′), and the fast wind itself (F). Throughout four regions, the mean charge states of iron measured by ACE/SWICS lies around 11+, consistent with typical values in the solar wind (Lepri et al. 2001). The stream interface (S′-F′) is indicated by the vertical line in Fig. 2 and is characterized by a drop of the O7+/O6+ abundance ratio measured with SWICS in the bulk solar wind (Wimmer-Schweingruber et al. 1997, 1999). The leading (S-S′) and trailing edge (F′-F) of the CIR were determined by the total pressure (Jian et al. 2006). Bučík et al. (2009) found that CIR boundaries can be well defined when the total pressure exceeds 50 pPa (indicated by the horizontal dashed line in Fig. 2), which is slightly higher than that in the background solar wind, which typically is 20−30 pPa, according to Jian et al. (2006). The total pressure P was obtained from the sum of plasma and magnetic field pressure, that is, \hbox{$P=n_{\rm p}v^{2}_{\rm th}m+B^2/2\mu_0$}P=npvth2m+B2/2μ0, where np and vth are the proton density and thermal speed, respectively, and B is the magnitude of the magnetic field. Because SOHO has no magnetometer, we used magnetic field data from ACE/MAG (which is also around L1). Comparing plasma parameters (bulk speed, thermal speed, and proton density) measured by PM with those of SWEPAM, we see that the physical conditions at SOHO and ACE were almost the same, and that the time difference between passages of the CIR boundaries is less than ten minutes. The CIR shown in Fig. 2 was bounded by a reverse shock (vertical line separating F′ from F). We clearly see that the suprathermal He++ intensity peaks inside the decelerated and compressed fast-wind region (F′), close to the reverse shock. In contrast, suprathermal particles are very rare in the S and S′ regions. After passage of the reverse shock, suprathermal particles continue to be observable for more than one day. They are commonly believed to be the sunward particles accelerated by the reverse shock far beyond the Earth orbit. In other words, the observer saw the duration of the CIR particle event, which was longer than that of the CIR compression region itself. The background level shown in green was estimated using the method described above. The signal-to-noise ratio (S/N) in the F and F′ regions is higher than 100, confirming that our observations are due to real He++ particles.
|
[
"Bučík et al. (2009)"
] |
[
"found that CIR boundaries can be well defined when the total pressure exceeds 50 pPa (indicated by the horizontal dashed line in Fig. 2), which is slightly higher than that in the background solar wind, which typically is 20−30 pPa, according to Jian et al. (2006)."
] |
[
"Uses"
] |
[
[
866,
885
]
] |
[
[
886,
1151
]
] |
2021MNRAS.506.5468Z__Begeman,_Broeils_&_Sanders_1991_Instance_1
|
Since the free-fall time-scale tff depends on the gravity law, we study how changing the gravity law affects the value of n in the KS law. Our main aim in this paper is to derive for the first time the KS law from a basic description in the framework of Milgromian dynamics (MOND; Milgrom 1983; Famaey & McGaugh 2012). MOND is an alternative approach to a cold dark matter-dominated universe deduced from the flattening of observed rotation curves of spiral galaxies under the assumption of Newtonian dynamics. In MOND, these dynamical discrepancies are addressed by a generalization of Newtonian gravity (for a thorough review, see e.g. Famaey & McGaugh 2012). Within the classical MOND framework, the Newtonian gravitational acceleration gN is replaced in the spherically symmetric case by $g=\sqrt{g_{\rm {N}} a_0}$ when the gravitational acceleration is far smaller than the critical acceleration a0 = 1.2 × 10−10 m s−2 (Begeman, Broeils & Sanders 1991; Gentile, Famaey & de Blok 2011). In less symmetric configurations, the equations of motion are derived from a Lagrangian, yielding standard equations of motion but with a generalized Poisson equation for the gravitational field (Bekenstein & Milgrom 1984; Milgrom 2010). MOND predicted the very tight radial acceleration relation (RAR) between the gravity g implied by disc galaxy rotation curves and the Newtonian gravity gN, resulting from their baryonic distribution (Lelli et al. 2017; Li et al. 2018). The RAR is also evident in stacked galaxy–galaxy weak lensing measurements that probe out to larger radii (Milgrom 2013; Brouwer et al. 2021). The external field effect (EFE) predicted by MOND (Bekenstein & Milgrom 1984) and required for consistency with data on Solar neighbourhood wide binaries (Banik & Zhao 2018c; Pittordis & Sutherland 2019) has recently been confirmed at high significance by comparing galaxies in isolated and more crowded environments (Haghi et al. 2016; Chae et al. 2020). Detailed numerical simulations of disc galaxy secular evolution in MOND have recently been conducted for the case of M33 (Banik et al. 2020) and for a Milky Way or M31-like surface density (Roshan et al. 2021), while star formation has also been explored with high-resolution simulations (Renaud, Famaey & Kroupa 2016). The possible cosmological context has been explored in e.g. Haslbauer, Banik & Kroupa (2020) and Asencio, Banik & Kroupa (2021). The MOND corrections to Newtonian gravity might be capturing effects of the quantum vacuum (Milgrom 1999; Smolin 2017; Verlinde 2017; Senay, Mohammadi Sabet & Moradpour 2021).
|
[
"Begeman, Broeils & Sanders 1991"
] |
[
"Within the classical MOND framework, the Newtonian gravitational acceleration gN is replaced in the spherically symmetric case by $g=\\sqrt{g_{\\rm {N}} a_0}$ when the gravitational acceleration is far smaller than the critical acceleration a0 = 1.2 × 10−10 m s−2"
] |
[
"Background"
] |
[
[
925,
956
]
] |
[
[
662,
923
]
] |
2020AandA...644A.108V__Birrer_&_Amara_2018_Instance_1
|
The most commonly used technique to create a mock lens system from simulated galaxies is to extract a mass map from a particle-based simulation and use it to calculate lensing quantities (i.e. lensing potential and its first and second derivatives) needed to emulate the gravitationally lensed images. For this purpose, galaxies from high resolution hydrodynamical simulations, including, for example, EAGLE (Evolution and Assembly of GaLaxies and their Environments, Schaye et al. 2015; Crain et al. 2015) or Illustris (Vogelsberger et al. 2014b,a), have been widely used. Different types of software, such as lenstronomy (Birrer & Amara 2018) and GLAMER (Gravitational Lensing Simulations with Adaptive Mesh Refinement, Metcalf & Petkova 2014), can handle the inference of lensing quantities from mass maps using fast Fourier transform convolution. Fast Fourier is a commonly used technique to speed up the calculation of lensing quantities, which imply computationally expansive numerical integration but it remains a demanding procedure (Metcalf & Petkova 2014; Plazas 2020). One could wonder what mass map resolution should be used and what size of map is relevant to be sufficiently precise in the mock creation while minimising the computational time. One generally considers that a strongly lensed system is determined by the projected mass inner to the lensed images. This would suggest that a region extending over a few Einstein radii (θE) is sufficient for the simulations. However, this consists in effectively ignoring any source of shear, and/or perturbations caused by substructures and/or anisotropy in the mass distributions. Moreover, depending on the symmetry of the problem, cutting the mass distribution at a given radius not only automatically removes the mass beyond that radius, but it may also introduce numerical artefacts that could wrongly be attributed to properties of the examined lens mass distribution. In this paper we focus on this latter point and quantify the impact of the shape (and size) of the integration domain on the lensing quantities inference.
|
[
"Birrer & Amara 2018"
] |
[
"Different types of software, such as lenstronomy",
"can handle the inference of lensing quantities from mass maps using fast Fourier transform convolution."
] |
[
"Background",
"Background"
] |
[
[
624,
643
]
] |
[
[
574,
622
],
[
747,
850
]
] |
2021AandA...648A.120R__Shimajiri_et_al._2014_Instance_1
|
We derived the mean values of the 13CO/C18O abundance ratio for the three star-forming regions, obtaining values of 13CO/C18O = 10.9 ± 7.5 for Taurus, 13CO/C18O = 17.3 ± 11.5 for Perseus, and 13CO/C18O = 23.7 ± 10.4 for Orion. The expected value is 13CO/C18O = 7.5−9.8, assuming 12C/13C = 57−67 and 16O/18O = 500−600 (Gerin et al. 2015; Langer & Penzias 1990; Wilson & Rood 1994). As mentioned in Sect. 5.7, the 13CO/C18O abundance ratio increases in regions of enhanced UV field (Shimajiri et al. 2014; Ishii et al. 2019; Areal et al. 2018). This variation is interpreted in terms of the selective photodissociation and isotopic fractionation (see Bron et al. 2018; Fuente et al. 2019). This effect is produced because the more abundant CO isotopolog shields itself from the effect of UV photons more efficiently than less abundant isotopologs (Stark et al. 2014; Visser et al. 2009). Our data confirm this trend in Orion, although the observed positions are located at a distance >1 pc from the ionized nebula M 42, which shows that the whole cloud is illuminated by a strong UV field emitted by young massive stars in the Trapezium cluster (Pabst et al. 2019). In contrast to X(C18O), which presents uniform abundance in the three regions, the mean 13CO abundance is a factor of approximately two higher in Orion than in Taurus. One interesting question pertains to whether the observed variation in the 13CO abundance is related to a variation in the 12CO/13CO ratio (Roueff et al. 2015; Colzi et al. 2020) or is revealing variations for a higher 12CO abundance in Orion. Because of the higher gas temperature and incident UV field in Orion, selective photodissociation would work in the direction of increasing N(12CO)/N(13CO) in Orion with respect to Taurus. Therefore, a higher 12CO abundance in Orion stands as the most likely explanation. Our results are based on significant approximations (one single phase, gas-dust thermalization), and therefore a detailed multi-transition study of these compounds is required to confirm this result.
|
[
"Shimajiri et al. 2014"
] |
[
"As mentioned in Sect. 5.7, the 13CO/C18O abundance ratio increases in regions of enhanced UV field"
] |
[
"Uses"
] |
[
[
481,
502
]
] |
[
[
381,
479
]
] |
2021AandA...653A.111R__Jones_et_al._(2021)_Instance_2
|
As done by Le Fèvre et al. (2020), we visually inspect the ancillary data, the intensity maps, the velocity and velocity dispersion fields presented in Sect. 3.1 to search for the presence of multiple components or disturbed morphology near the position of the targets. The channel maps, the spectra and the PVDs are checked together searching for consistent emission features. By taking into account the results of the initial qualitative classification by Le Fèvre et al. (2020) and of the more recent quantitative analysis of a subsample of the ALPINE targets by Jones et al. (2021), we proceed with a more in-depth characterization of the [CII]-detected galaxies aimed at obtaining a robust merger fraction at z ∼ 5. Adopting the same criteria described in Sect. 2 to differentiate the targets and considering the S/N of the minor merger component as described in Sect. 3.1, we find a slightly lower fraction (∼31%, 23 out of 75 [CII]-detected sources) of mergers7 if compared to the 40% found by Le Fèvre et al. (2020), with 12, 20 and 7% of the sample made by rotating, extended and compact dispersion dominated sources, respectively. To be more conservative in the classification of the galaxies (especially for obtaining a more robust merger statistics), we define the remaining 30% of the sample as ‘uncertain’, a new category that includes the weak galaxies (as described in Le Fèvre et al. 2020) and also objects that, by visual inspection, present features that are intermediate to those of various classes. This category is similar to the ‘uncertain’ (UNC) class introduced in Jones et al. (2021) that, based on the results of the 3DBarolo fits, contains sources they are unable to classify because of the low S/N and/or spectral resolution, or contrasting evidence in their classification criteria. Although the uncertain category is populated by a significantly larger fraction of sources with respect to the weak class (∼16%) in Le Fèvre et al. (2020), we recover the same qualitative morpho-kinematic distribution of the previous analysis, confirming the high fraction of rotators and mergers at these early epochs.
|
[
"Jones et al. (2021)"
] |
[
"This category is similar to the ‘uncertain’ (UNC) class introduced in",
"that, based on the results of the 3DBarolo fits, contains sources they are unable to classify because of the low S/N and/or spectral resolution, or contrasting evidence in their classification criteria."
] |
[
"Similarities",
"Similarities"
] |
[
[
1590,
1609
]
] |
[
[
1520,
1589
],
[
1610,
1812
]
] |
2020MNRAS.499.4394M__Bate_et_al._2014_Instance_1
|
For these four remaining FHSC candidates (L1451-mm, MC35-mm, SM1N, and Oph A-N6) that have been observed at intermediate scales (few 100 au to few 1000 au) a final confirmation of their true evolutionary state requires higher resolution observations. For L1451-mm, the compact outflow needs to be resolved to investigate its morphology and kinematics, as a higher velocity component (an indication of protostellar nature) could be revealed by observations with a beam smaller than 100 au, similar to the case of B1b-N (Hirano 2019). An additional goal of high-resolution observations for L1451-mm and the remaining youngest candidates should be to investigate the temperature and density profiles of the envelope at scales from few au to 100 au. This is because simulations show that the temperature remains lower than ∼30 K even at several tens of au up to 100 au from the centre (Bate et al. 2014; Tomida et al. 2015; Hincelin et al. 2016; Young et al. 2019) during the FHSC stage. On the other hand, Class 0 sources show temperatures of 20–30 K or higher at scales of several 100 au (sufficient for thermal evaporation of CO) that results in the inner envelope and disc being easily detected using C18O observations (Yen et al. 2015, 2017; Stephens et al. 2018). This holds even in very low luminosity objects, for which the unexpected large extent of C18O is interpreted as evidence of a previous burst of accretion (Frimann et al. 2017; Hsieh et al. 2018). As for the density profile, simulations of the FHSC stage show a flat inner region, corresponding to the FHSC structure and extending up to ∼10 au (Tomida et al. 2013; Bate et al. 2014). For a protostar, on the other hand, the density profile should increase towards the central ∼1 au region (Young et al. 2019). Observations of the continuum emission with a resolution better than a few tens of au are likely required to model the emission and provide a density and temperature profile that can probe the relevant scales. Additional line observations with a similar resolution can also help to further distinguish between the different models. We note that, as pointed out in Young et al. (2019), distinguishing a dense core with only an FHSC and one that has recently formed a protostar but in which the FHSC structure is still present is likely not possible, even with high-resolution observations. Given the optically thick nature of the FHSC core, it is difficult to probe the physical properties within the FHSC structure. Despite this, finding a source with density and temperature profiles as well as with outflow properties consistent with the theoretical predictions will provide convincing evidence in support of a bona fide FHSC.
|
[
"Bate et al. 2014"
] |
[
"An additional goal of high-resolution observations for L1451-mm and the remaining youngest candidates should be to investigate the temperature and density profiles of the envelope at scales from few au to 100 au. This is because simulations show that the temperature remains lower than ∼30 K even at several tens of au up to 100 au from the centre"
] |
[
"Future Work"
] |
[
[
882,
898
]
] |
[
[
533,
880
]
] |
2015MNRAS.451.4290S__Torrey_et_al._2014_Instance_1
|
Hydrodynamical simulations of evolving galaxies allow us to calibrate these diagnostics by measuring their observability given a set of formation scenarios and physical processes (e.g. Jonsson et al. 2006; Rocha et al. 2007; Lotz et al. 2008a; Bush et al. 2010; Narayanan et al. 2010; Hayward et al. 2013; Snyder et al. 2013; Lanz et al. 2014). The quality and breadth of these experiments are limited by the availability of computational resources and the fidelity of models for galaxy physics such as star formation, supernovae, and the interstellar medium (ISM). It has only recently become widespread to model the formation of galaxies ab initio (e.g. Governato et al. 2004; Agertz, Teyssier & Moore 2011; Guedes et al. 2011; Marinacci, Pakmor & Springel 2013; Ceverino et al. 2014), and the realism continues to improve (Stinson et al. 2012; Hopkins et al. 2014; Torrey et al. 2014), albeit with still widely varying physics models (e.g. Scannapieco et al. 2012; Kim et al. 2014). Prior to these advances, studies were limited to small numbers of isolated galaxies or mergers to inform common diagnostics of galaxy evolution, an approach with a significant limitation: they do not fully account for cosmological context, such as gas accretion and the breadth of assembly histories. In addition to mergers, models of high-redshift galaxy formation (e.g. Dekel, Sari & Ceverino 2009; Dekel et al. 2013) have recently appreciated the tight coupling between gas accretion and disc evolution (e.g. Cacciato, Dekel & Genel 2012; Danovich et al. 2012; Dekel & Krumholz 2013), as well as bulge and super-massive black hole (SMBH) growth mediated by turbulent motions or violent disc instability (e.g. Bournaud et al. 2011; Porter et al. 2014) and the evolution of giant clumps (Dekel & Burkert 2013). These important processes likely complicate interpretation of a given observation, and recent studies of galaxy morphology have begun to exploit simulations including them (e.g. Scannapieco et al. 2010; Pedrosa, Tissera & De Rossi 2014).
|
[
"Torrey et al. 2014"
] |
[
"The quality and breadth of these experiments are limited by the availability of computational resources and the fidelity of models for galaxy physics such as",
"It has only recently become widespread to model the formation of galaxies ab initio",
"and the realism continues to improve"
] |
[
"Motivation",
"Motivation",
"Motivation"
] |
[
[
868,
886
]
] |
[
[
345,
502
],
[
566,
649
],
[
788,
824
]
] |
2021MNRAS.508.2583Z___2016_Instance_1
|
Located in the star-forming region ρ-Ophiuchi, inside the dark cloud L1689N and at a distance of 141 pc (Dzib et al. 2018), IRAS16293−2422 is a well-studied Young Stellar Object (YSO) classified as a Class 0 source with less than 104 yr (Andre, Ward-Thompson & Barsony 1993), and represents one of the very early stages of low-mass star formation. It was the first source identified as a hot corino (Blake et al. 1994; van Dishoeck et al. 1995) based on the detection of Complex Organic Molecules (COMs) in the source, which was later supported by follow-up studies (Ceccarelli et al. 1998, 2000; Schöier et al. 2002; Crimier et al. 2010; Jørgensen et al. 2011, 2016; Pineda et al. 2012; Oya et al. 2016; Jacobsen et al. 2018; van der Wiel et al. 2019). Higher resolution observations revealed that IRAS16293−2422 is in fact a triple system, composed of sources A1 and A2, separated by 54 au from each other (Maureira et al. 2020) and source B, 738 au (5 arcse; Wootten 1989) away from source A. Due to this larger separation, tidal truncation between the three protostars is discarded and therefore source B is considered to have evolved as an isolated source (Rodríguez et al. 2005). It was initially proposed to be either an evolved T Tauri star (Stark et al. 2004; Takakuwa et al. 2007) or a very young object (Chandler et al. 2005), however, Chandler et al. (2005) suggested that source B has large-scale infalls based on SO line emission. Pineda et al. (2012) confirmed the infall of an inner envelope, with mass accretion rates of 4.5 × 10−5 M⊙yr−1, based on ALMA detections of inverse P-Cygni profiles in CH3OCHO-E, CH3OCHO-E-A and H2CCO, ruling out the possibility of it being a T Tauri star. The interpretations of infall from these profiles was also suggested by Jørgensen et al. (2012) and Zapata et al. (2013). Unlike the A1 and A2 protostars, source B has not shown clear signs of outflow launching, explained by the lack of free–free emission at low frequencies (Chandler et al. 2005; Rodríguez et al. 2005; Loinard et al. 2007; Rao et al. 2009; Liu et al. 2018; Hernández-Gómez et al. 2019b) and also based on molecular lines (Loinard et al. 2002; van der Wiel et al. 2019).
|
[
"Jørgensen et al.",
"2016"
] |
[
"It was the first source identified as a hot corino",
"based on the detection of Complex Organic Molecules (COMs) in the source, which was later supported by follow-up studies"
] |
[
"Background",
"Background"
] |
[
[
639,
655
],
[
662,
666
]
] |
[
[
348,
398
],
[
445,
565
]
] |
2021AandA...647A.140C__Gianninas_et_al._2016_Instance_2
|
In recent years, numerous low-mass and ELM WDs have been detected in the context of relevant surveys, such as the SDSS, ELM, SPY and WASP (see, e.g., Koester et al. 2009; Brown et al. 2010, 2016, 2020; Kilic et al. 2011, 2012; Gianninas et al. 2015; Kosakowski et al. 2020). The discovery of their probable precursors, namely, the so-called low-mass pre-WDs, has triggered an interest in these types of objects because of the possibility of studying the evolution of the progenitors that lead to the WD phase. Even more interestingly, the detection of multi-periodic brightness variations in low-mass WDs (Hermes et al. 2012, 2013a,b; Kilic et al. 2015, 2018; Bell et al. 2017, 2018; Pelisoli et al. 2018), and low-mass pre-WDs (Maxted et al. 2013, 2014; Gianninas et al. 2016; Wang et al. 2020) has brought about new classes of pulsating stars known as ELMVs and pre-ELMVs, respectively (ELM and pre-ELM variables, respectively). It has allowed for the study of their stellar interiors using the tools of asteroseismology, similarly to the case of other pulsating WDs such as ZZ Ceti stars or DAVs –pulsating WDs with H-rich atmospheres – and V777 Her or DBVs – pulsating WDs with He-rich atmospheres (Winget & Kepler 2008; Fontaine & Brassard 2008; Althaus et al. 2010; Córsico et al. 2019). The pulsations observed in ELMVs are compatible with global gravity (g)-mode pulsations. In the case of pulsating ELM WDs, the pulsations have large amplitudes mainly at the core regions (Steinfadt et al. 2010; Córsico et al. 2012; Córsico & Althaus 2014a), allowing for the study of their core chemical structure. According to nonadiabatic computations (Córsico et al. 2012; Van Grootel et al. 2013; Córsico & Althaus 2016), these modes are probably excited by the κ − γ (Unno et al. 1989) mechanism acting at the H-ionization zone. In the case of pre-ELMVs, the nonadiabatic stability computations for radial (Jeffery & Saio 2013) and nonradial p- and g-mode pulsations (Córsico et al. 2016; Gianninas et al. 2016; Istrate et al. 2016b) revealed that the excitation is probably due to the κ − γ mechanism, acting mainly in the zone of the second partial ionization of He, with a weaker contribution from the region of the first partial ionization of He and the partial ionization of H. The presence of He in the driving zone is crucial to having the modes destabilized by the κ − γ mechanism (Córsico & Althaus 2016; Istrate et al. 2016b).
|
[
"Gianninas et al. 2016"
] |
[
"In the case of pre-ELMVs,",
"and nonradial p- and g-mode pulsations",
"revealed that the excitation is probably due to the κ − γ mechanism, acting mainly in the zone of the second partial ionization of He, with a weaker contribution from the region of the first partial ionization of He and the partial ionization of H."
] |
[
"Background",
"Background",
"Background"
] |
[
[
1988,
2009
]
] |
[
[
1828,
1853
],
[
1927,
1965
],
[
2033,
2281
]
] |
2018MNRAS.477.3520L__Abolfathi_et_al._2018_Instance_1
|
Over time, the data releases have treated the Balmer line regions in different ways. The presence of the artificial curvature was first reported by Busca et al. (2013) in the context of the DR9 data release. To minimize this effect, a different scheme was used in DR12 (Alam et al. 2015, see their table 2) by using a linear function (instead of an iterative b-spline procedure) to interpolate the flux over the masked regions. Surprisingly, we observe that this data reduction change was only applied to the Balmer β, γ, and δ lines but not applied to the Balmer α line. For the latter one, the original iterative spline interpolation scheme has been used all along. As a result, an absorption-like feature at the location of Balmer α is found in SDSS data releases 9 up to now, i.e. the latest data release 14 (Abolfathi et al. 2018). To illustrate this, we show examples of calibration vectors for SDSS BOSS DR9 (Ahn et al. 2012; Dawson et al. 2013), DR12 (Alam et al. 2015), eBOSS DR14 (Dawson et al. 2016; Abolfathi et al. 2018) data release as well as calibration vectors for the MaNGA survey (Bundy et al. 2015) DR14 data release (Abolfathi et al. 2018) in Fig. 6. In the upper panel, we show small-scale features in calibration vectors from the same plate (number 3647) processed by the DR9 (blue), DR12 (green), and DR14 (red) pipelines. The spectra are obtained by normalizing the large-scale features in calibration vectors with a cubic b-spline with break points separated by 50 Å. The three spectra are mostly identical while the smooth curving features at the wavelengths of Balmer series in DR9 disappear in DR12 and DR14. From DR12, the pipeline corrects the features with linear interpolation across wavelength regions with Balmer β, γ, and δ lines (Alam et al. 2015). One can also observe the change in the median residual spectra of DR9 and DR12 in Fig. 1. However, the Hα feature remains uncorrected from DR9 to DR14. We also show a DR7 calibration vector (black) in which most of the wiggles are absent. As pointed previously, this is due to the fact that the DR7 pipeline interpolates the calibration vectors using an effective scale larger than that used in subsequent data releases.
|
[
"Abolfathi et al. 2018"
] |
[
"Surprisingly, we observe that this data reduction change was only applied to the Balmer β, γ, and δ lines but not applied to the Balmer α line. For the latter one, the original iterative spline interpolation scheme has been used all along. As a result, an absorption-like feature at the location of Balmer α is found in SDSS data releases 9 up to now, i.e. the latest data release 14"
] |
[
"Compare/Contrast"
] |
[
[
813,
834
]
] |
[
[
428,
811
]
] |
2018ApJ...869..121M__Manfroid_et_al._2009_Instance_1
|
An example of a low [14N/15N] value is potentially interesting for understanding the origins of material in our own solar system. The observed value in N2 of ∼200 is lower than values measured for both atmospheres of Earth and Venus (∼270; Junk & Svec 1958; Hoffman et al. 1979) and inferred for the protosolar nebula (∼440; Owen et al. 2001; Fouchet et al. 2004; Meibom et al. 2007; Marty et al. 2010). It is more similar to the extremely low values of [14N/15N] found in primitive solar system material, like cometary CN and HCN (∼140; Arpigny et al. 2003; Bockelée-Morvan et al. 2008; Manfroid et al. 2009) and in organic material in chondritic meteorites (∼50–150; Briani et al. 2009; Bonal et al. 2010), and that recently observed in HCN in a sample of T Tauri and Herbig disks (Guzmán et al. 2017). Low [14N/15N] values are seen to be correlated with high D abundances in carbonaceous chondrites, “cluster” interplanetary dust particles, and comets Wild2 and Hale-Bopp (Messenger 2000; Busemann et al. 2006; Floss et al. 2006; Aléon 2010), an enrichment that occurred either in the protosolar disk or in the precursor cloud core. Recent ISM modeling suggests that this correlation may not occur in the core stage, as a correlation between D and 15N is not expected (due to, e.g., sensitive dependence of chemistry on ortho-to-para variations in collisional partners like H2; Wirström et al. 2012; De Simone et al. 2018). This is supported by observations showing a hint of anticorrelation between the abundances of D and 15N using N2H+ in a sample of massive prestellar cores (Fontani et al. 2015b). However, the potential presence of simultaneously high 15N and D in N2 suggests that it could still be possible for this to occur in an individual core, allowing for an interstellar origin for this solar system feature. As the solar system is suggested to have formed in a rich stellar cluster experiencing a nearby supernova (Adams 2010; Dukes & Krumholz 2012; Pfalzner 2013), the environment of a massive protocluster like Sgr B2 may actually be quite relevant for understanding the neighborhood of the protosolar nebula.
|
[
"Manfroid et al. 2009"
] |
[
"It is more similar to the extremely low values of [14N/15N] found in primitive solar system material, like cometary CN and HCN"
] |
[
"Similarities"
] |
[
[
588,
608
]
] |
[
[
404,
530
]
] |
2020AandA...639A..88C__Chatzistergos_et_al._2019b_Instance_1
|
To overcome these limitations, in our previous paper (Chatzistergos et al. 2018b, Paper I, hereafter) we introduced a novel approach to process the historical and modern Ca II K observations, to perform their photometric calibration, to compensate for the intensity centre-to-limb variation (CLV, hereafter), and to account for various artefacts. By using synthetic data, we also showed that our method can perform the photometric calibration and account for image artefacts with higher accuracy than other methods presented in the literature. More importantly, we showed that, as long as the archives are consistent with each other, for example, they are centred at the same wavelength and employing the same bandwidth for the observations, the method can be used to derive accurate information on the evolution of plage areas without the need of any adjustments in the processing of the various archives (Chatzistergos et al. 2019b, Paper II, hereafter). In Paper II, we applied our method to 85 972 images from 9 Ca II K archives to derive plage areas and produce the first composite of plage areas over the entire 20th century. In particular, we analysed the Ca II K archives from the Arcetri, Kodaikanal (8-bit digitisation), McMath-Hulbert, Meudon, Mitaka, Mt Wilson, Rome/PSPT, Schauinsland, and Wendelstein sites. Five out of the nine analysed archives were amongst the most studied and prominent ones, while the remaining archives were from less studied data sources. There are, however, many other Ca II K archives that are available and still remain largely unexplored. These archives harbour the potential to fill gaps in the available plage series as well as to address inconsistencies among the various archives and within individual archives (e.g. change in data quality, or in the measuring instrument with time). Moreover, since the work presented in Paper II, more data from various historical and modern archives became available in digital form. In particular, historical data that have been made available in the meantime include those from the latest 16-bit digitisation of the Kodaikanal archive, Catania, Coimbra, Kenwood, Kharkiv, Kyoto, Manila, Rome, Sacramento Peak, and Yerkes observatories, as well as additional data from the Meudon and Mt Wilson archives. In this light, here we present results from the most comprehensive analysis to date of historical and modern Ca II K observations taken between 1892 and 2019 from 43 different datasets for the purposes of producing a composite plage area series.
|
[
"Chatzistergos et al. 2019b"
] |
[
"More importantly, we showed that, as long as the archives are consistent with each other, for example, they are centred at the same wavelength and employing the same bandwidth for the observations, the method can be used to derive accurate information on the evolution of plage areas without the need of any adjustments in the processing of the various archives",
"Paper II, hereafter)."
] |
[
"Background",
"Background"
] |
[
[
907,
933
]
] |
[
[
544,
905
],
[
935,
956
]
] |
2022ApJ...931...70B__Gabrielse_et_al._2012_Instance_1
|
RFs can propagate from the magnetotail to Earth over a long distance more than 10 R
E together with BBFs behind them (Runov et al. 2009; Cao et al. 2010). Studies have suggested that RFs are crucial regions for particle acceleration, pitch-angle evolution, wave–particle interactions, and electromagnetic energy conversion during their Earthward propagation. For instance, rapid increases in energy fluxes of electrons and ions from tens to hundreds of keV are a typical feature of RF events (Khotyaintsev et al. 2011; Liu et al. 2013, 2018c, 2021a, 2022b; Zhou et al. 2018; Liu & Fu 2019; Gabrielse et al. 2021), pitch-angle distribution of suprathermal electrons can evolve dramatically around RFs (Runov et al. 2013; Liu et al. 2020), strong particle and wave activity can occur in the vicinity of RFs (Ono et al. 2009; Zhou et al. 2009, 2014; Fu et al. 2014; Breuillard et al. 2016; Greco et al. 2017; Yang et al. 2017), and RFs are associated with energy conversion from electromagnetic fields to particles (Sitnov et al. 2009; Huang et al. 2015; Khotyaintsev et al. 2017; Liu et al. 2018a, 2022a). The energetic plasma in the vicinity of RFs plays a key role in connecting the magnetotail with the inner magnetosphere because they carry a large amount of energy and can be injected into the inner magnetosphere to affect the ring current and radiation belt (Gabrielse et al. 2012; Duan et al. 2014; Turner et al. 2014). Possible mechanisms responsible for the energization of particles around RFs have been widely investigated based on both spacecraft observations and numerical simulations during the past decade. The strong convection electric field induced by the strong magnetic field gradient of RFs provides significant adiabatic acceleration of the ambient particles (Birn et al. 2004, 2013, 2015; Gabrielse et al. 2012, 2014, 2016; Ganushkina et al. 2013; Liu et al. 2016; Turner et al. 2016). Nonadiabatic effects, caused by particle reflection ahead of the RFs (Zhou et al. 2018), resonance with RFs (Ukhorskiy et al. 2013, 2017), and scattering by wave emissions (Zhou et al. 2009; Greco et al. 2017), are also significant for particle energization. These above studies usually assumed that the RF surface has a planar boundary at a typical thickness comparable to the ion gyroradius and below (Nakamura et al. 2002; Sergeev et al. 2009; Zhou et al. 2009; Schmid et al. 2011; Liu et al. 2013; Vapirev et al. 2013). Divin et al. (2015b) revealed that the RF surface is unstable to instabilities ranging from electron scales to ion scales. Simulation studies found that RFs can be unstable to interchange instability and that finger-like structures on ion–electron hybrid scales can develop at the RF (Vapirev et al. 2013). Such finger-like structures are found to play a role in modulating the electron acceleration process (Wu et al. 2018). Bai et al. (2022) also reported significant ion trapping acceleration at the RF with ion-scale ripples. Unlike these surface structures with ion or ion–electron hybrid scales, Liu et al. (2018b) recently reported that the RF layer has electron-scale density gradients, currents, and electric fields, based on the MMS mission, which consists of four spacecraft separated by 30 km. Such electron-scale ripple structure can be generated by lower hybrid drift instability (Divin et al. 2015b; Pan et al. 2018). Liu et al. (2021c) presented a detailed investigation of energy flux densities at two RFs with/without the electron-scale surface ripples and indicated that surface ripples may play an important role in the particle dynamics. But how such electron-scale RF structure impacts the electron energization and transport still remains unknown. In this paper, with the aid of observation-based test-particle simulation, we aim to investigate in detail the effect of the front surface ripples on the local electron dynamics.
|
[
"Gabrielse et al. 2012"
] |
[
"The energetic plasma in the vicinity of RFs plays a key role in connecting the magnetotail with the inner magnetosphere because they carry a large amount of energy and can be injected into the inner magnetosphere to affect the ring current and radiation belt"
] |
[
"Motivation"
] |
[
[
1364,
1385
]
] |
[
[
1104,
1362
]
] |
2022MNRAS.511.1714T__Peñarrubia_et_al._2005_Instance_1
|
While these similar characteristics suggest that the OA and C could have a relationship, the nature of this relationship is not clear. Kawata, Thom & Gibson (2003) used numerical simulations to investigate whether Complex C could have been produced by the passage of a satellite galaxy through the Milky Way disc, and while they found that such an event could produce a structure with the general properties of Complex C and the OA, they did not favour this hypothesis because they could not identify compelling evidence of the putative satellite. However, subsequently several stellar streams and structures have been discovered in this region of the outer Galaxy, and the progenitor(s) of these stellar structures could potentially also explain the origin of Complex C and the OA in a scenario like the one investigated by Kawata et al. (2003). The most well-known stellar stream in this region is the Monoceros Ring (e.g. Newberg et al. 2002; Peñarrubia et al. 2005), but surveys such as the Sloan Digital Sky Survey (SDSS; York et al. 2000) and the Gaia Survey (Gaia Collaboration 2016) have revealed additional stellar streams in this part of the Galaxy. Table 1 summarizes some properties of stellar streams in this direction including Monoceros ‘North’, the Anticentre Stream, Hríd, Gaia 9, and GD1. It is intriguing to note that some of the metallicities and locations of these stellar streams are similar to the published metallicities of Complex C and the OA, but as we will see in this paper, the metallicities of these HVCs are not securely measured – there are degeneracies in the ionization models used to derive metallicities, and it is possible that these HVCs have significantly higher metallicities than the published abundances. It is also worth noting that some of these stellar streams, including the northern Monoceros Ring, Hríd, and GD1, have similar velocities to the HVCs. The Anticentre Stream, on the other hand, exhibits significantly different kinematics with mean velocities of $\it {v}_{\rm LSR} = +48$ and +78 km s−1 in two directions studied by Grillmair, Carlin & Majewski (2008). However, spatially overlapping streams with different kinematics could lead to a roiled, turbulent region, which in turn could make precipitation more likely (Voit 2021). To more visually show the spatial correspondence of the stellar structures and the gas clouds, Fig. 1 shows a schematic map of the projected locations of the streams and clouds from Table 1. Note that Fig. 1 is centred and zoomed in on a set of active galactic nuclei employed in this paper (see below) and does not show the full extent of the HVCs and streams.
|
[
"Peñarrubia et al. 2005"
] |
[
"The most well-known stellar stream in this region is the Monoceros Ring (e.g."
] |
[
"Background"
] |
[
[
946,
968
]
] |
[
[
847,
924
]
] |
2021MNRAS.506.1045M__Marshall_et_al._2013_Instance_1
|
Discovered in 1977 from its bright H α emission (Stephenson & Sanduleak 1977), SS433’s defining characteristics are undoubtedly the helical motion of highly collimated jets of plasma launched from its innermost regions, and mass-loaded, non-polar outflows (Fabian & Rees 1979; Margon et al. 1979) which together inflate the surrounding W50 supernova remnant. Knots in SS433’s jet can be resolved at radio frequencies using very long baseline interferometry (VLBI) and indicate the presence of highly relativistic electrons (Vermeulen et al. 1987), while the baryon content is revealed by emission lines ranging from H and He lines in the optical through to highly ionized Fe lines in the X-rays (Kotani et al. 1994; Marshall et al. 2013). The Doppler shifts of the lines indicate precession of the accreting system with a period of ≈ 162 d, also seen in optical (He ii) emission lines originating from the non-polar wind (Fabrika 1997). Both the jets and winds carry a large kinetic luminosity (>1038 erg s−1, e.g. Marshall et al. 2002), which requires extraction of energy via accretion on to a compact object. While the nature of the compact object in SS433 remains somewhat unknown (although dynamical arguments suggest the presence of a black hole – Blundell, Bowler & Schmidtobreick 2008), the rate of mass transfer from the companion star, as inferred from the IR excess (Shkovskii 1981; Fuchs et al. 2006), is thought to be ∼1 × 10−4 M⊙ yr−1, orders of magnitude in excess of the Eddington limit for any plausible stellar remnant (>300 times the Eddington mass accretion rate for a typical stellar mass black hole of around 10 M⊙). Classical theory and radiation magnetohydrodynamic (RMHD) simulations agree that such ‘super-critical’ rates of accretion will lead to a radiatively supported, large scale height (H/R ≈ 1, where H is the height of the disc at distance R from the compact object) accretion disc with powerful winds launched from the surface at mildly relativistic speeds (Shakura & Sunyaev 1973; Poutanen et al. 2007; Ohsuga & Mineshige. 2011; Takeuchi et al. 2013; Jiang et al. 2014; Sadowski et al. 2014).
|
[
"Marshall et al. 2013"
] |
[
"while the baryon content is revealed by emission lines ranging from H and He lines in the optical through to highly ionized Fe lines in the X-rays"
] |
[
"Background"
] |
[
[
716,
736
]
] |
[
[
548,
694
]
] |
2020MNRAS.498..385J__Chomiuk_&_Povich_2011_Instance_1
|
In Fig. 9, we show the total galactic star formation rate as a function of simulation time t for each isolated disc galaxy. Following the initial vertical collapse of the disc and the subsequent star formation ‘burst’ from t ∼ 30 Myr to t ∼ 250 Myr, the SFR settles down to a rate of ∼2–4 M⊙ yr−1. We make absolutely sure to consider each isolated disc in its equilibrium state by examining the cloud population during a later time interval, between t = 600 Myr and t = 1 Gyr (grey-shaded region). Over this period, the SFR declines only gradually, by a total of around 0.5 M⊙ yr−1. These values are consistent with the current observed SFR in the Milky Way (Murray & Rahman 2010; Robitaille & Whitney 2010; Chomiuk & Povich 2011; Licquia & Newman 2015). We may also consider the resolved star-forming behaviour on scales of 750 pc, as studied in nearby galaxies by Bigiel et al. (2008). In Fig. 10, we display the star formation rate surface density as a function of gas surface density for each of our simulated galaxies, at a simulation time of 600 Myr. The top row shows the 2D projection maps of the CO-bright molecular gas column density $\Sigma _{{\rm H}_2}$. These are computed via the total gas column density in equation (35) and degraded using a Gaussian filter of FWHM = 750 pc. The corresponding projections of the star formation rate surface densities ΣSFR are displayed in the central row. Details for the production of all maps are given in Appendix A. In the bottom row, the values of $\Sigma _{{\rm H}_2}$ and ΣSFR in each pixel of the spatially degraded projection maps are compiled to produce a single histogram. The loci of our simulation data fall close to the observed star formation relations obtained by Bigiel et al. (2008), denoted by the orange contours, though with a population of points at lower densities and star formation rates than are reached by the observations. These points arise because we consider all CO-emitting gas down to a molecular hydrogen surface density of $\Sigma _{{\rm H}_2} = 10^{-3.5} \: {\rm M}_\odot {\rm pc}^{-2}$ (see Fig. 4). This avoids taking an arbitrary cut on $\Sigma _{\rm H_2}$, but also captures much lower levels of CO emission than could be detected by current observatories.
|
[
"Chomiuk & Povich 2011"
] |
[
"These values are consistent with the current observed SFR in the Milky Way"
] |
[
"Similarities"
] |
[
[
708,
729
]
] |
[
[
583,
657
]
] |
2022ApJ...928..167Z__Krüger_&_Foucart_2020_Instance_1
|
Compared with BNS mergers, which would definitely eject a certain amount of materials to produce EM signals, some NSBH binaries may not tidally disrupt the NS component and, hence, would not make bright EM counterparts such as sGRBs and kilonovae.
9
9
During the final merger phase for plunging NSBH binaries, some weak EM signals may be produced because of the charge and magnetic field carried by the NS (e.g., Dai 2019; Pan & Yang 2019; Zhang 2019; D’Orazio et al. 2022; Sridhar et al. 2021). The tidal disruption probability of NSBH mergers and the brightness of NSBH EM signals are determined by the BH mass, BH spin, NS mass, and NS equation of state (EoS; e.g., Belczynski et al. 2008; Kyutoku et al. 2011, 2013, 2015; Fernández et al. 2015; Kawaguchi et al. 2015, 2016; Foucart 2012; Foucart et al. 2018; Barbieri et al. 2019; Krüger & Foucart 2020; Fragione & Loeb 2021; Fragione 2021; Zhu et al. 2020, 2021c, 2021e; Raaijmakers et al. 2021; Li & Shen 2021; Tiwari et al. 2021). An NSBH merger tends to be a disrupted event and produces bright EM signals if it has a low-mass BH with a high projected aligned spin and a low-mass NS with a stiff EoS. The parameter space in which an NSBH merger can undergo tidal disruption may be very limited. Recently, LIGO–Virgo–KAGRA (LVK) Collaboration reported three high-confidence GWs from NSBH candidates, i.e., GW190814, GW200105_162426, and GW200115_042309 (Abbott et al. 2020, 2021a; Nitz et al. 2021). In spite of many efforts for follow-up observations of these three events, no confirmed EM counterpart candidate has been identified (e.g., Coughlin et al. 2020; Gompertz et al. 2020; Kasliwal et al. 2020;Page et al. 2020; Thakur et al. 2020; Alexander et al. 2021; Anand et al. 2021; Dobie et al. 2022; Kilpatrick et al. 2021). Abbott et al. (2021a), Zhu et al. (2021c), and Fragione (2021) showed that the parameter space of these GW candidates mostly lies outside the disrupted parameter region, so these candidates are likely plunging events with a high probability. There are many mysteries surrounding NSBH binaries, such as the proportion of disrupted events in cosmological NSBH mergers, their cosmological contribution to elements heavier than iron, the formation channel of NSBH binaries, and so on. Systemic research on the population properties of NSBH binaries can help us address these mysteries and unveil the nature of cosmological NSBH binaries.
|
[
"Krüger & Foucart 2020"
] |
[
"The tidal disruption probability of NSBH mergers and the brightness of NSBH EM signals are determined by the BH mass, BH spin, NS mass, and NS equation of state (EoS; e.g."
] |
[
"Uses"
] |
[
[
837,
858
]
] |
[
[
498,
669
]
] |
2019ApJ...877...33Z__Dunn_et_al._2010_Instance_1
|
To investigate the emission-line properties in the blue system, we employed the photoionization code Cloudy (Version 17.01, Ferland et al. 2017) and applied the measured emission-line ratios to these models to simulate the possible physical conditions and processes in the medium. The simple model is a slab-shaped gas with a unique density and homogeneous chemical composition of solar values, irradiated directly by the central ionization continuum source. For the BLR around one of the hypothesized binary black holes, such primordial models can be simply described using parameters such as density nH, hydrogen column density NH, ionization parameter U, and spectral energy distribution (SED) of the incident radiation. A typical AGN ionization continuum is applied as the incident SED, which is a combination of a blackbody “Big Bump” and power laws.6
6
See details in Hazy, a brief introduction to Cloudy; http://www.nublado.org.
The component peaks at ≈1 Ryd and is parameterized by TBB = 1.5 × 105 K. The slope of the X-ray component, the X-ray to UV ratio, and the low-energy slope are set as αx = −2 (beyond 100 keV) and −1 (between 1.36 eV and 100 keV), αox = −1.4, and αUV = −0.5, respectively. This UV-soft SED is regarded as more realistic for radio-quiet quasars than the other available SEDs provided by Cloudy (see the detailed discussion in Section 4.2 of Dunn et al. 2010). We calculated a series of photoionization models with different ionization parameters, densities, and hydrogen column densities. The ranges of these parameters are −4.0 ≤ log10 U ≤ 1.0, 3.0 ≤ log10 nH (cm−3) ≤ 11.0, and 19.0 ≤ log10 NH (cm−2) ≤ 24.0, with a step of 0.2 dex, which could well cover the possible parameter space of the broad-line and narrow-line regions. In Figure 3, the flux ratios of He iλ10830/Hα are presented in the log U–log NH space for the five densities (log nH (cm−3) = 3, 5, 7, 9, and 11). The extensive parameter space is almost enough to cover all the routine possibilities of the AGN “normal” emission-line regions. However, the typical BLR gases with nH ∼ 109–1010 cm−3, NH ∼ 1022 cm−2 and U ∼ 10−2–10−1 present the relatively strong He i λ10830 emission. If the blue system of SDSS J1536+0441 is also compared, the observed ratio of He i10830/Hα6564 is at least larger than 0.1, which is two times the 3σ upper limit. Further extensive calculation in a larger parameter space suggests that a parameter combination range of UnH ≈ 1011.5 cm−3 with an exceedingly high density of nH ≥ 1012 cm−3 would reproduce the observed flux ratios of the hydrogen and helium lines. If the blue system emits from such a high-density medium, the size would be 50 times less than that of the red system BLR (based on the sub-parsec binary hypothesis, in which both BLRs are illuminated by the same ionizing flux), and the number ratios of ions in the blue system and red system would be only approximately five per thousand. Even accounting for the higher emission efficiency (within an order of magnitude), the high-density medium is not sufficient to emit blue system lines comparable to those of the red system.
|
[
"Dunn et al. 2010"
] |
[
"This UV-soft SED is regarded as more realistic for radio-quiet quasars than the other available SEDs provided by Cloudy (see the detailed discussion in Section 4.2 of"
] |
[
"Uses"
] |
[
[
1376,
1392
]
] |
[
[
1209,
1375
]
] |
2021MNRAS.500.3438O__Vidotto_et_al._2014b_Instance_1
|
In addition to the wind models of λ And, we also present here the first full surface magnetic field observations of this star, finding a strong magnetic field for such an evolved star. These observations, carried out with the NARVAL spectropolarimeter, allow us to constrain the surface magnetic field of λ And. These derived surface magnetic fields can constrain the lower boundary of the 3D magnetohydrodynamic wind simulations that we run. Usually, we see a decay in magnetic field strength as solar-type stars evolve, as their activity decreases along with their rotation (Skumanich 1972; Vidotto et al. 2014b; Booth et al. 2020). However, this subgiant star seems to have a relatively strong large-scale magnetic field compared to the Sun. The exact process through which this star would reach this stage in its evolution with such a magnetic field is yet unknown. Potential reasons are that it began with a much stronger dynamo in its past than anticipated, or perhaps the secondary companion had some effect on the primary star at a point in the past. λ And differs from the Sun as it is an RS Canum Venaticorum (RS CVn) variable, meaning it is a variable binary system. The variability on this star is likely due to magnetic spots coming in and out of view due to stellar rotation (Baliunas & Dupree 1979, 1982; Donati, Henry & Hall 1995; Henry et al. 1995; O’Neal et al. 2001; Sanz-Forcada, Brickhouse & Dupree 2001; Frasca et al. 2008; Drake et al. 2011). RS CVn systems, in particular, can present observed levels of chromospheric and coronal activity that are orders of magnitude higher than in single stars with similar spectral types (Ayres & Linsky 1980; Walter & Bowyer 1981). This is likely caused by the increase in activity when the two stars interact with each other, which can lead to rotational synchronization of the system (e.g. Lanza & Rodonò 2004; an analogous process has been inferred to take place in close-in planet–star systems, Cuntz, Saar & Musielak 2000). For the purposes of this work, we assume the binarity of this system does not affect our wind models. Compared to the Sun, λ And is metal-poor ([Fe/H] = −0.46 ± 0.04 dex, Maldonado & Villaver 2016). We do not include the effects of different metal abundances on the stellar wind and stellar evolution, but the effects of which have been examined in other works (Suzuki 2018).
|
[
"Vidotto et al. 2014b"
] |
[
"Usually, we see a decay in magnetic field strength as solar-type stars evolve, as their activity decreases along with their rotation",
"However, this subgiant star seems to have a relatively strong large-scale magnetic field compared to the Sun. The exact process through which this star would reach this stage in its evolution with such a magnetic field is yet unknown."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
593,
613
]
] |
[
[
443,
575
],
[
635,
869
]
] |
2022MNRAS.509.3599T__Matteo,_Springel_&_Hernquist_2005_Instance_1
|
The typical value of the mass outflow rate for sources accreting below or close to the Eddington limit is $\dot{M}_{\rm out} \gtrsim 5\!-\!10{{\ \rm per\ cent}} \,\dot{M}_{\rm acc}$, for both UFOs and non-UFOs (Tombesi et al. 2012). In this scenario, even for the UFOs with the lowest allowed velocity, the mechanical power is enough to exercise a significant feedback impact on the surrounding environment. Looking at the comparison between mass accretion rate and mass outflow rate for our source (see Table 3), the upper limit on the mass outflow rate for Wind 1 is extremely high, but the values for Wind 2 and Wind 3 are still comparable with the values of quasars and Seyfert galaxies (Tombesi et al. 2012). If we consider the lower limits instead, their values are well below the average. Theoretical works (Di Matteo, Springel & Hernquist 2005; King 2010; Ostriker et al. 2010; Debuhr, Quataert & Ma 2011) showed that, in order to have a significant feedback impact in the environment surrounding an AGN, it is required a minimum ratio between the mechanical power of the outflow and the bolometric luminosity of ${\sim}0.5{{\ \rm per\ cent}}$. Tombesi et al. (2012) showed that actually the lower limit of this value for UFOs is ${\sim}0.3{{\ \rm per\ cent}}$ and for non-UFOs is ${\sim}0.02 \!-\! 0.8{{\ \rm per\ cent}}$. According to what found for the ratio between the mass outflow rate and the mass accretion rate, looking at the upper limits on $\dot{K}/L_{\rm b,out}$ of our source, i.e. multiphase and multiscale X-ray winds, IRAS 04416+1215 fits well in this scenario in which the outflowing winds can impress a feedback. Indeed, the upper limits on $\dot{K}/L_{\rm b,out}$ is comparable with the kinetic coupling efficiency, defined as the ratio of the kinetic luminosity of outflows to the AGN radiative luminosity (Eout/Lrad), calculated with the feedback model for hyper-Eddington accretion by Takeo, Inayoshi & Mineshige (2020), using the outflow velocities and the $\dot{M}/\dot{M}_{\rm acc}$ values of IRAS 04416+1215. Instead, the lower limits are below the minimum value required to generate at least a weak feedback. Considering only the values derived from the lower limits on the distance, i.e. the situation in which the X-ray winds are cospatial, we would be in a scenario in which the source loses much luminosity due to advection inside the disc, resulting in a much lower efficiency for wind production as most of the radiation remains trapped inside the disc. This deduction is supported also by the results of the lower limits on the ratio between the momentum rate of the outflows and the momentum of the radiation. Outflows accelerated through the continuum radiation pressure are expected to have a $\dot{p}_{\rm out}/\dot{p}_{\rm rad}\sim 1$ (King & Pounds 2015). The median value of this ratio for UFOs is ∼0.96 after the relativistic correction and ∼0.64 without the relativistic corrections (Luminari et al. 2020). The values we found for the lower limits on $\dot{p}_{\rm out}/\dot{p}_{\rm rad}$ of IRAS 04416+1215 are again well below the median. Thus, the outcoming luminosity of the source is not enough to accelerate the material to the escape velocity, which is required for a wind to leave the system, suggesting that likely in the scenario of the cospatial winds the outflows observed in IRAS 04416+1215 could be accelerated by other mechanisms such as magnetohydrodynamic processes.
|
[
"Di Matteo, Springel & Hernquist 2005"
] |
[
"Theoretical works",
"showed that, in order to have a significant feedback impact in the environment surrounding an AGN, it is required a minimum ratio between the mechanical power of the outflow and the bolometric luminosity of ${\\sim}0.5{{\\ \\rm per\\ cent}}$.",
"Instead, the lower limits are below the minimum value required to generate at least a weak feedback. Considering only the values derived from the lower limits on the distance, i.e. the situation in which the X-ray winds are cospatial, we would be in a scenario in which the source loses much luminosity due to advection inside the disc, resulting in a much lower efficiency for wind production as most of the radiation remains trapped inside the disc."
] |
[
"Uses",
"Uses",
"Uses"
] |
[
[
815,
851
]
] |
[
[
796,
813
],
[
914,
1152
],
[
2044,
2495
]
] |
2020ApJ...892...68N__Laurikainen_et_al._2004_Instance_1
|
We also examine our
measurement of NGC 3504 in the context of the empirical compilations of
and
scaling relations (Kormendy & Ho 2013; McConnell et al. 2013; Scott et al. 2013; Saglia et al. 2016) in Figure 16. The stellar-bulge velocity dispersion of NGC 3504 is determined from Ho et al. (2009), while the bulge mass of
M⊙ was derived from Section 2.2. For a sanity check, we estimate the bulge mass of NGC 3504 using our H-band MGE model and adapting the effective radius of the bulge derived from the bulge-disk-bar decomposition model (Laurikainen et al. 2004). Accounting for the dynamical M/LH (Section 5.2), we obtain
M⊙. Our result shows that the best-fit
of NGC 3504 is fully consistent with the empirical
and
relations of Kormendy & Ho (2013), McConnell et al. (2013), and Saglia et al. (2016), but outside
uncertainty of the Scott et al. (2013) and Savorgnan et al. (2016) empirical
relations for “Sérsic” galaxies (those without central cores). Compared with the theoretical predictions of the bimodality in the BH accretion efficiency model (e.g., Pacucci et al. 2015, 2018), our measurement is consistent with the
correlation, but a positive outlier is the
relation up to 1σ. At the mass of
M⊙, the SMBH of NGC 3504 lies within the same mass regime as the BHs derived in the Combes et al. (2019) sample and lies between the samples of lower- and higher-mass galaxies previously and currently studied with ALMA (Davis et al. 2013, 2017, 2018; Onishi et al. 2015, 2017; Barth et al. 2016a, 2016b; Boizelle et al. 2019; Nagai et al. 2019; North et al. 2019; Smith et al. 2019, T. Davis et al. 2020, in preparation, D. Nguyen et al. 2020, in preparation), respectively. All of these works prove that the cold gas-dynamical method observed with ALMA at high spatial resolution now can work well in a wide range of BH masses covering six orders of magnitude from 105 to
.
|
[
"Laurikainen et al. 2004"
] |
[
"For a sanity check, we estimate the bulge mass of NGC 3504 using our H-band MGE model and adapting the effective radius of the bulge derived from the bulge-disk-bar decomposition model"
] |
[
"Uses"
] |
[
[
568,
591
]
] |
[
[
382,
566
]
] |
2018ApJ...861...28S__Temmer_et_al._2011_Instance_1
|
In addition, the graduated cylindrical shell (GCS) model, which was developed by Thernisien et al. (2006, 2009) and Thernisien (2011), is applied to obtain the three-dimensional parameters of these CMEs. Figure 5 shows the fitting results of these CMEs. Seen from these images, the GCS model can well represent the topology of these CMEs. The last three columns in Table 1 show the fitting results of these CMEs, including the propagation directions, velocities and face-on angular widths. Assuming a constant velocity and considering the influence of the propagation direction and angular width on the prediction of the arrival time suggested by Shen et al. (2014), CME-1 would arrive at the Earth around the time of September 6 22:27 UT. However, previous results show that fast CME would decelerate during their propagation in interplanetary space (e.g., Gopalswamy et al. 2001a, 2005; Vršnak 2001; Vršnak & Žic 2007; Temmer et al. 2011; Lugaz & Kintner 2012, and reference therein). Such deceleration may make the CME-1 arrive at the Earth later than September 6 22:27 UT. Thus, CME-1 is more likely to be the solar source of ICME-1. Seen from the Figure 3, the front edge of the CME-2 is lower than the front edge of CME-1 indicating that CME-2 would arrive at the Earth later than CME-1. Thus, CME-2 was the solar source of the ICME-2. It should be noted that, based on the fitting results of GCS model, CME-2 is faster than CME-1 and their propagation directions are close to each other. Thus, these two CMEs are expected to interact in interplanetary space. Seen from the in situ observations, possible interaction region signatures were detected between these two ICMEs with lower magnetic field, higher velocity, high density, and higher plasma beta. Furthermore, similar analysis shows that CME-3 might arrive at Earth after September 7 19:21 UT. Considering the long duration of Ejecta-4 and the larger angular width of CME-3, we verify that CME-3 is the solar source of ICME-4 and the driver of the second shock. It should be noted that, no obvious Earth-directed CME could be identified as the solar source of Ejecta-3. A possible explanation is that this ejecta structure is formed in the sheath region of ICME-4 during its propagation outward (e.g., Zheng & Hu 2018, and reference therein).
|
[
"Temmer et al. 2011"
] |
[
"However, previous results show that fast CME would decelerate during their propagation in interplanetary space (e.g.,"
] |
[
"Compare/Contrast"
] |
[
[
921,
939
]
] |
[
[
740,
857
]
] |
2018AandA...613A...3Q__Kelly_et_al._2017_Instance_1
|
As a prototypical Seyfert 2 galaxy with starburst at a distance of 14.4 Mpc (1″ = 72 pc, Bland-Hawthorn et al. 1997), NGC 1068 was observed at radio (Greenhill et al. 1996), millimeter (Schinnerer et al. 2000), infrared (Jaffe et al. 2004), optical (Antonucci & Miller 1985), UV (Antonucci et al. 1994), and X-ray (Kinkhabwala et al. 2002). High spatial resolution CO (1–0) observations show two molecular spiral arms with a diameter of ~40″ and a northern half-bar, while a CO (2–1) map reveals a nuclear ring with two bright knots in the CND region (Schinnerer et al. 2000). The dense gas fraction as traced by HCN (1–0) (Tacconi et al. 1994; Helfer & Blitz 1995) and CS (2–1) (Tacconi et al. 1997; Takano et al. 2014) in the nuclear region is higher than the two arms. Observations of CO (3–2) (Krips et al. 2011; Tsai et al. 2012; García-Burillo et al. 2014) showed that the difference of molecular gas temperatures between the nuclear region and the two arms was not as large as that of densities. Dozens of molecular lines at millimeter wavelength were detected at CND with single-dish observations (Usero et al. 2004; Nakajima et al. 2011, 2013; Aladro et al. 2013). Moreover, several molecules were detected and resolved toward NGC 1068 with interferometers in the past few years (Tosaki et al. 2017; Kelly et al. 2017; Furuya & Taniguchi 2016; Izumi et al. 2016; Imanishi et al. 2016; Nakajima et al. 2015; Viti et al. 2014; Takano et al. 2014; García-Burillo et al. 2014, 2016). The molecular gas in the CND region was denser and hotter than that in the starburst ring, while chemical properties in the two regions were also different (Viti et al. 2014). The highest molecular gas temperature was higher than 150 K, and the gas density was above 105 cm−3 in the CND region (Viti et al. 2014). The distribution of different species of molecules were also different: CO isotopic species, for instance, were enhanced in the starburst ring, while the shock/dust related molecules were enhanced in the CND region (Nakajima et al. 2015). The spatially resolved observations showed that the CND region was a complex dynamical system. For instance, the east and west dots were dominated by a fast shock and a slower shock (Kelly et al. 2017), while the dust torus also showed complex kinematics (García-Burillo et al. 2016). Gas inflow was driven by a past minor merger (Furuya & Taniguchi 2016), while the outflow was AGN driven (García-Burillo et al. 2014). We conducted adeeper survey of millimeter lines toward the CND region of NGC 1068 with the IRAM 30 m telescope, with the goal to quantify the gas properties in the CND. Compared to previous single-dish observations, our data probe weaker transition lines, which could place more constraints on the physical and chemistry properties of the CND.
|
[
"Kelly et al. 2017"
] |
[
"Moreover, several molecules were detected and resolved toward NGC 1068 with interferometers in the past few years"
] |
[
"Background"
] |
[
[
1309,
1326
]
] |
[
[
1174,
1287
]
] |
2016MNRAS.461..666K__Joshi_et_al._2011_Instance_2
|
C-statistic (e.g. Jang & Miller 1997) is the most commonly used and the one-way analysis of variance (ANOVA; de Diego 2010) the most powerful test for verifying the presence of variability in a DLC. However, we did not employ either of these tests because, de Diego (2010) has questioned the validity of the C-test by arguing that the C-statistics does not have a Gaussian distribution and the commonly used critical value of 2.567 is too conservative. On the other hand, the ANOVA test requires a rather large number of data points in the DLC, so as to have several points within each sub-group used for the analysis. This is not feasible for our DLCs which typically have no more than about 30–45 data points. Therefore, we have instead used the F-test which is based on the ratio of variances, F = variance(observed)/variance(expected) (de Diego 2010; Villforth, Koekemoer & Grogin 2010), with its two versions : (i) the standard F-test (hereafter Fη-test, Goyal et al. 2012) and (ii) scaled F-test (hereafter Fκ-test, Joshi et al. 2011). The Fκ-test is preferred when the magnitude difference between the object and comparison stars is large (Joshi et al. 2011). Onward Paper II, we have only been using the Fη-test because our objects are generally quite comparable in brightness to their available comparison stars. An additional gain from the use of the Fη-test is that we can directly compare our INOV results with those deduced for other major AGN classes (Goyal et al. 2013). An important point to keep in mind while applying the statistical tests is that the photometric errors on individual data points in a given DLC, as returned by the algorithms in the iraf and daophot softwares are normally underestimated by the factor η which ranges between 1.3 and 1.75, as estimated in independent studies (e.g. Gopal-Krishna, Sagar & Wiita 1995; Garcia et al. 1999; Sagar et al. 2004; Stalin et al. 2004a; Bachev, Strigachev & Semkov 2005). Recently, using a large sample, Goyal et al. (2013) estimated the best-fitting value of η to be 1.5, which is adopted here. Thus, the Fη statistics can be expressed as
\begin{equation*}
F_{1}^{\eta } = \frac{\sigma ^{2}_{({\rm q-s1})}}{ \eta ^2 \langle \sigma _{{\rm q-s1}}^2 \rangle }, \hspace{5.69046pt} F_{2}^{\eta } = \frac{\sigma ^{2}_{({\rm q-s2})}}{ \eta ^2 \langle \sigma _{{\rm q-s2}}^2 \rangle }, \hspace{5.69046pt} F_{{\rm s1-s2}}^{\eta } = \frac{\sigma ^{2}_{({\rm s1-s2})}}{ \eta ^2 \langle \sigma _{{\rm s1-s2}}^2 \rangle },\end{equation*}
where $\sigma ^{2}_{({\rm q-s1})}$, $\sigma ^{2}_{({\rm q-s2})}$ and $\sigma ^{2}_{({\rm s1-s2})}$ are the variances of the ‘quasar–star1’, ‘quasar–star2’ and ‘star1–star2’ DLCs and $\langle \sigma _{{\rm q-s1}}^2 \rangle =\sum _{\boldsymbol {i}=0}^{N}\sigma ^2_{i,{\rm err}}({\rm q-s1})/N$, $\langle \sigma _{{\rm q-s2}}^2 \rangle$ and $\langle \sigma _{{\rm s1-s2}}^2 \rangle$ are the mean square (formal) rms errors of the individual data points in the ‘quasar–star1’, ‘quasar–star2’ and ‘star1–star2’ DLCs, respectively. η is the scaling factor (= 1.5) as mentioned above.
|
[
"Joshi et al. 2011"
] |
[
"The Fκ-test is preferred when the magnitude difference between the object and comparison stars is large"
] |
[
"Compare/Contrast"
] |
[
[
1147,
1164
]
] |
[
[
1042,
1145
]
] |
2019AandA...630A...2H__Kofman_et_al._(2015)_Instance_1
|
Several laboratory experiments have been performed to determine the mechanical properties of possible building blocks of comets. In particular, Güttler et al. (2009), Schräpler et al. (2015), Lorek et al. (2016), and Katsuragi & Blum (2017) investigated how the compressibility of ice layers and different dust structures created by random ballistic deposition as well as pebble sub-structures (with scales of centimeters, millimeters, and 0.1 mm) depend on the volume filling factor. The respective results are summarized in Fig. 6. Remote observations lead to multiple independent estimates for the volume filling factor. Kofman et al. (2015) derived values of 0.15–0.25, Pätzold et al. (2016) estimated 0.25–0.30, and a study by Fulle et al. (2016) resulted in 0.21–0.37. The corresponding range is depicted in Fig. 6 by the horizontal bar shaded in blue. As thevolume filling factor is known, the laboratory measurements can be used to determine the most likely material of the surface. For the observed volume filling factor, dust and ice layers have a compressive strength above 103 Pa. A surface made up of such dust or ice layers is therefore inconsistent with the results of Groussin et al. (2015) or the compressive strength derived as part of this work. In contrast, a surface consisting of layers of dust and ice aggregates could explain all observations, except for the MPa range derived by Spohn et al. (2015). This is strong evidence for the presence of aggregate layers on the surface of 67P, that is, a surface composed of sub-decimeter-sized aggregates, as inferred from the gravitational collapse scenario. This conclusion can be drawn without precise knowledge of the actual value of the compressive strength, as the upper limit derived as part of this work and by Groussin et al. (2015), even considering all possible errors, is well below 103 Pa for different parts of 67P. The presence of such aggregate layers can only be explained by a formation dominated by relatively low-velocity collisions. Hierarchical agglomeration would cause a significantly higher compressive strength as a result of impact-compaction of the pebbles and possible break-up of aggregates (Blum 2018). The derived mechanical properties combined with the available laboratory models therefore suggest that the gentle collapse of aggregate ensembles played a major role in the formation ofcomets like 67P.
|
[
"Kofman et al. (2015)"
] |
[
"derived values of 0.15–0.25",
"The corresponding range is depicted in Fig. 6 by the horizontal bar shaded in blue. As thevolume filling factor is known, the laboratory measurements can be used to determine the most likely material of the surface."
] |
[
"Uses",
"Uses"
] |
[
[
624,
644
]
] |
[
[
645,
672
],
[
775,
990
]
] |
2020AandA...637A..44N__Kraus_(2018)_Instance_3
|
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)"
] |
[
"are at the levels of >4σ for the analysis of"
] |
[
"Compare/Contrast"
] |
[
[
1345,
1357
]
] |
[
[
1300,
1344
]
] |
2020MNRAS.491..903W__Rosotti_et_al._2014_Instance_1
|
The process of planet formation is strongly dependent on the stellar birth environment. The majority of stars exist in clusters or associations within their first few Myr of evolution (Lada & Lada 2003; Longmore et al. 2014; Krumholz, McKee & Bland-Hawthorn 2019), during which time they also host protoplanetary discs (PPDs – e.g. Haisch, Lada & Lada 2001; Ribas et al. 2014). Multiple feedback mechanisms influence disc evolution. In sufficiently dense environments, star–disc encounters can truncate the disc and induce increased accretion rates (Clarke & Pringle 1993; Ostriker 1994; Hall, Clarke & Pringle 1996; Pfalzner et al. 2005a; Olczak, Pfalzner & Spurzem 2006; Pfalzner, Olczak & Eckart 2006; de Juan Ovelar et al. 2012; Breslau et al. 2014; Rosotti et al. 2014; Winter et al. 2018a). Recent studies indicate that in the solar neighbourhood such interactions only have a significant effect in the early stages of cluster evolution due to enhanced stellar multiplicity and substructure, and therefore set initial conditions rather than destruction time-scales (Bate 2018; Winter et al. 2018b; Winter, Booth & Clarke 2018c). However, in regions with massive stars, external photoevaporation by far-ultraviolet (FUV) and extreme-ultraviolet (EUV) photons can rapidly disperse PPDs (Johnstone, Hollenbach & Bally 1998; Störzer & Hollenbach 1999; Armitage 2000; Clarke 2007; Fatuzzo & Adams 2008; Adams 2010; Facchini, Clarke & Bisbas 2016; Ansdell et al. 2017; Haworth et al. 2018b; Winter et al. 2018b). Additionally, before the dispersal of the parent giant molecular cloud (GMC), ram pressure stripping can truncate PPDs (Wijnen et al. 2017a) or additional material can be accreted (Moeckel & Throop 2009; Scicluna et al. 2014), leading to the destruction and reforming of discs during the embedded phase (Bate 2018). If a PPD is destroyed quickly by feedback in dense stellar environments, planets may be unable to form, depending on the efficiency of the formation mechanisms (Youdin & Goodman 2005; Johansen & Lambrechts 2017; Ormel, Liu & Schoonenberg 2017; Haworth et al. 2018a). Given the apparent ubiquity of grouped star formation, quantifying the destruction time-scales for PPDs due to neighbour feedback is of great relevance for understanding the demographics of PPDs and exoplanetary systems.
|
[
"Rosotti et al. 2014"
] |
[
"Multiple feedback mechanisms influence disc evolution. In sufficiently dense environments, star–disc encounters can truncate the disc and induce increased accretion rates"
] |
[
"Background"
] |
[
[
754,
773
]
] |
[
[
378,
548
]
] |
2021AandA...647A.132K__Kouloumvakos_et_al._2015_Instance_1
|
Studies of SEP events are important for different reasons. On one hand, solar eruptive events are well-observed processes of energetic-particle acceleration (Vainio & Afanasiev 2018), which can be studied in detail using a multi-messenger approach, complementing particle data with observations in different wavelengths (e.g., Plainaki et al. 2014; Cliver 2016; Kocharov et al. 2017). For this purpose, the peak flux intensity and detailed temporal variability of the particle flux are important as signatures of the acceleration process in the solar corona and the interplanetary medium (e.g., Desai & Giacalone 2016; Kong et al. 2017). Accordingly, numerous studies were focused on peak fluxes of SEPs and corresponding acceleration and transport processes (e.g., Kouloumvakos et al. 2015; Kocharov et al. 2017). On the other hand, enhanced fluxes of energetic particles affect the radiation environment near the Earth (e.g., Webber et al. 2007; Mishev et al. 2015), making not only the peak fluxes but also the fluence (event-integrated flux) and its spectral shape of significant importance, especially for extreme events (e.g., Cliver et al. 2020). We emphasise that SEP fluences can not be used for the detailed study of SEP acceleration processes, because (i) the observations at 1 AU are also modified by transport, and (ii) different energies in the fluence spectrum can be dominated by different acceleration mechanisms or by the same mechanism operating under different conditions. It is evident from the proton time-intensity profiles alone that the fluence at MeV and 10 MeV energies is often dominated by acceleration at interplanetary shocks (e.g., Reames 1999). However, the question is more open at 100 MeV and GeV energies peaking much earlier, with possible contributions from flares and/or coronal shocks as the main candidates to account for the acceleration (see Cliver 2016, and references therein). Even if the same CME-driven shock were responsible for the acceleration of 10 MeV and 1 GeV protons, the former would typically be accelerated mainly in the solar wind and the latter in the corona, and there is no reason to suggest that the spectral form of the fluence would reveal something common about the accelerator properties.
|
[
"Kouloumvakos et al. 2015"
] |
[
"Accordingly, numerous studies were focused on peak fluxes of SEPs and corresponding acceleration and transport processes (e.g.,"
] |
[
"Background"
] |
[
[
766,
790
]
] |
[
[
638,
765
]
] |
2015MNRAS.450.4364N__Wu_et_al._2004_Instance_1
|
Low- and intermediate-mass stars are formed by the gravitational collapse of the parental giant molecular cloud (GMC), followed by the accretion process (Palla 1996). During the accretion phase, material is ejected as well via collimated bipolar jets. However, when a YSO reaches 8 M⊙, the radiative flux becomes so intense (using ϕ = L/4πd2, the ratio between the radiative fluxes of an O5 and a B3 star – masses of ∼40 and ∼8 M⊙, respectively – is ≈250) that it may interrupt the accretion flow. A process that constrains the outcoming radiation field to narrower angles may leave some room for the accretion process to continue in some directions. This seems to be the case for the outflows driven by young stars from a very broad mass range, as previous reported by several authors (Bachiller 1996; Bontemps et al. 1996; Shepherd & Churchwell 1996; Beuther et al. 2002; Wu et al. 2004). Outflows associated with high-mass objects are expected to be more energetic than the outflows observed in lower mass YSOs (Beuther et al. 2005; Zhang et al. 2005; López-Sepulcre et al. 2009), with velocities greater than ∼100 km s−1 (Martí, Rodríguez & Reipurth 1998). Some authors have found evidences that outflows associated with massive stars are scaled up versions of their low-mass counterparts (Vaidya et al. 2011; Codella et al. 2013) while other works have reported that no well-collimated outflows have been found towards MYSOs (Shepherd, Testi & Stark 2003; Sollins et al. 2004). Massive YSO outflows mapped in high-velocity CO lines have collimation factors R = length/width ∼2.05 ± 0.96 as compared to R ∼ 2.81 ± 2.16 for low-mass stars (Wu et al. 2004), indicating a weak tendency that outflows associated with massive stars are less collimated than those from low-mass stars as previously thought (Richer et al. 2000). Besides the degree of collimation, these massive outflows would work removing mass from the plane of the accretion disc, lowering the density on the plane and, therefore, facilitating the accretion flow to reach the stellar core as shown in the recent 3D simulations presented by Krumholz et al. (2009). Although these authors have not included the outflow activity on their simulations, they argue that the presence of outflows would decrease the star formation efficiency from 70 per cent (considering purely radiation effects) to 50 per cent.
|
[
"Wu et al. 2004"
] |
[
"A process that constrains the outcoming radiation field to narrower angles may leave some room for the accretion process to continue in some directions. This seems to be the case for the outflows driven by young stars from a very broad mass range, as previous reported by several authors"
] |
[
"Compare/Contrast"
] |
[
[
874,
888
]
] |
[
[
498,
785
]
] |
2016AandA...589A..44G__within_2000_Instance_1
|
W51e2 is the strongest and best-studied HC HII region in the W51 Main cluster, and it is believed to be powered by an O8-type young star (e.g., Shi et al. 2010a). A number of interferometric studies conducted with varying angular resolutions, at centimetre (cm) and (sub)millimetre (mm) bands, identified molecular and ionized gas undergoing infall and rotation toward W51e2. VLA observations of the NH3 inversion lines (1, 1) and (2, 2) seen in absorption (1\hbox{$\farcs$}.̋1 beamsize) revealed radial infall on scales larger than 5000 AU toward the W51e2 core (Zhang & Ho 1997). Higher angular resolution observations of the (3, 3) NH3 absorption line (0\hbox{$\farcs$}.̋3 beamsize) showed signatures of rotation within 2000 AU based on a position-velocity (pv) diagram (Zhang & Ho 1997). Zhang et al. (1998) identified a velocity gradient in a CH3CN transition at 2 mm, deriving a position angle (PA) of 20 ± 20°. Keto & Klaassen (2008) imaged the H53α radio recombination line (RL) with the VLA (0\hbox{$\farcs$}.̋45 beamsize) and they claimed rotation in the ionized gas along the axis of a molecular bipolar outflow (oriented NW-SE) imaged with the SMA in the CO (2−1) line (1′′ beamsize), suggesting a simple inflow/outflow picture in a single high-mass young stellar object (YSO). However, higher resolution observations, using the SMA at the wavelengths of 0.85 mm (0\hbox{$\farcs$}.̋3 beamsize) and 1.3 mm (0\hbox{$\farcs$}.̋7 beamsize), revealed a more complex picture, by resolving W51e2 into three subcores (Shi et al. 2010a): W51e2-W, corresponding to the HC HII region, W51e2-E, located about 1′′ east of the HC HII region and corresponding to the brightest dust continuum source, and W51e2-NW, the weakest continuum component, located about 1′′ NW of the HC HII region. Shi et al. (2010b) imaged the CO (3−2) line (with a 0\hbox{$\farcs$}.̋7 beamsize) and established that the driving source of the powerful molecular outflow in this region is the protostellar core W51e2-E, and not the HC HII region W51e2-W, challenging the scenario proposed by Keto & Klaassen (2008). Etoka et al. (2012) used MERLIN to image the Class II 6.7 GHz CH3OH masers (typical signpost of HMSF), and found that the bulk of maser emission is indeed concentrated toward W51e2-E, and not the HC HII region W51e2-W. This further supports the scenario proposed by Shi et al. (2010a), where the ongoing star formation activity in the region is not concentrated on the HC HII region but toward its companion 1′′ to the east.
|
[
"Zhang & Ho 1997"
] |
[
"VLA observations of the NH3 inversion lines (1, 1) and (2, 2) seen in absorption (1\\hbox{$\\farcs$}.̋1 beamsize) revealed radial infall on scales larger than 5000 AU toward the W51e2 core"
] |
[
"Background"
] |
[
[
564,
579
]
] |
[
[
376,
562
]
] |
2021ApJ...920..139M__in_2002_Instance_1
|
Cepheus X-4 was discovered as a transient source using the X-ray telescope of the OSO-7 satellite during 1972 June–July (Ulmer et al. 1973). Ginga observed the source during the 1988 March outburst and detected a spin period of 66.25 s for its neutron star for the first time (Koyama et al. 1991). Spectroscopic studies from the same Ginga observations led to the detection of a cyclotron resonance scattering feature corresponding to a centroid energy at 30.5 ± 0.4 keV (Mihara et al. 1991). The ROSAT observations during the 1993 June outburst refined the source coordinates and also determined its pulsar spin period (Schulz et al. 1995). The observations by BATSE during 1993 June–July and during the subsequent outburst of 1997 June–July, which was also followed by RXTE, determined the pulse characteristics of Cepheus X-4. Additionally, a possible range of its orbital period from 23 to 147.3 days was suggested using RXTE data (Wilson et al. 1999). From observed characteristic features and outburst activities, it was predicted that Cepheus X-4 could possibly have a massive early-type Be star with its circumstellar disk as a companion, which was thought to be the most likely cause of its long outburst of about 40 days, as seen for other Be binaries. Optical observations of Cepheus X-4 subsequently confirmed it as a Be binary system and estimated its location at a distance of 3.8 ± 0.6 kpc (Bonnet-Bidaud & Mouchet 2005). But this distance estimate was later challenged by Riquelme et al. (2012), who proposed a distance of either 7.9 kpc or 5.9 kpc according to whether the stellar type of the companion is a B1 or B2 star, respectively. The distance of Cepheus X-4 was later reported by Gaia to be
10.2
−
1.6
+
2.2
kpc (Malacaria et al. 2020). Luminosity dependent changes in the pulse profile of Cepheus X-4 were studied during the declining phase of the 1997 outburst (Mukerjee & Agrawal et al. 2000) by combining observations by the Indian X-ray Astronomy Experiment (Agrawal et al. 1996) and RXTE (Rothschild et al. 1998). The RXTE observed another outburst in 2002 and re-established its cyclotron resonance feature corresponding to a centroid energy at 30.7 ± 1.8 keV (as established earlier by Ginga), which did not show a significant dependence on X-ray luminosity, although the continuum became harder with increasing source luminosity (McBride et al. 2007). The source went into outburst again in 2014 and was observed with the Nuclear Spectroscopic Telescope Array (NuSTAR; Harrison et al. 2013) and Suzaku (Mitsuda et al. 2007). The Suzaku observation of the 2014 outburst overlapped with the second observation with NuSTAR on 2014 July 1–2, and detected an additional absorption feature at ≈45 keV in the phase resolved spectra of the pulsar, which was identified as the first harmonic of the fundamental cyclotron line detected at ≈28 keV (Jaiswal & Naik 2015). The source spectra obtained from the NuSTAR observations in 2014 were well fitted by a Fermi–Dirac cutoff (FD-cutoff) model along with an iron emission line, and a cyclotron absorption feature was clearly detected in both of the observations at
30.39
−
0.14
+
0.17
keV and
29.42
−
0.24
+
0.27
keV, respectively. Although, the averaged source luminosity differed by a factor of about 3 between these two observations, it only showed a marginal variation in its centroid energy (Furst et al. 2015). Using the same NuSTAR observations of the 2014 outburst, Vybornov et al. (2017) reported that the spectrum of Cepheus X-4 showed two cyclotron resonance scattering features, with the fundamental line at ≈30 keV and its harmonic at ≈55 keV. They also showed that the energy of the fundamental cyclotron absorption feature increases and the continuum becomes harder with increasing X-ray luminosity. Pulse phase resolved spectroscopic studies of Cepheus X-4 were conducted by Bhargava et al. (2019) at these two different intensities of the source using the same two NuSTAR observations of the 2014 outburst. It was found that the observed cyclotron line profile of Cepheus X-4 had an asymmetric shape in the phase averaged spectrum. However, for phase resolved spectra, a single symmetric cyclotron profile fitted the data well. The spectral continuum and the parameters of the cyclotron line showed some variations with respect to the pulse phase only within a limited pulse phase (Bhargava et al. 2019).
|
[
"McBride et al. 2007"
] |
[
"The RXTE observed another outburst in 2002 and re-established its cyclotron resonance feature corresponding to a centroid energy at 30.7 ± 1.8 keV (as established earlier by Ginga), which did not show a significant dependence on X-ray luminosity, although the continuum became harder with increasing source luminosity"
] |
[
"Compare/Contrast"
] |
[
[
2379,
2398
]
] |
[
[
2060,
2377
]
] |
2022MNRAS.512..439C__Lian_et_al._2021_Instance_2
|
It is still unclear whether this incompatibility is evidence against the spatially flat ΛCDM model or is caused by unidentified systematic errors in one of the established cosmological probes or by evolution of the parameters themselves with the redshift (Dainotti et al. 2021b, 2022). Newer, alternate cosmological probes could help alleviate this issue. Recent examples of such probes include reverberation-mapped quasar (QSO) measurements that reach to redshift z ∼ 1.9 (Czerny et al. 2021; Khadka et al. 2021a,b; Yu et al. 2021; Zajaček et al. 2021), H ii starburst galaxy measurements that reach to z ∼ 2.4 (Mania & Ratra 2012; Chávez et al. 2014; González-Morán et al. 2019, 2021; Cao, Ryan & Ratra 2020, 2022a; Cao et al. 2021a; Johnson, Sangwan & Shankaranarayanan 2022; Mehrabi et al. 2022), QSO angular size measurements that reach to z ∼ 2.7 (Cao et al. 2017, 2020, 2021a; Ryan, Chen & Ratra 2019; Lian et al. 2021; Zheng et al. 2021), QSO flux measurements that reach to z ∼ 7.5 (Risaliti & Lusso 2015, 2019; Khadka & Ratra 2020a,b, 2021, 2022; Lusso et al. 2020; Yang, Banerjee & Ó Colgáin 2020; Li et al. 2021; Lian et al. 2021; Luongo et al. 2021; Rezaei, Solà Peracaula & Malekjani 2021; Zhao & Xia 2021),1 and the main subject of this paper, gamma-ray burst (GRB) measurements that reach to z ∼ 8.2 (Amati et al. 2008, 2019; Cardone, Capozziello & Dainotti 2009; Cardone et al. 2010; Samushia & Ratra 2010; Dainotti et al. 2011, 2013a,b; Postnikov et al. 2014; Wang, Dai & Liang 2015; Wang et al. 2016, 2022; Fana Dirirsa et al. 2019; Khadka & Ratra 2020c; Hu, Wang & Dai 2021; Dai et al. 2021; Demianski et al. 2021; Khadka et al. 2021c; Luongo et al. 2021; Luongo & Muccino 2021; Cao et al. 2021a). Some of these probes might eventually allow for a reliable extension of the Hubble diagram to z ∼ 3–4, well beyond the reach of Type Ia supernovae. GRBs have been detected to z ∼ 9.4 (Cucchiara et al. 2011), and might be detectable to z = 20 (Lamb & Reichart 2000), so in principle GRBs could act as a cosmological probe to higher redshifts than 8.2.
|
[
"Lian et al. 2021"
] |
[
"QSO flux measurements that reach to z ∼ 7.5"
] |
[
"Background"
] |
[
[
1125,
1141
]
] |
[
[
947,
990
]
] |
2021AandA...647A..35B__Laffon_et_al._(2010)_Instance_2
|
When we now compare the photodesorption yields at 541 eV between fluences 5 × 1015 photon cm−2 and at 3 × 1017 photon cm−2 in Fig. 2, the CO2 photodesorption first increases from 2.6 × 10−2 molecule/photon to 7.3 × 10−2 molecule/photon. We also observed this phenomenon for CO photodesorption yield (the data are not shown for more clarity), which increased from 1.9 × 10−2 molecule/photon to 3.5 × 10−2 molecule/photon. Second, the estimated yield for the X-ray photodesorption of CH3OH from pure methanol ice decreased by almost one order of magnitude from 9.0 × 10−3 molecule/photon to 1.3 × 10−3 molecule/photon. This indi- cates that the photodesorption of CH3OH is higher for a lower fluence received by the ice when more intact methanol molecules are present in the ice. This aging process favors the photodesorption of simpler molecules such as CO2 or CO. Laffon et al. (2010) estimated with NEXAFS spectroscopy (at the C K-edge) that X-ray irradiation at 150 eV of pure methanol ice at 20 K leads to a survival rate of 50% for methanol after an absorbed dose of 1.1 MGy. In our fixed-energy experiments, we irradiated pure methanol ice with fluences between 5 × 1015 photon cm−2 and 2 × 1016 photon cm−2. Because we irradiated a volume of 0.1 cm2 × 100 ML, with a meanenergy of ~550 eV, and when we consider a volumic mass of condensed methanol of ~ 0.64 g cm−3 (at 20 K; Luna et al. 2018) and an X-ray absorption cross section of ~ 0.6 Mbarn (Ishii & Hitchcook 1988), the absorbed doses used in our fixed energy experiments change from ~ 2 to ~ 15 MGy, whichis quite similar to the absorbed doses in Laffon et al. (2010). This indicates that we could expect a methanol destruction rate of about 50% for our low-fluence experiments. In similar experiments, when irradiating a H2 O:CH4:NH3 (2:1:1) ice mixture covered by a layer of CO:CH3OH (3:1) with 250–1250 eV X-rays during 120 min with a flux of 7.6 × 1014 photon s−1, higher by almost two order of magnitudes than our experiments, Ciaravella et al. (2020) did not detect a desorption signal on the mass channel 31 (attributed to methanol desorption) and estimated that only ~ 20% of methanol molecules remained intact in the first minutes of the irradiation. The irradiation flux therefore appears to be critical for detecting methanol desorption in X-ray irradiation experiments of methanol-containing ices. A lower X-ray flux appears to favor methanol desorption because the methanol destruction rate is lower. This destruction of methanol molecules could also have a significant effect on the formation and desorption of more complex molecules.
|
[
"Laffon et al. (2010)"
] |
[
"the absorbed doses used in our fixed energy experiments change from ~ 2 to ~ 15 MGy, whichis quite similar to the absorbed doses in"
] |
[
"Similarities"
] |
[
[
1611,
1631
]
] |
[
[
1479,
1610
]
] |
2022MNRAS.515.5267B__Landau_&_Lifshitz_1959_Instance_1
|
Recent studies have shown that one can use energy balance arguments that include the large-scale magnetic field, ${\mathrm{{\boldsymbol {\mathit {B}}}}_0}$, to derive scaling laws between the Alfvénic and kinetic fluid quantities (Federrath 2016; Beattie et al. 2020; Skalidis & Tassis 2021; S+2021). The dimensionless magnetic energy density, by which we mean the magnetic energy density normalized to the mean thermal pressure3$\rho _0 c_{\rm s}^2$, is
(5)$$\begin{eqnarray}
{e_{\mathrm{ mag}}}= \frac{B^2}{8\pi c_{\rm s}^2 \rho _0} = \frac{1}{8\pi c_{\rm s}^2 \rho _0}\Bigg (B_0^2+ \underbrace{2\delta \mathrm{{{\mathit {B}}}}\cdot \mathrm{{{\mathit {B}}}}_0}_{{\substack{\text{coupling} \\\text{term}}}} {} + \delta B^2 \Bigg),
\end{eqnarray}$$where $B_0^2$ is the large-scale field contribution to the total energy, δB2 is the turbulent field contribution, and $2\delta \mathrm{{\boldsymbol {\mathit {B}}}}\cdot \mathrm{{\boldsymbol {\mathit {B}}}}_0$ is the coupling term between the two field components. In the linear perturbation theory limit of the MHD equations, δB2 includes contributions from shear Alfvén, fast and slow magnetosonic compressive eigenmodes (e.g. Landau & Lifshitz 1959). Because $\delta \mathrm{{\boldsymbol {\mathit {B}}}}\cdot \mathrm{{\boldsymbol {\mathit {B}}}}_0 = \delta B_{\parallel }B_0$, the coupling term only contains the component of magnetic field fluctuations that are parallel to the large-scale field. In linear theory, both fast and slow magnetosonic compressible modes are able to perturb the field variables parallel to ${\mathrm{{\boldsymbol {\mathit {B}}}}_0}$, so under the lens of linear theory, the coupling term is the fluctuation contribution from the compressible modes in the turbulence scaled by B0 (Bhattacharjee, Ng & Spangler 1998). Furthermore, for sub-Alfvénic turbulence Beattie et al. (2021b) showed that converging, shocked flows along magnetic field lines excite strong δB∥ fluctuations, which travel roughly at the theoretical fast Alfvén mode speed. Therefore, it is likely, assuming that δB∥/B0 ≪ 1 (this is indeed the case for ${\mathcal {M}_{\text{A0}}}\lt 1$ plasmas; see left panel of fig. 5 in Beattie et al. 2022) where a linear theory may be valid for the magnetic field, the coupling term contains significant energy contributions from fast magnetosonic modes excited by shocked gas that converges and forms dense filaments perpendicular to magnetic field lines.
|
[
"Landau & Lifshitz 1959"
] |
[
"In the linear perturbation theory limit of the MHD equations, δB2 includes contributions from shear Alfvén, fast and slow magnetosonic compressive eigenmodes (e.g."
] |
[
"Uses"
] |
[
[
1178,
1200
]
] |
[
[
1014,
1177
]
] |
2020AandA...641A.155V__Puglisi_et_al._2019_Instance_2
|
The scenario presented above has been formulated in various flavors to individually explain several of the properties reported here. The main addition of this work, namely the excitation of CO in distant main-sequence and starburst galaxies, fits in the general picture that we sketched. The ensemble of properties and correlations that we reported here can be also used to revisit the definition of what a starburst is. A standard operational classification is based on the distance from the observed empirical M⋆-SFR correlation, the main sequence. This proved to be a useful distinction and an excellent predictor of several trends (e.g., Sargent et al. 2014), but recent results, including our present and previous analysis (Puglisi et al. 2019), show that the demarcation between starburst and main-sequence galaxies is more blurred that we previously considered. We do detect starburst-like behaviors in galaxies on the main sequence (Elbaz et al. 2018), likely linked to the existence of transitional objects (Popping et al. 2017; Barro et al. 2017b; Gómez-Guijarro et al. 2019; Puglisi et al. 2019, and in prep. to limit the references to recent works based on submillimeter observations). Such transition might well imply an imminent increase of the SFR, driving the object in the realm of starbursts (e.g., Barro et al. 2017b), or its cessation, bringing the system back onto or even below the main sequence (Gómez-Guijarro et al. 2019; Puglisi et al. 2019), with the CO properties potentially able to distinguish between these two scenarios. Regardless of these transitional objects, a definition of starburst based on ΣSFR, rather than ΔMS, would naturally better account for the observed molecular gas excitation properties, dust temperatures and opacities, or SFE (see also Elbaz et al. 2011; Rujopakarn et al. 2011; Jiménez-Andrade et al. 2018; Tacconi et al. 2020). As an example, in Fig. 8 we show the mean SLED of the subsample of galaxies with both CO (2 − 1) and CO (5 − 4) coverage, split at its median ΣSFR. While only tentative at this stage, this suggests a trend of increasing CO excitation with ΣSFR, consistently with Fig. 7 and what mentioned above.
|
[
"Puglisi et al. 2019"
] |
[
"We do detect starburst-like behaviors in galaxies on the main sequence",
"likely linked to the existence of transitional objects",
"to limit the references to recent works based on submillimeter observations"
] |
[
"Compare/Contrast",
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
1086,
1105
]
] |
[
[
869,
939
],
[
961,
1015
],
[
1120,
1195
]
] |
2018AandA...615A..61C__Gladders_et_al._2013_Instance_1
|
To tackle this problem, one must study the recent star formation history (SFH) of galaxies. This information is embeddedin their spectral energy distribution (SED). However, recovering it through SED modeling is complex and subject to many uncertainties and degeneracies. Although an average SFH of galaxies can be derived assuming that they follow the MS (Heinis et al. 2014; Ciesla et al. 2017), galaxies are expected to exhibit complex SFHs, with short-term fluctuations, requiring sophisticated SFH parametrizations to model them (e.g., Lee et al. 2010; Pacifici et al. 2013; Behroozi et al. 2013; Pacifici et al. 2016). However, implementation of these models is complex and a large library is needed to model all galaxies properties. Instead, numerous studies have used simple analytical forms to model galaxies SFH (e.g., Papovich et al. 2001; Maraston et al. 2010; Pforr et al. 2012; Gladders et al. 2013; Simha et al. 2014; Buat et al. 2014; Boquien et al. 2014; Ciesla et al. 2015; Abramson et al. 2016; Ciesla et al. 2016, 2017). Furthermore, SFH parameters are known to be difficult to constrain from broadband SED modeling with a general agreement on the difficulty to constrain the age of the galaxy, here defined as the age of the oldest star, from broad-band SED fitting (e.g., Maraston et al. 2010; Pforr et al. 2012; Buat et al. 2014; Ciesla et al. 2015, 2017). To understand the origin of the scatter of the MS, we need to use an analytical SFH that is able to recover recent variations of the SFR with a precision better than the scatter of the MS itself, that is, 0.3 dex. Recently, Ciesla et al. (2017) showed that a delayed SFH to which we add a flexibility on the recent SFH provides SFRs that are more accurate than those estimated by other typical analytical SFHs (τ models, delayed, etc.). This SFH was tested in Ciesla et al. (2016). Studying a sample of local galaxies from the Virgo cluster, that is, galaxies known to have undergone a fast drop of star formation activity due to ram pressure stripping, we showed that the amplitude of the flexibility can be constrained by broadband SED modeling as long as ultraviolet (UV) rest frame and near-IR data are available.
|
[
"Gladders et al. 2013"
] |
[
"However, implementation of these models is complex and a large library is needed to model all galaxies properties. Instead, numerous studies have used simple analytical forms to model galaxies SFH (e.g.,"
] |
[
"Background"
] |
[
[
892,
912
]
] |
[
[
625,
828
]
] |
2022ApJ...938...92B__Zikanov_&_Thess_1998_Instance_1
|
Flows with
Rem≪1
and
N∼(1)
have the distinct property that the induced magnetic field is quickly diffused away, yet the Lorentz force is not negligible. This limit is referred to as the quasi-static approximation to MHD (which we call “QMHD” henceforth; Moffatt 1967; Sommeria & Moreau 1982; Davidson 1995; Knaepen & Moreau 2008; Davidson 2013), and has been studied mainly in metallurgy and in MHD experiments due to the typically low conductivity of liquid metals (Alemany et al. 1979; Sommeria 1988; Gallet et al. 2009; Klein & Pothérat 2010; Pothérat & Klein 2014; Baker et al. 2018), although recent numerical studies on its turbulent properties and anisotropy have been done as well (Zikanov & Thess 1998; Burattini et al. 2008; Favier et al. 2010, 2011; Reddy & Verma 2014; Verma 2017). After nondimensionalizing the equations of MHD using the uniform density ρ, ℓ, and u, and taking the limits above, one is left with a single dynamical equation for the velocity
2
∂v∂t+v·∇v=−∇p*−Ro−1xˆ∥Ω×v−N∇−2(xˆ∥B0·∇)2v+F,
where p
* is the total pressure modified by rotation and magnetic pressure, Ro
−1 ≡ 2Ωℓ/u is the inverse Rossby number (quantifying the relative strength of the Coriolis force),
xˆ∥Ω
and
xˆ∥B0
are unit vectors in the direction of rotation and the background magnetic field, respectively, and
F
is a generic forcing term that can include dissipation such as viscosity and a body force (to be specified in Section 3). The background magnetic field is fixed in time and is uniform in space, such that
∇×B0=B0∇×xˆ∥B0=0
. Care must be taken if considering a spatially dependent background magnetic field
B
0
(
x
), as the resulting equation will not be the same. See the discussion in Section 5. This equation is accompanied with the incompressibility condition ∇ ·
v
= 0. The induced magnetic field can be found using a diagnostic relation
3
b=−∇−2xˆ∥B0·∇v,
which would be
b=−∇−2B0·∇v/η
in dimensional variables.
|
[
"Zikanov & Thess 1998"
] |
[
"This limit is referred to as the quasi-static approximation to MHD",
"although recent numerical studies on its turbulent properties and anisotropy have been done as well"
] |
[
"Background",
"Background"
] |
[
[
703,
723
]
] |
[
[
166,
232
],
[
602,
701
]
] |
2016MNRAS.461.2328M__Smith_et_al._2005_Instance_1
|
This work, which focuses on mass reconstruction from gravitational lensing, is only the first in a series, which will exploit our mesh-free numerical techniques. Two regimes are typically distinguished in lensing mass reconstruction. Strong lensing is usually confined to the inner-most core of the gravitational lens and produces spectacular observational constraints such as multiple images of the same source, gravitational arcs or even rings. The domain of weak lensing is further away from the centre of the lens but spans large areas and manifests itself by the weak distortion in the shape of background galaxies behind the lens. Reconstruction techniques are divided into two classes, although this distinction is by no means unique or even consistent in some cases. Parametric techniques (e.g. Kneib et al. 1996; Broadhurst et al. 2005; Smith et al. 2005; Halkola, Seitz & Pannella 2006; Jullo et al. 2007; Zitrin et al. 2009; Oguri 2010; Newman et al. 2013; Jullo et al. 2014; Monna et al. 2014; Johnson et al. 2014, for some recent examples) assume a parametric form of the underlying mass density distribution for the lens and typically make the assumption that light traces mass in the positioning of these parametric forms. On the other hand, free-form1 methods (see e.g. Broadhurst, Taylor & Peacock 1995; Bartelmann et al. 1996; Abdelsalam, Saha & Williams 1998; Bridle et al. 1998; Seitz, Schneider & Bartelmann 1998; Bradač et al. 2005a; Cacciato et al. 2006; Liesenborgs, De Rijcke & Dejonghe 2006; Diego et al. 2007; Jee et al. 2007; Coe et al. 2008; Bradač et al. 2009; Merten et al. 2009; Williams & Saha 2011; Merten et al. 2011, 2015; Diego et al. 2015, for some recent examples) usually do not make this assumption and purely rely on the input data either based on weak lensing, strong lensing or a combination of the two. This is possible while using a reconstruction mesh and directly inverting the underlying equations describing lensing on this mesh. In the following, we introduce a free-form method combining weak and strong lensing, which uses our new mesh-free numerical framework. This method translates original ideas by Bartelmann et al. (1996), Seitz et al. (1998), Bradač et al. (2005a), Cacciato et al. (2006) and Merten et al. (2009) into the flexible and efficient mesh-free numerical domain. Alternative implementations of such ideas can e.g. be found in Bradač et al. (2009).
|
[
"Smith et al. 2005"
] |
[
"Reconstruction techniques are divided into two classes, although this distinction is by no means unique or even consistent in some cases. Parametric techniques (e.g.",
"assume a parametric form of the underlying mass density distribution for the lens and typically make the assumption that light traces mass in the positioning of these parametric forms."
] |
[
"Background",
"Background"
] |
[
[
846,
863
]
] |
[
[
637,
802
],
[
1053,
1237
]
] |
2020ApJ...903L..12H__Zhong_et_al._2020_Instance_1
|
Magnetic reconnection (MR) may occur in various space and astrophysical plasma environments, among which the planetary magnetopause boundaries separating the solar wind and magnetospheric origins of plasmas and magnetic field are some of the most likely sites for the occurrence of MR. Due to the easy access to the in situ spacecraft observations the Earth’s magnetopause is the most widely studied space plasma environment for MR (Paschmann et al. 1979; Vaivads et al. 2004; Graham et al. 2014). In particular, the Magnetospheric Multiscale (MMS) mission has contributed greatly to the kinetic physics of magnetopause reconnection (Burch et al. 2016; Hasegawa et al. 2017; Zhong et al. 2020). Many studies have shown that an initial Harris type equilibrium profile with constant total pressures
and antiparallel magnetic field with or without a guide field (Harris 1962) may tend to develop MR geometry. In particular, two major categories of MR have been proposed: the steady state model with a single X line and the outflow approaching the Alfvén speed (Petschek 1964), and the tearing mode instability with a series of X and O lines and mild plasma velocity (Furth et al. 1963). Numerous fluid and kinetic simulations have been carried out to examine the various aspects of MR processes for the past 50 yr (Hau & Chiou 2001; Guo et al. 2015; Landi et al. 2015). In particular, the effects of pressure or temperature anisotropy on MR have been examined by a number of authors (Chen & Palmadesso 1984; Shi et al. 1987; Birn & Hesse 2001; Chiou & Hau 2002, 2003; Hung et al. 2011). In the MHD models the double-polytropic (DP) laws are widely adopted as the energy closures to study the effects of temperature anisotropy and energy closures on MR and tearing mode instability (Chiou & Hau 2002, 2003; Hung et al. 2011). It is shown that the mirror type temperature anisotropy of
may greatly enhance the growth rate of tearing mode instability and the merging rate of single X-line reconnection. In particular, the coupling of tearing and mirror instabilities may lead to relatively larger magnetic islands as compared to the cases with isotropic pressure and the mirror waves with anticorrelated density and magnetic field may be present in the vicinity of X lines.
|
[
"Zhong et al. 2020"
] |
[
"In particular, the Magnetospheric Multiscale (MMS) mission has contributed greatly to the kinetic physics of magnetopause reconnection"
] |
[
"Background"
] |
[
[
675,
692
]
] |
[
[
498,
632
]
] |
2015MNRAS.450.3458C__Cichowolski_et_al._2001_Instance_7
|
The kinetic energy stored in the CO shell can be estimated as $E_{\rm kin} = 0.5\, M_{\rm shell}\, V^2_{\rm exp}$, where Vexp is the expansion velocity of the shell and Mshell is the total (molecular, atomic, and ionized) shell mass. Adopting an expansion velocity equal to half the velocity interval where the structure is observed, Vexp = 7.0 ± 1.3 km s− 1 , the molecular mass given in Table 1 and the atomic and ionized masses estimated by Cichowolski et al. (2001), 1450 and 3000 M⊙, respectively, we obtain Ekin = (2.5 ± 1.0) × 1049 erg, assuming a 40 per cent error for the masses.. Although Cichowolski et al. (2001) concluded that WR 130 could have alone created the observed structure, it is important to note that they did not take into account the molecular mass present in the shell, which considerably increases the kinetic shell energy. Thus, we can compare now the new value obtained for Ekin with the mechanical energy deposited in the ISM by the wind of the WR star, Ew = (0.7–2.2) × 1050 erg (Cichowolski et al. 2001). We obtain ϵ = Ekin/Ew = 0.007–0.5. The ratio ϵ measures the energy conversion efficiency in the shell, and according to evolutionary models ϵ ≤ 0.2 (Koo & McKee 1992). Thus, not all the possible values of ϵ are compatible with the scenario where the energy injected during the WR phase is enough to create the structure. In this case, the contribution of the energy injected during the O-star phase and/or other massive stars, should be considered. As mentioned in the Introduction, WR 130 is a WNH star, and according to Smith & Conti (2008) its age would be of about 2–3 Myr and its initial mass of at least 60 M⊙. A rough estimation of the energy injected by such a star during its main sequence yields Ew = (2.5–3.5) × 1050 erg (de Jager, Nieuwenhuijzen & van der Hucht 1988), which would be enough to create the observed structure. We have nevertheless looked for the presence of other massive stars in the region. We queried the available catalogues such as the Galactic O-Star Catalog (Maíz Apellániz et al. 2013), the Early-Type Emission-Line Stars Catalogue (Wackerling 1970), the Catalogue of Be stars (Jaschek & Egret 1982), the H-alpha Stars in the Northern Milky Way Catalogue (Kohoutek & Wehmeyer 1997), and the Catalog of Galactic OB Stars (Reed 2003), for early-type and emission stars. No stars were found in any catalogue. The only massive star located nearby is, as mentioned by Cichowolski et al. (2001), an OB star, which has an uncertain spectral type and no distance estimate (Stock, Nassau & Stephenson 1960). It is located in projection not in the centre of the structure but on to the shell (there is a second OB star mentioned by Cichowolski et al. 2001 but its location is actually outside the structure, see fig. 1 of Cichowolski et al. 2001). Although we cannot completely rule out the possibility that the OB star may be playing a role in creating the shell structure, we think that the action of WR 130 is sufficient and most likely dominant in the region.
|
[
"Cichowolski et al. 2001"
] |
[
"see fig. 1 of"
] |
[
"Background"
] |
[
[
2785,
2808
]
] |
[
[
2771,
2784
]
] |
2022MNRAS.509.1504M__Komissarov_2006_Instance_1
|
Energy flows induced into magnetically dominated relativistic magnetospheres of compact objects are commonly modelled by numerical simulations in the force-free electrodynamics (FFE) limit. Fueled by the track record of observations in the era of multimessenger astrophysics, current targets for such simulations include the magnetospheres of rapidly spinning black holes, spiraling neutron stars, magnetars, and pulsars. The tenuous, magnetically dominated atmosphere (magnetosphere) of pulsars is an active field of scientific interest. They fascinate both observers (e.g. Lorimer et al. 1995; Ransom et al. 2005; Abdo et al. 2013; Jankowski et al. 2018) and theorists (e.g. Kennel & Coroniti 1984; Lyubarskii 1996; Contopoulos, Kazanas & Fendt 1999; Goodwin et al. 2004; Timokhin 2006; Timokhin & Arons 2013; Contopoulos 2019; Pétri 2020). With the remarkable progress in scientific computing, their rotating magnetosphere has captured designers of numerical methods that integrate FFE and magnetohydrodynamics (MHD) with ever-improving accuracy (e.g. Komissarov 2006; Spitkovsky 2006; Tchekhovskoy, Spitkovsky & Li 2013; Parfrey, Spitkovsky & Beloborodov 2017; Carrasco & Shibata 2020). Recently, particle-in-cell (PIC) simulations were able to resolve a broad range of scale separations and allow for unprecedented insight into the microphysics of pulsar magnetospheres across the global scale (Cerutti et al. 2015; Philippov, Spitkovsky & Cerutti 2015a; Kalapotharakos et al. 2018; Philippov & Spitkovsky 2018; Guépin, Cerutti & Kotera 2020). In this fascinating flurry of outcomes, only few references scrutinized whether the results from ideal plasma simulations are the best possible model for the pulsar magnetosphere that contains an inherently non-ideal region, namely the equatorial current sheet (ECS) beyond the closed zone (Contopoulos 2016; Contopoulos 2019; Contopoulos, Pétri & Stefanou 2020). Here, we study with rigorous technical depth how this non-ideal region can affect the global dynamics of the force-free aligned rotator magnetosphere – effectively serving as a blueprint for force-free magnetospheres of other compact objects.
|
[
"Komissarov 2006"
] |
[
"With the remarkable progress in scientific computing, their rotating magnetosphere has captured designers of numerical methods that integrate FFE and magnetohydrodynamics (MHD) with ever-improving accuracy (e.g."
] |
[
"Background"
] |
[
[
1055,
1070
]
] |
[
[
843,
1054
]
] |
2018MNRAS.480.1081K__Hunter_et_al._2008_Instance_1
|
Work in recent years has shown that rotation is a key ingredient in shaping the evolution of massive stars with very low metallicities (Z SMC metallicity down to Population III stars; see Meynet & Maeder 2017, and references therein). Fast rotators (with initial vrot ≳300 km s−1) are expected to lead to chemically homogeneous evolution (CHE) in which the star becomes brighter and hotter, and thus more ionizing photons, specially in the extreme UV, are emitted than in the corresponding non-rotating case (e.g. Brott et al. 2011; Levesque et al. 2012; Yoon, Dierks & Langer 2012). Presently, although we still know little about rotation velocities of massive stars and their variation with environment, observations seem to favor fast rotators at low Z (e.g. Martayan et al. 2007; Hunter et al. 2008). Model predictions suggest that the effects of rotation, like CHE, should be enhanced at lower metallicities. There is an increase of theoretical and observational evidence which supports the significant role of rotation among the generations of first stars, with Z=0 or extremely low metallicities, and fast rotating massive stars were likely common phenomena in the early Universe (e.g. Leitherer 2008; Chiappini et al. 2008; Maeder & Meynet 2012; Yoon et al. 2012). Modern models for low-metallicity massive fast-rotating single stars which undergo CHE have been published by Szécsi et al. (2015); the authors called them transparent wind UV intense stars (or TWUIN stars). Considering all this and the extremely low-Z of SBS 0335 − 052E, we compare our observations with the TWUINs predictions,6 following the same approach that we used previously for non-rotating WCE and massive O stellar models. Taking the computed Q(He ii)=7.37 × 1048 photon s−1 for the most massive 294 M⊙ TWUIN, we require that ∼430 (∼300) such stars are necessary to explain the Q(He ii)Int (Q(He ii)$_{\rm He\, \small {II}-MB}$). Again, hundreds of these super-massive TWUINs do not match the M⋆, SSCs of SBS 0335 − 052E. If we instead, apply the same approach using lower mass TWUIN models, it makes even harder to account for the observations. Besides the Q(He ii) budget, we have measured values of He ii λ4686/Hβ as high as 0.06 within the He ii λ4686 main body region. Under ionization-bounded conditions, even the most massive TWUIN models cannot reproduce these values of He ii λ4686/Hβ (see table B.1 from Szécsi et al. 2015).
|
[
"Hunter et al. 2008"
] |
[
"Presently, although we still know little about rotation velocities of massive stars and their variation with environment, observations seem to favor fast rotators at low Z (e.g."
] |
[
"Background"
] |
[
[
785,
803
]
] |
[
[
585,
762
]
] |
2019MNRAS.488.1066C__Link_et_al._1992_Instance_1
|
A pulsar rotates at an angular velocity Ω with inertial moment I and magnetic moment $\boldsymbol {\mathcal {M}}$. It is described by the canonical rotation-down equation in the magnetic dipole model:
(4)
\begin{eqnarray*}
\dot{\Omega } = - K \Omega ^3,
\end{eqnarray*}
where $K = \frac{\mu _0 | \boldsymbol{\mathcal {M}} |^2 \sin ^2 \theta }{6 \pi c^3 I}$, μ0 is vacuum permeability, and c is speed of light. θ is inclination angle, namely, the angle between $\boldsymbol {\mathcal {M}}$ and $\boldsymbol {\Omega }$. The possible change of macroscopic magnetic moment of pulsars at the moment of a glitch occurring was discussed by Link and Epstein (Link et al. 1992; Link & Epstein 1997). Namely, K and Ω changed while the rotation-down equation remains the same at glitches,
(5)
\begin{eqnarray*}
\frac{\Delta \dot{\Omega }}{\dot{\Omega }} \approx \frac{\Delta K}{K} = 2 \frac{\Delta (| \boldsymbol{\mathcal {M}} | \sin \theta)}{| \boldsymbol{\mathcal {M}} | \sin \theta } - \frac{\Delta I}{I} .
\end{eqnarray*}
From observations, we know $\frac{\Delta \dot{\Omega }}{\dot{\Omega }} \sim 10^{-3}$. Merely variation of orbit angular momentum cannot describe the magnetic moment changing, as $- \frac{\Delta I}{I} = \frac{\Delta \Omega }{\Omega } \lesssim 10^{-6} \ll \frac{\Delta \dot{\Omega }}{\dot{\Omega }}$. It motivates us to consider the origin of the equation (5) related with conservation of total angular momentum. As Einstein-de Haas effect suggests that $\boldsymbol {S}$ and $\boldsymbol {\Omega }$ can transfer into each other caused by the total angular momentum conservation,
(6)
\begin{eqnarray*}
\Delta \boldsymbol {J} = \Delta (\boldsymbol {S} + I \boldsymbol {\Omega }) = \boldsymbol {T} \Delta t \approx 0,
\end{eqnarray*}
where $\boldsymbol {S}=(1/\gamma)\boldsymbol {\mathcal {M}}$, Δt is the duration of a glitch, and $\boldsymbol {T}$ is the torque. The total angular momentum conservation would turn to be reasonable, when glitches regarding as instantaneous processes, namely Δt → 0, as the traditional glitch model (Baym et al. 1969) does. In principal, $\boldsymbol {T} \Delta t \ne 0$ is permitted in our model. For the first step, we wish to obtain approximated results with the simplest case of equation (6).
|
[
"Link et al. 1992"
] |
[
"The possible change of macroscopic magnetic moment of pulsars at the moment of a glitch occurring was discussed by Link and Epstein"
] |
[
"Background"
] |
[
[
655,
671
]
] |
[
[
522,
653
]
] |
2016MNRAS.459.1422E__Ebrahimi_&_Bhattacharjee_2014_Instance_1
|
We begin with our results from global DNS MHD simulations of the MRI in cylindrical (r, ϕ, z) geometry using the DEBS (Schnack et al. 1987; Ebrahimi et al. 2009) initial-value code to solve the non-linear, viscous and resistive MHD equations
(1)
\begin{eqnarray}
\frac{\mathrm{\partial} \boldsymbol A }{ \mathrm{\partial} t } &=& -{\boldsymbol E} = S\boldsymbol V\times \boldsymbol B - \eta \boldsymbol J
\end{eqnarray}
(2)
\begin{eqnarray}
\rho \frac{\mathrm{\partial} \boldsymbol V }{ \mathrm{\partial} t } &=& -S \rho \boldsymbol V\cdot \ \nabla \boldsymbol V + S\boldsymbol J \times \boldsymbol B +P_{\rm m} \nabla ^2 \boldsymbol V -S \frac{\beta _0}{2}\nabla P
\end{eqnarray}
(3)
\begin{eqnarray}
\frac{\mathrm{\partial} P }{ \mathrm{\partial} t } &=& -S\nabla \cdot (P \boldsymbol V) - S (\Gamma -1) P \nabla \cdot \boldsymbol V
\end{eqnarray}
(4)
\begin{eqnarray}
\frac{\mathrm{\partial} \rho }{ \mathrm{\partial} t } &=& -S\nabla \cdot (\rho \boldsymbol V)
\end{eqnarray}
(5)
\begin{eqnarray}
\boldsymbol B &=& \nabla \times \boldsymbol A
\end{eqnarray}
(6)
\begin{eqnarray}
\boldsymbol J &=& \nabla \times \boldsymbol B,
\end{eqnarray}
where the variables, ρ, P, V, B, J, and Γ are the density, pressure, velocity, magnetic field, current, and ratio of the specific heats, respectively. We use the same normalization (Schnack et al. 1987; Ebrahimi et al. 2009; Ebrahimi & Bhattacharjee 2014), where time, radius and velocity are normalized to the outer radius a, the resistive diffusion time τR = a2/μ0η, and the Alfvén velocity $V_{\rm A} = B_0/\sqrt{\mu _0 \rho _0}$, respectively. The dimensionless parameters, S = τRVA/a and Pm, are the Lundquist number and the magnetic Prandtl number (the ratio of viscosity to resistivity), respectively. the initial state satisfies the equilibrium force balance condition $\frac{\beta _0}{2}\nabla p = \rho V_{\phi }^2/r$, where $\beta _0 \equiv 2 \mu _0 P_0/B_0^2$ is normalized to the axis value, and the initial pressure and density profiles are assumed to be radially uniform and unstratified. Pressure and density are evolved, however, they remain fairly uniform during the computations. A no-slip boundary condition is used for the poloidal flow and flow fluctuations. The inner and outer radial boundaries are perfectly conducting so that the tangential electric field, the normal component of the magnetic field, and the normal component of the velocity vanish. The tangential component of the velocity is the rotational velocity of the wall. The azimuthal (ϕ) and axial (z) boundaries are periodic. We assume a radial pressure gradient balances the centrifugal force in equilibrium, but radial gravity and a radial pressure force are interchangeable for our incompressible, unstratified circumstance. The pressure gradient, rather than gravity, is what balances the centrifugal force in cylindrical laboratory experiments designed to test the MRI (Goodman & Ji (2002)).
|
[
"Ebrahimi & Bhattacharjee 2014"
] |
[
"We use the same normalization",
"where time, radius and velocity are normalized to the outer radius a, the resistive diffusion time τR = a2/μ0η, and the Alfvén velocity $V_{\\rm A} = B_0/\\sqrt{\\mu _0 \\rho _0}$, respectively."
] |
[
"Uses",
"Uses"
] |
[
[
1436,
1465
]
] |
[
[
1362,
1391
],
[
1468,
1658
]
] |
2021MNRAS.500.1772N__Siegel_&_Ciolfi_2015_Instance_1
|
While these early studies demonstrated the viability of neutron star mergers as a major r-process site, they identified only one ejection channel: ‘dynamical ejecta’ that are tidally flung out by gravitational torques. Since they are never substantially heated, these ejecta carry their original β −equilibrium electron fraction from the original neutron star, Ye ≈ 0.05, and this enormous neutron-richness allows them to undergo a ‘fission cycling’ process (Goriely, Bauswein & Janka 2011; Korobkin et al. 2012), which produces a very robust r-process abundance distribution close to the solar pattern for A ≥ 130, but hardly any lighter r-process elements. Oechslin, Janka & Marek (2007) pointed out that there is a second channel of mass ejection that also happens on a dynamical time-scale: shock-heated matter from the interface where the stars come into contact. As of today, many more mass ejection channels have been discussed: matter that becomes unbound on secular time-scales (∼1 s) from the post-merger accretion torus (Beloborodov 2008; Metzger, Piro & Quataert 2008; Fernandez & Metzger 2013; Fernandez et al. 2015; Just et al. 2015; Siegel & Metzger 2017, 2018; Fernandez et al. 2019; Miller et al. 2019a), as MHD-driven winds (Siegel & Ciolfi 2015) and by viscous effects (Shibata, Kiuchi & Sekiguchi 2017; Radice et al. 2018a; Shibata & Hotokezaka 2019) from a long-lived neutron star merger remnant. Similar to the case of proto-neutron stars, the enormous neutrino luminosities (>1053 erg s−1) after a neutron star merger can also drive substantial matter outflows (Ruffert et al. 1997; Rosswog & Ramirez-Ruiz 2002; Dessart et al. 2009; Perego et al. 2014; Martin et al. 2015; Radice et al. 2018b). The secular torus ejecta contain approximately 40 per cent of the initial torus mass and, although the latter may vary substantially from case to case, they likely contribute the lion’s share to the total ejecta mass. Due to their different thermal histories and exposure times to neutrinos, the ejecta channels can have different electron fractions Ye and therefore different nucleosynthesis yields.1 For electron fractions below a critical value, $Y_{\rm e}^{\rm crit}\approx 0.25$ (Korobkin et al. 2012; Lippuner & Roberts 2015), lanthanides and actinides are efficiently produced, which, due to their open f-shells, have particularly high bound–bound opacities (Barnes & Kasen 2013; Kasen, Badnell & Barnes 2013; Tanaka & Hotokezaka 2013; Tanaka et al. 2020) and therefore lead to red transients that peak days after the merger. Ejecta with electron fractions above $Y_{\rm e}^{\rm crit}$, in contrast, only produce ‘lighter’ elements with lower opacities and thus result in bluer transients that peak after about 1 d. Opaque, low-Ye ejecta blocking the view on high-Ye ejecta can lead to a ‘lanthanide curtaining’ effect (Kasen, Fernández & Metzger 2015; Wollaeger et al. 2018), which will efficiently block blue light. Therefore, it is important to understand the layering, dynamics, interaction and potential mixing of different ejecta channels.
|
[
"Siegel & Ciolfi 2015"
] |
[
"As of today, many more mass ejection channels have been discussed:",
"as MHD-driven winds"
] |
[
"Background",
"Background"
] |
[
[
1243,
1263
]
] |
[
[
869,
935
],
[
1222,
1241
]
] |
2016AandA...586A..80O__Fornasier_et_al._2015_Instance_3
|
Figure 1 shows that in the regions where activity was detected visually, i.e., Hapi, Seth, and Ma’at pits have lower (8–13%/100 nm) spectral slopes than the rest of the comet surface (13–22%/100 nm). In addition to those places, Seth alcoves, the wall of the large Anuket alcove, around the circular features, both clustered and isolated bright features (see Thomas et al. 2015b; Auger et al. 2015; Pommerol et al. 2015b, for definitions) show similar lower spectral slopes than the rest of the surface, even though there was no visual detection of activity features rising from them at the time of the observations used in this study4. This may be because the observing geometry was not suited for their detections during the observations. In the regions we investigated, the Hapi region displays the lowest spectral slopes 8–11%/100 nm (see also Fornasier et al. 2015) together with the isolated bright features (IBFs) and the clustered bright features in the Imhotep region. The locations of the bright features on the Imhotep image (image #4) are shown in Fig. B.4. According to the spectral slope values, the IBFs of Imhotep seem to be more similar to the Hapi region than the active pits of Seth and Ma’at regions. Active pits, alcoves, and the large alcove of Anuket have slope values of typically 10–13%/100 nm. The Ma’at region, which is located on the smaller lobe (head) of the comet, displays higher spectral slope values than the Seth region, which is located on the larger (body) lobe of the comet. In the investigated regions, the highest slope values are detected in the Imhotep region (see Fig. 1d). Here it should be mentioned that the comparison of spectral slopes is performed under the assumption of no spectral reddening between the phase angles of the images we investigated, although the spectral slopes show reddening by phase as presented in Fornasier et al. (2015). Unfortunately, the previous work does not cover all the phase angles of the images we investigated, but the spectral slope variation between 35–54° (Fig. 3 of Fornasier et al. 2015) is small so that we can make this comparison. However, if we follow the linear trend of the phase reddening, for the image taken in 70.45° phase angle (image #4), the spectral slopes would vary from 15%/100 nm to 18%/100 nm in the observations we used.
|
[
"Fornasier et al. 2015"
] |
[
"Unfortunately, the previous work does not cover all the phase angles of the images we investigated, but the spectral slope variation between 35–54° (Fig. 3 of",
"is small so that we can make this comparison. However, if we follow the linear trend of the phase reddening, for the image taken in 70.45° phase angle (image #4), the spectral slopes would vary from 15%/100 nm to 18%/100 nm in the observations we used."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
2053,
2074
]
] |
[
[
1893,
2051
],
[
2076,
2328
]
] |
2020MNRAS.492.3241V__Frebel_et_al._2006_Instance_1
|
Progress in this field will require large statistical samples of metal-poor stars in a variety of environments within the Local Group. Unfortunately, metal-poor stars are exceedingly rare and difficult to find, being overwhelmed by the more numerous metal-rich populations in the Galaxy. Examination of the Besançon model of the Galaxy (Robin et al. 2003), which is guided by a theoretical framework for the formation and evolution of the main stellar populations, suggests that a typical halo field has only one in ∼2000 stars with [Fe/H] −3 between 14 V 18 (Youakim et al. 2017). Enormous effort has gone into the discovery and study of extremely, ultra, and hyper metal-poor stars with [Fe/H] −3.0, −4.0, and −5.0, respectively. Most of the known metal-poor stars have been found in dedicated surveys, such as objective prism surveys (the HK survey and Hamburg-ESO survey, Beers, Preston & Shectman 1992; Christlieb et al. 2002, 2008; Beers & Christlieb 2005; Frebel et al. 2006; Schörck et al. 2009), wide-band photometric surveys (Schlaufman & Casey 2014), and blind spectroscopic surveys, such as the Sloan Digital Sky Survey (SDSS) SEGUE and BOSS surveys (Yanny et al. 2009; Eisenstein et al. 2011; Dawson et al. 2013), and from the Large Sky Area Multi-Object Fibre Spectroscopic Telescope (LAMOST; Cui et al. 2012). According to the SAGA data base (see Suda et al. 2017, and references therein), there are ∼500 stars with [Fe/H] −3.0, though fewer than half have detailed chemical abundances. Recently, narrow-band photometric surveys have shown higher success rates for finding metal-poor stars, particularly SkyMapper (Keller et al. 2007; DaCosta et al. 2019) and the Pristine survey (Starkenburg et al. 2017b; Youakim et al. 2017; Aguado et al. 2019a). Pristine photometry with follow-up Keck II/DEIMOS spectroscopy has also been used to increase sample sizes and improve the chemodynamical studies of faint satellites (Draco II and Sgr II, Longeard et al. 2018, 2019). At the same time, Simon (2018) has shown that Gaia DR2 proper motion cleaning may also be a promising way to find new metal-poor members of ultra-faint dwarf galaxies.
|
[
"Frebel et al. 2006"
] |
[
"Most of the known metal-poor stars have been found in dedicated surveys, such as objective prism surveys (the HK survey and Hamburg-ESO survey"
] |
[
"Background"
] |
[
[
967,
985
]
] |
[
[
736,
878
]
] |
2018ApJ...860...88B__Paradijs_1978_Instance_1
|
Thermonuclear (type-I) X-ray bursts are intermittently observed from many neutron star low-mass X-ray binaries (LMXBs; Strohmayer & Bildsten (2006) and references therein). Such a burst originates from an unstable thermonuclear burning of the accreted matter accumulated on the neutron star surface (Joss 1977; Lamb & Lamb 1978; Strohmayer & Bildsten 2006). For most bursts, the observed X-ray intensity rises in ≈0.5–5 s, decays in ∼10–100 s as the neutron star surface cools down after the nuclear burning, and the typical recurrence time is a few hours to days (Galloway et al. 2008). The burst spectrum is traditionally described with a blackbody model, and the best-fit burst blackbody normalization, which can be identified as the burst emission area, usually matches well with the expected surface area of a neutron star (Hoffman et al. 1977; Swank et al. 1977). These motivated an effort to measure the neutron star radius using the burst continuum spectrum, where the normalization of the burst blackbody is expected to be proportional to the square of the stellar radius (e.g., van Paradijs 1978; Goldman 1979; van Paradijs 1979; van Paradijs & Lewin 1986). Note that such a radius measurement is extremely important to probe the superdense and degenerate core matter of neutron stars, which is a fundamental problem of physics (e.g., Lattimer & Prakash 2007; Bhattacharyya et al. 2017). However, a reliable radius measurement using this method has so far not been possible due to a number of systematic uncertainties, which include (1) burst emission from and the visibility of an unknown fraction of the neutron star surface and (2) plausible deviation of the burst spectrum from a blackbody, etc. (Bhattacharyya 2010; Bhattacharyya et al. 2010; Kajava et al. 2017a). Nevertheless, the use of continuum burst spectrum remains a promising method to measure the neutron star radius for the following reasons: (1) neutron star LMXBs provide many complementary methods to measure neutron star parameters, the joint application of which has a potential to significantly reduce the systematics (Bhattacharyya 2010) and (2) it was generally believed that the much more intense burst emission could be reliably distinguished from the persistent (i.e., non-burst) emission during the burst.
|
[
"van Paradijs 1978"
] |
[
"These motivated an effort to measure the neutron star radius using the burst continuum spectrum, where the normalization of the burst blackbody is expected to be proportional to the square of the stellar radius (e.g.,"
] |
[
"Background"
] |
[
[
1088,
1105
]
] |
[
[
870,
1087
]
] |
2019AandA...632A.104G__Hirabayashi_et_al._2016_Instance_1
|
Finally, our observations are consistent with the bilobate shape of the nucleus of comet 8P/Tuttle. As noted in Sect. 1, this shape is likely common among comets because it was found for four out of the six comets for which we have spatially resolved images. This is also the case of the trans-Neptunian object 2014 MU69 (Ultima Thule) observed by the New Horizon spacecraft (Stern et al. 2019). This binary configuration has some implications for the formation and evolution of 8P/Tuttle. A contact binary could result from (i) the accretion at low velocity of two primordial objects (Massironi et al. 2015; Davidsson et al. 2016), (ii) the disruption of a monolithic object due to excessive spin-up resulting from non-gravitational forces or YORP5 effect followed by a reaccretion (Boehnhardt 2004; Ćuk 2007; Hirabayashi et al. 2016), or (iii) the catastrophic disruption of a monolithic object by a collision followed by a re-accretion (Jutzi & Benz 2017; Schwartz et al. 2018). On the one hand, with a low thermal inertia compared with NEAs, the YORP effect is low for comets, in particular for NIC, which have an elongated orbit and spend most of their time far from the Sun, and it may not be sufficient to increase the spin rate of the nucleus to the point where centrifugal exceed gravitational forces. On the other hand, comet 8P/Tuttle has been on a very stable orbit for centuries, and it is likely an evolved comet, as suggested by its low activity, so that it could have been much more active in the past. For cometary nuclei, the primary cause for spin-up is torques caused by outgassing, therefore it is possible that 8P/Tuttle formed as a monolithic body and became a contact binary after its injection into the inner Solar System as a result of excessive spin-up resulting from non-gravitational forces. This scenario has been proposed for comet 67P/Churyumov-Gerasimenko by Hirabayashi et al. (2016). Alternatively, if the binary nature of comet 8P/Tuttle is the result of a primordial accretion or a catastrophic collision in the early Solar Sytem, it could have persisted until now. Similar examples are offered by some binary asteroids that can be stable over the age of the Solar System (Chauvineau et al. 1991), or as proposed by Davidsson et al. (2016) for comet 67P/Churyumov-Gerasimenko. For comet 8P/Tuttle, it is however not possible to distinguish the solution of a binary nucleus that formed in the first billion years of our Solar System (e.g., Matonti et al. 2019) from a more recent origin following its injection into the inner Solar System (e.g., Hirabayashi et al. 2016).
|
[
"Hirabayashi et al. 2016"
] |
[
"A contact binary could result from",
"(ii) the disruption of a monolithic object due to excessive spin-up resulting from non-gravitational forces or YORP5 effect followed by a reaccretion"
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
811,
834
]
] |
[
[
490,
524
],
[
633,
782
]
] |
2018ApJ...867..120H__Takahashi_et_al._1990_Instance_1
|
Although the magnetic (i.e., one-photon) pair production is also taken into account, most electron–positron pairs are found to be produced via photon–photon (i.e., two-photon) collisions, which take place via two paths. One path is through the collisions of two MeV photons, both of which were emitted from the equatorial ADAF. Another path is through the collisions of TeV and eV photons; the former photons were emitted by the gap-accelerated leptons via inverse Compton process, while the latter were emitted from the ADAF via the synchrotron process. There is, indeed, a third path, in which the gap-emitted GeV curvature photons collide with the ADAF-emitted keV inverse Compton photons; however, this path is negligible, particularly when
. If the pairs are produced via TeV–eV collisions (i.e., via the second path) outside the gap outer boundary, they have outward ultrarelativistic momenta to easily “climb up the hill” of the potential k0 (Takahashi et al. 1990; see also Figure 2 of HP16) and propagate to large distances without turning back. However, if the pairs are produced via MeV–MeV collisions (i.e., via the first path), they are produced with subrelativistic outward momenta; thus, they eventually return to fall onto the horizon owing to the strong gravitational pull inside the separation surface (Figure 2 of HP16). When the returned pairs arrive at the gap outer boundary, only positrons can penetrate into the gap because of E∥ 0. Accordingly, electrons accumulate at the boundary, whose surface charge leads to the jump of the normal derivative of E∥. Thus, although the stationary gap solutions show that the γ-ray spectrum little depends on the injected current density (Section 5.5), the gap solution inevitably becomes time dependent owing to the increasing discontinuity of
with an accumulated surface charge (in this case, electrons) at the outer boundary. If the injected current is much small compared to the GJ current, the time dependence will be mild. However, if the injected current becomes a good fraction of the GJ current, the assumption of the stationarity becomes invalid, as pointed out by Levinson & Segev (2017). In this sense, a caution should be made in the applicability of the stationary solutions presented in this paper, when the injected current is non-negligible compared to the current created within the gap.
|
[
"Takahashi et al. 1990"
] |
[
"If the pairs are produced via TeV–eV collisions (i.e., via the second path) outside the gap outer boundary, they have outward ultrarelativistic momenta to easily “climb up the hill” of the potential k0",
"see also Figure 2 of HP16) and propagate to large distances without turning back."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
956,
977
]
] |
[
[
753,
954
],
[
979,
1060
]
] |
2021AandA...655A..12T__Tang_et_al._2017b_Instance_3
|
Using the RADEX3 non local thermodynamic equilibrium (LTE) modeling program (van der Tak et al. 2007) with collisional rate coefficients from Wiesenfeld & Faure (2013), we modeled the relation between the gas kinetic temperature and the measured average of para-H2CO 0.5 × [(322–221 + 321–220)/303–202] ratios, adopting a 2.73 K background temperature, an average observational linewidth of 4.0 km s−1, and column densities N(para-H2CO) = 2.7 × 1012 and 3.7 × 1012 cm−2 for N113 and N159W, respectively. The results are shown in Fig. 5. The values of the para-H2CO column density were obtained with APEX data (beam size ~30″; Tang et al. 2017b), which cover similar regions. Different column densities of para-H2CO only weakly affect derived kinetic temperatures (see Fig. 3 in Tang et al. 2017b or Fig. 4 in Tang et al. 2018a; this was also shown in Fig. 13 and discussed in Sect. 4.3.1 of Mangum & Wootten 1993) as long as all lines are optically thin. Considering that the relation between the gas temperature and the para-H2CO line ratio may vary at different spatial densities (see Fig. 2 in Tang et al. 2017b), we modeled it at spatial densities 104, 105, and 106 cm−3 in Fig. 5. It appears that Tkin at n(H2) = 105 cm−3 is consistently lower than values at 104 and 106 cm−3 by ≲23% and ≲34%, respectively, for Tkin ≲ 100 K. Local thermodynamic equilibrium (LTE) is a good approximation for the H2CO level populations under optically thin and high-density conditions (Mangum & Wootten 1993; Tang et al. 2017a,b, 2018b). Following the method applied by Tang et al. (2017b) in their Eq. (2), we plot the relation between the LTE kinetic temperature, TLTE, and the para-H2CO (3–2) line ratio in Fig. 5. Apparently, TLTE agrees well with Tnon-LTE at volume densities n(H2) ~ 105 cm−3 as long as Tkin ≲ 100 K. Previous observations show that para-H2CO (3–2) is sensitive to gas temperature at density 105 cm−3 (Ginsburg et al. 2016; Immer et al. 2016; Tang et al. 2017b). The spatial density measured with para-H2CO (303–202) and C18O (2–1) in N113 and N159W is n(H2) ~ 105 cm−3 on a size of ~30″ (Tang et al. 2017b). Therefore, here we adopt 105 cm−3 as an averaged spatial gas density in the N113 and N159W regions.
|
[
"Tang et al. 2017b"
] |
[
"Considering that the relation between the gas temperature and the para-H2CO line ratio may vary at different spatial densities (see Fig. 2 in",
"we modeled it at spatial densities 104, 105, and 106 cm−3 in Fig. 5."
] |
[
"Uses",
"Uses"
] |
[
[
1097,
1114
]
] |
[
[
955,
1096
],
[
1117,
1185
]
] |
2020MNRAS.492.5675C__Pérez-Montero_et_al._2019_Instance_2
|
In regarding AGNs, the Te method tends to underestimate the oxygen abundance by an average value of about 0.6 dex in comparison to estimations based on strong-line methods and it produces subsolar O/H values for most of these objects (Dors et al. 2015; Dors, Freitas-Lemes & Âmores 2020). An alternative method to derive the metallicity or abundances in the nuclear regions of spiral galaxies is the extrapolation of the radial oxygen abundance. Along decades, results based on this indirect method have indicated Z near or slightly above the solar value in nuclear regions (Vila-Costas & Edmunds 1992; Zaritsky, Kennicutt & Huchra 1994; van Zee et al. 1998; Pilyugin, Vílchez & Contini 2004; Gusev et al. 2012; Dors et al. 2015; Zinchenko et al. 2019), in consonance with predictions of chemical evolution models (e.g. Mólla & Díaz 2005) and with the use of strong-line methods (e.g. Groves, Dopita & Sutherland 2004; Groves, Heckman & Kauffmann 2006; Feltre, Charlot & Gutkin 2016; Pérez-Montero et al. 2019; Thomas, Kewley & Dopita 2019; Dors et al. 2020). Therefore, Temethod does not seem to work for AGNs. The origin of the discrepancy between Z values calculated via Te method and via strong-line methods, the so-called Teproblem, could be attributed, in part, to the presence of heating/ionization by gas shock in the narrow-line region (NLR) of AGNs. In fact, Contini (2017) carried out detailed modelling of AGN optical emission lines by using the SUMA code (Contini & Aldrovandi 1983) and suggested the presence of gas shock with low velocity ($v \: \lesssim \: 400 \: \rm km \:s^{-1}$) in a sample of Seyfert 2 nuclei. This result is supported by recent spatially resolved observational studies of Seyfert 2 nuclei, in which the presence of gas outflows with velocity of the order of 100–300 $\rm km \: s^{-1}$ have been found (e.g. Riffel, Storchi-Bergmann & Riffel 2017; Riffel, Hekatelyne & Freitas 2018). Moreover, the Te problem can also be originated due to the use of an unappropriate calculation of the ionization correction factor (ICF) for oxygen in AGNs (Pérez-Montero et al. 2019; Dors et al. 2020).
|
[
"Pérez-Montero et al. 2019"
] |
[
"Moreover, the Te problem can also be originated due to the use of an unappropriate calculation of the ionization correction factor (ICF) for oxygen in AGNs"
] |
[
"Compare/Contrast"
] |
[
[
2078,
2103
]
] |
[
[
1921,
2076
]
] |
2021MNRAS.500.3957E__Springel_&_Hernquist_2003_Instance_1
|
The results presented in this paper are based on data from IllustrisTNG,1the next-generation suite of state-of-the-art magnetohydrodynamical cosmological simulations of galaxy formation (Marinacci et al. 2018; Naiman et al. 2018; Nelson et al. 2018; Pillepich et al. 2018b; Springel et al. 2018). Building on the success of its predecessor Illustris (Genel et al. 2014; Vogelsberger et al. 2014a,b; Nelson et al. 2015; Sijacki et al. 2015), IllustrisTNG follows the same fundamental approach but includes improved aspects and novel features in its galaxy formation model and expands its scope to several simulated volumes and improved resolution. The models for galaxy formation include physical processes such as gas heating by a spatially uniform and time-dependent UV background, primordial and metal-line gas cooling, a subgrid model for star formation, and the unresolved structure of the interstellar medium (Springel & Hernquist 2003), as well as models for the evolution and chemical enrichment of stellar populations, which track nine elements (H, He, C, N, O, Ne, Mg, Si, and Fe) in addition to europium and include yields from supernovae Ia, II, and AGB stars (Vogelsberger et al. 2013; Torrey et al. 2014). Furthermore, IllustrisTNG incorporates improved feedback implementations for galactic winds caused by supernovae as well as accretion and feedback from black holes. In particular, depending on accretion, black hole feedback occurs in two modes: low accretion rates result in purely kinetic feedback while high accretion rates invoke thermal feedback (Weinberger et al. 2017). Galactic winds are injected isotropically and the wind particles’ initial speed scales with the one-dimensional dark matter velocity dispersion (Pillepich et al. 2018a). Magnetic fields are amplified self-consistently from a primordial seed field and follow ideal magnetohydrodynamics (Pakmor & Springel 2013). The TNG simulations were run using the moving mesh code arepo (Springel 2010). Here, concepts from adaptive mesh refinement and smooth particle hydrodynamics are combined to create an unstructured, moving Voronoi tessellation. IllustrisTNG follows the ΛCDM framework, adopting cosmological parameters according to recent constraints from Planck data: matter density $\Omega _\rm {m} = 0.3089$, baryonic density $\Omega _\rm {b} = 0.0486$, cosmological constant $\Omega _\Lambda = 0.6911$, Hubble constant h = 0.6774, normalization σ8 = 0.8159, and spectral index $n_\rm {s} = 0.9667$ (Planck Collaboration I 2016).
|
[
"Springel & Hernquist 2003"
] |
[
"The models for galaxy formation include physical processes such as gas heating by a spatially uniform and time-dependent UV background, primordial and metal-line gas cooling, a subgrid model for star formation, and the unresolved structure of the interstellar medium"
] |
[
"Uses"
] |
[
[
915,
940
]
] |
[
[
647,
913
]
] |
2017AandA...607A.126Y__Gunawardhana_et_al._2011_Instance_1
|
Observed high mass end power-law index of the galaxy-wide IMF resulting from the calculated IGIMF, \hbox{$\alpha_3^{\mathrm{gal}}$}α3gal (i.e., αgal in Eq. (15) for m> 1 M⊙), for a constant SFR over δt = 10 Myr in dependence of the galaxy-wide SFR. In Fig. 5, \hbox{$\alpha_3^{\mathrm{gal}}$}α3gal values diverge for different SFRs and also vary for different m at the high mass end. As at each m value there exists a different \hbox{$\alpha_3^{\mathrm{gal}}$}α3gal–SFR relation, we plot solid lines for log 10(m/M⊙) = 0.2, 0.4, ..., 2, i.e., 1.58, 2.51, ..., 100 M⊙ from black to gray (top to bottom) for the fiducial model and dotted lines for the corresponding SolarMetal model defined in Sect. 2.1. The blue squares are data from the GAMA galaxy survey (Gunawardhana et al. 2011). The red triangles and red dash-dotted line are data from Weidner et al. (2013b); the left triangle is for the MW field, the middle three triangles are galaxy studies, the right triangle is for the bulges of the MW and M 31, and the dash-dotted line is their IGIMF model assuming β = 2. A recent study has suggested that the 2 M⊙/yr SFR for MW is overestimated (Chomiuk & Povich 2011), but we leave this data point the same as in Weidner et al. (2013b). Gargiulo et al. (2015) report consistency between their IGIMF model assuming β = 2 (thick yellow dashed line) and the [α/Fe] abundance ratios of elliptical galaxies. The purple diamond is an individual analysis for the dwarf galaxy NGC 2915 (Bruzzese et al. 2015). Green stars are based on the Lee et al. (2009) 11HUGS observations of dwarf galaxies. The black circle is an observation for the solar neighborhood from Rybizki & Just (2015) with adopted MW SFR from Robitaille & Whitney (2010) as an upper limit of the solar neighborhood SFR because the Sun is located in an inter-arm region where the relevant SFR is significantly smaller (toward the direction indicated by the arrow; see Sect. 4.6 for further details). The thin horizontal dashed line represents the canonical IMF index α2 = α3 = 2.3.
|
[
"Gunawardhana et al. 2011"
] |
[
"The blue squares are data from the GAMA galaxy survey"
] |
[
"Uses"
] |
[
[
758,
782
]
] |
[
[
703,
756
]
] |
2015AandA...576L..16P__Li_et_al._(2013)_Instance_1
|
To constrain the size of outflows that we could have missed, we performed simple simulations. We placed artificial, unipolar secondary sources next to a primary point source model representing Sgr A* and compared the closure phases obtained from the resulting artificial visibility data with the observations. We considered two geometries: a single-point source and a jet composed of ten equally spaced point sources (knots) with equal fluxes. We probed four orientations for the simulated outflows (see Fig. 2): along the major axis, along the minor axis of the beam, the jet direction claimed by Li et al. (2013), and the jet direction claimed by Yusef-Zadeh et al. (2012). We used total fluxes of 0.2 Jy and 0.55 Jy for the artificial sources; these values ensure that our simulated outflows are sufficiently faint to not violate the constraints given by the known recent brightness evolution of Sgr A* (0.2 Jy from the mean variability of ≈15% from June 2013 to February 2014 at 41 GHz, with 0.55 Jy corresponding to the strongest variation in the same period, Chandler & Sjouwerman 2014). For each simulation setup, we measured the average of absolute values of the closure phases for each triangle. We varied the distances of the model sources (for the jet model: the largest distance) from Sgr A* until we found a critical distance at which the absolute values of the simulated closure phases exceeded those of the observations by more than the 1σ error at all triangles. We summarize our results in Table 1. As expected, the critical distances are smaller for brighter outflows. Jet-like structures lead to larger critical distances than equally luminous single, compact sources. As a consequence of the very elongated beam, the critical distances for sources located along the major axis of the beam are larger by a factor of ≈7 than for those located along the minor axis. In a few cases (denoted “N/A” in Table 1), the absolute values of the simulated closure phases were similar to those of the observations for all distances of the model sources, meaning that we were unable to identify a critical distance. Overall, our observations limit the extension of asymmetric (in the observer frame) jet-like outflows from Sgr A* to projected distances of ≈2.5 mas along the major axis and ≈0.4 mas along the minor axis.
|
[
"Li et al. (2013)"
] |
[
"We probed four orientations for the simulated outflows",
"the jet direction claimed by"
] |
[
"Uses",
"Uses"
] |
[
[
598,
614
]
] |
[
[
444,
498
],
[
569,
597
]
] |
2020MNRAS.491.5881Y__McConnell_et_al._2011_Instance_1
|
There is good observational and theoretical evidence that supermassive black holes (SMBHs) exist in nearly every galaxy in universe. Understanding the properties of these SMBHs will clarify their roles in galaxy formation and evolution across the cosmology history (e.g. Kormendy & Ho 2013). There are mainly two parameters for an SMBH, i.e. mass (MBH) and spin, which need to be determined. For a few very nearby (100 Mpc) quiescent galaxies, including our Galaxy, SMBH masses can be measured through the stellar or gaseous dynamics method (e.g. Tremaine et al. 2002; McConnell et al. 2011). It has been found that nearby quiescent galaxies follow a tight correlation between the central SMBH mass and the bulge or spheroid stellar velocity dispersion (σ*), which is called MBH–σ* relation (e.g. Kormendy & Ho 2013). Active galactic nuclei (AGNs) can be classified into type 1 or type 2 AGNs, depending on whether the broad-line regions (BLRs) can be viewed directly. For type 1 AGNs, the BLR can be used as a probe of the gravitational potential of the SMBHs. The SMBH mass can be weighed through the BLR clouds for type I AGNs across cosmos time. The SMBH masses in type I AGNs can be calculated as follows (e.g. Kaspi et al. 2000; Bian & Zhao 2002; Peterson et al. 2004; Collin et al. 2006; Du et al. 2016a; Yu et al. 2019):
(1)$$\begin{eqnarray*}
M_{\rm BH} =f_{\rm BLR}\frac{R_{\rm BLR}~(\Delta V)^2}{G},
\end{eqnarray*}$$where G is the gravitational constant. RBLR is the distance from black hole to the BLRs, and can be estimated from the reverberation mapping (RM) method (e.g. Blandford & McKee 1982; Peterson 1993). ΔV is the velocity of the BLR clouds, and usually traced by the full width at half-maximum (FWHM) or the line dispersion (σH β) of the broad H β emission line. fBLR is a virial factor to characterize the kinematics, geometry, and inclination of the BLR clouds. Using the MBH–σ* relation, we recently did the calibration of fBLR and found $f_{\rm BLR} \propto \rm FWHM^{-1.11}$ when FWHM(H β) is used as the tracer of ΔV in equation (1) (Mejia-Restrepo et al. 2018; Yu et al. 2019). It is consistent with the results by the BLR dynamical model to fit simultaneously the AGNs continuum/H β light curves and H β line profiles (e.g. Li et al. 2018; Pancoast et al. 2018; Williams et al. 2018).
|
[
"McConnell et al. 2011"
] |
[
"For a few very nearby (100 Mpc) quiescent galaxies, including our Galaxy, SMBH masses can be measured through the stellar or gaseous dynamics method (e.g."
] |
[
"Background"
] |
[
[
569,
590
]
] |
[
[
392,
546
]
] |
2016ApJ...831...37K__Dadina_2008_Instance_1
|
Based on previous works (Kawamuro et al. 2013; Tazaki et al. 2013), we start with a base-line model,
in the XSPEC terminology. This model includes absorbed primary X-ray emission (i.e., a cut-off power law), a scattered component, and a reflection continuum from distant cold matter accompanied by a narrow iron-Kα line. Optically thin thermal emission from the host galaxy (apec in XSPEC) and other emission/absorption lines (zgauss) are also added if they are significantly required with a confidence level above 90% in terms of Δχ2. Because it is difficult to determine the cut-off energy from our data, we fix it at 300 keV, which is a typical value measured in nearby AGNs (Dadina 2008). The first constant factor, NXIS, is applied to the primary power-law component in the Suzaku spectra to absorb possible time variability between the Suzaku (one epoch) and Swift/BAT (averaged for 70 months) observations. The second constant term represents the scattered fraction, fscatt. As a reflection component from the torus, we employ the pexrav model, which calculates a reflected spectrum from an optically thick slab with a solid angle of Ω irradiated by a point source (Magdziarz & Zdziarski 1995). We set the reflection strength, R = Ω/2π, as a free parameter, and fix the inclination angle at 60°. It is confirmed that even if 30° is adopted, best-fit parameters do not significantly change. The shape of the incident spectrum is assumed to be the same as the power-law component. We basically assume that the reflection and scattered components did not vary between the Suzaku and Swift/BAT observations, considering that the size of the reflector most likely has a parsec scale. In the three low-mass LLAGNs (NGC 4395, NGC 5273, and NGC 5643), however, we assume that the reflection component varied in accordance with the primary emission because of a smaller size of the emitting regions. Thus, in these targets, R is defined with respect to the primary component in the Suzaku data. The zgauss component represents an iron-Kα fluorescence line. The line width is fixed at 20 eV, which corresponds to a typical velocity dispersion of ∼2000 km s−1 measured with Chandra/HETGS in local Seyfert galaxies (Shu et al. 2010). We always consider the Galactic absorption
, which is calculated with the nh command (Kalberla et al. 2005) in FTOOLS.
|
[
"Dadina 2008"
] |
[
"Because it is difficult to determine the cut-off energy from our data, we fix it at 300 keV, which is a typical value measured in nearby AGNs"
] |
[
"Uses"
] |
[
[
684,
695
]
] |
[
[
541,
682
]
] |
2019ApJ...875...61M__Marco_2006_Instance_1
|
A substantial fraction of metal-poor stars that have recently evolved off the MS, e.g., giants and planetary nebulae (PNe), have been influenced by binary interactions. The IMF is significantly weighted toward low-mass stars (Bastian et al. 2010; Kroupa et al. 2013), and the MW star formation rate was ≈3 times larger ≈10 Gyr ago than it is now (Governato et al. 2007; De Lucia et al. 2014). Based on the measured IMF and modeled galactic star formation history, we estimate that ≈55% of MW giants and PNe have old, solar-type progenitors (τ* > 7 Gyr, M ≈ 0.8–1.2
). Such old, low-mass giants tend to be metal-poor (Ratnatunga & Yoss 1991; Carollo et al. 2010; Mackereth et al. 2017). The metallicity trend therefore dramatically affects the properties of low-mass evolved stars. For example, the enhanced close binary fraction of metal-poor solar-type stars substantially strengthens the conclusion that the shaping of PN morphologies is the result of binary interactions (Moe & De Marco 2006; De Marco 2009; Jones & Boffin 2017). Providing further corroboration, Badenes et al. (2015) measured the delay-time distribution of bright PNe in the LMC and discovered two distinct populations of PN progenitors: an old channel (τ* = 5–8 Gyr) deriving from solar-type stars (M ≈ 1.0–1.2
) and a young channel (35–800 Myr) evolving from late-B/early-A stars (≈2–8
). According to the measured age–metallicity relation of the LMC (Olszewski et al. 1991; Pagel & Tautvaisiene 1998; Cole et al. 2005; Carrera et al. 2011; Piatti & Geisler 2013), the old, solar-type progenitors are metal-poor ([Fe/H] ≲ −1.0) and hence have a large close binary fraction of Fclose = 40%–50%. The young progenitors have a higher metallicity of [Fe/H] ≈ −0.4 but are sufficiently massive so that they also have a large close binary fraction of Fclose = 40%–60%. Meanwhile, evolved stars with intermediate masses (M ≈ 1.2–2.0
) in the LMC have intermediate metallicities and therefore a smaller close binary fraction of Fclose ≈ 30%. If PNe derive from interactions in close binaries, then the variations in Fclose with respect to mass and metallicity can explain the observed bimodal mass/age distribution of PN progenitors in the LMC.
|
[
"Moe & De Marco 2006"
] |
[
"For example, the enhanced close binary fraction of metal-poor solar-type stars substantially strengthens the conclusion that the shaping of PN morphologies is the result of binary interactions"
] |
[
"Similarities"
] |
[
[
981,
1000
]
] |
[
[
787,
979
]
] |
2020MNRAS.492.2510L__Galsgaard_et_al._2007_Instance_1
|
We suggest that the jet–facula collision does not cause a change in the field-line connectivity and only leads to the redistribution of jet material. There are two main reasons supporting this viewpoint. The first is the 3D magnetic configuration at the collision region. The penumbra and facula are both located at negative-polarity magnetic fields and the relative orientation of the two flux systems is nearly parallel (Fig. 5). When the magnetic field lines are parallel it is not easy to have reconnection (Feynman & Martin 1995; Galsgaard et al. 2007). As suggested by Gopalswamy et al. (2009), there are two possibilities during the CME–CH interaction. When the magnetic fields in the CME and the CH are parallel, deflection of the CME takes place. When the CME field lines are antiparallel to the CH field lines, magnetic reconnection is possible. In our event, the jet–facula collision is similar to the first situation of CME–CH interaction. The second reason is the absence of reconnection signatures. The jet–facula collision does not produce any impulsive brightenings in the collision region in the H α and EUV images. Although RFJ was heated after the collision, the heating of RFJ is less intense and concentrated than the reconnection heating, and is probably caused by collision-induced conversion of kinetic energy to thermal energy. Due to the interaction of the primary FJs with the facula structure, the material of primary FJs is redistributed, with partial material propagating along the QSL fan plane and forming the AFJs. It is the apparent material end that differentiates the primary FJs and AFJs observationally. In the aspect of magnetic fields, the primary FJs and AFJs are along the same magnetic field lines from the sunspot. Direct observations of the jet–facula collision process and apparent fan-shaped jets are very rare. Thus it is difficult to give an exact physical explanation based on the present observations. If more related cases could be observed in future, we expect to be able to interpret in depth the physical mechanism of collision-induced apparent jets.
|
[
"Galsgaard et al. 2007"
] |
[
"When the magnetic field lines are parallel it is not easy to have reconnection"
] |
[
"Uses"
] |
[
[
535,
556
]
] |
[
[
432,
510
]
] |
2022ApJ...933..243F__Woosley_&_Bloom_2006b_Instance_2
|
Gamma-ray bursts (GRBs) are among the most powerful gamma-ray sources in the universe. They could be generated from the merger of binary compact objects (BCOs; Duncan & Thompson 1992; Usov 1992; Thompson 1994; Metzger et al. 2011) or the death of massive stars (Woosley 1993; Paczyński 1998; Woosley & Bloom 2006a). The merger of BCOs; a black hole (BH)–a neutron star (NS) or NS–NS, leading to kilonovae (KNe), is correlated with short-duration gamma-ray bursts (sGRBs; T
90
10
10
T
90 is defined as the time during which the cumulative number of collected counts above background rises from 5% to 95%. ≲ 2 s; Li & Paczyński 1998; Rosswog 2005; Metzger et al. 2010; Kasen et al. 2013; Metzger 2017). On the other hand, long-duration gamma-ray bursts (lGRBs; T
90 ≳ 2 s; Kouveliotou et al. 1993) are associated with the core collapse (CC) of dying massive stars (Woosley 1993; Galama et al. 1998) leading to supernovae (SNe; Bloom et al. 1999; Woosley & Bloom 2006b). It is believed that in both scenarios large quantities of materials with a wide range of velocities are ejected. In the framework of CC-SNe (depending on the type of SN association), several materials ejected with sub-relativistic velocities less than β ≲ 0.4
11
11
Hereafter, we adopt natural units c = ℏ = 1. have been reported (see, e.g., Kulkarni et al. 1998; Bloom et al. 1999; Woosley & Bloom 2006b; Valenti et al. 2008; Gal-Yam 2017; Izzo et al. 2019, 2020; Modjaz et al. 2020; Nicholl et al. 2020). Regarding the merger of two NSs, sub-relativistic materials such as the cocoon, the shock breakout, and the dynamical and wind ejecta are launched with velocities in the range 0.03 ≲ β ≲ 0.8
12
12
Some authors have considered the shock breakout material in the sub-, trans-, and ultra-relativistic regimes (see, e.g., Kyutoku et al. 2014; Metzger et al. 2015; Fraija et al. 2019c). (see, e.g., Dessart et al. 2009; Metzger & Fernández 2014; Fernández et al. 2015; Kyutoku et al. 2014; Metzger et al. 2015; Nagakura et al. 2014; Murguia-Berthier et al. 2014; Lazzati et al. 2017, 2018; Goriely et al. 2011; Hotokezaka et al. 2013; Bauswein et al. 2013; Wanajo et al. 2014). While the mass and velocity inferred for the first GRB/KN association
13
13
GRB 170817A/AT 2017gfo. were M
ej ≈ (10−4−10−2)M
⊙ and β ≈ (0.1−0.3), respectively (Coulter et al. 2017; Arcavi et al. 2017; Cowperthwaite et al. 2017; Nicholl et al. 2017; Metzger 2019), the mass and velocity inferred for the first GRB/SN association
14
14
GRB 980425/SN1998bw. was M
ej ≈ 10−5
M
⊙ and β ≈ (0.2–0.3), respectively (Kulkarni et al. 1998).
|
[
"Woosley & Bloom 2006b"
] |
[
"Hereafter, we adopt natural units c = ℏ = 1. have been reported (see, e.g.,"
] |
[
"Uses"
] |
[
[
1357,
1378
]
] |
[
[
1240,
1315
]
] |
2016ApJ...831..200W__Lorén-Aguilar_&_Bate_2015_Instance_1
|
HD 97048 is yet another disk for which the dust emission at (sub-)mm wavelengths shows evidence of axisymmetric ring-like structures, here on spatial scales of around tens of au (Walsh et al. 2014; ALMA Partnership et al. 2015; Andrews et al. 2016; Nomura et al. 2016; Zhang et al. 2016). We predict that this substructure will be clearly evident in images of HD 97048 at higher spatial resolution (≈10–20 au, see Figure 13). There remains much debate in the literature on the origin of such axisymmetric substructure in protoplanetary disks including gaps and dust traps carved by forming planets (see, e.g., Dipierro et al. 2015; Pinilla et al. 2015; Rosotti et al. 2016), a change in dust opacity properties at the positions of snow lines (e.g., Banzatti et al. 2015; Zhang et al. 2015; Guidi et al. 2016; Okuzumi et al. 2016), and toroidal dust traps created by hydrodynamic or magnetohydrodynamic effects (see, e.g., Pinilla et al. 2012b; Lorén-Aguilar & Bate 2015; Ruge et al. 2016). To distinguish between each of the scenarios requires observations of dust emission at multiple and well-separated frequencies to determine the radial dust size and density distribution (and dust opacity index) along with emission from optically thin gas tracers to determine the gas surface density. Planets will create deep gaps in the gas surface density as well as influencing the dust (note that this is dependent on the planet mass, see e.g., Rosotti et al. 2016), toroidal instabilities will create much shallower features in the gas surface density, and opacity changes at snow lines will affect only the dust emission and will have no effect on the gas. We note that the ringed substructure seen here has very recently been confirmed in scattered light images of HD 97048 taken with VLT/SPHERE (Ginski et al. 2016). An initial (and shallow) comparison of the data sets shows remarkable coincidence between the positions of the (sub-)mm peaks and gaps and those seen in scattered light. That such structure is seen in both small (≈μm-sized) dust grains in the disk atmosphere and large (≈mm-sized) dust grains in the disk midplane points toward a (proto)planetary system origin; however, further data, particularly to better constrain the gas structure, are needed for confirmation. Since this paper has been accepted for publication, ALMA Cycle 2 data of HD 97048, for which longer baseline data were available and imaged with a uv clip (>160 kλ), resulted in a beam of 048 × 026 (18°) and resolved the inner dust cavity (40 – 46 au) and the bright dust ring at ≈150 au (van der Plas et al. 2016).
|
[
"Lorén-Aguilar & Bate 2015"
] |
[
"There remains much debate in the literature on the origin of such axisymmetric substructure in protoplanetary disks including",
"and toroidal dust traps created by hydrodynamic or magnetohydrodynamic effects (see, e.g.,"
] |
[
"Motivation",
"Motivation"
] |
[
[
944,
969
]
] |
[
[
426,
551
],
[
831,
921
]
] |
2022AandA...665L...1F__Peck_et_al._2001_Instance_1
|
In our model, we compute the tidal response of the oceans and the solid-Earth to luni-solar semi-diurnal forcing, both combined with mimetic continental drift driven by plate tectonics. We focus on the dependence of dissipation on the Earth’s spin rate. We combine two analytical approaches that describe long-wavelength barotropic tidal flows over shallow spherical and hemispherical shells. The spherical shell describes a global ocean that we assume had existed in the earliest eons of the lifetime of the Earth (Motoyama et al. 2020). The existence of an early ocean is supported by evidence from the analysis of detrital zircon around 4.4 Ga (Wilde et al. 2001), from the interaction between the ocean and continental crust 4 billion years ago (Mojzsis et al. 1996), and from records of the oxygen isotope composition of seawater (Peck et al. 2001; Johnson & Wing 2020). The “globality” of this ocean is justified by the analysis of continental crust growth curves based on geochemical evidence in zircon crystallization ages (Dhuime et al. 2012; Hawkesworth et al. 2020). In compliance with these curves, we consider that a hemispherical oceanic shell has taken over in the most recent times. In our model, the center of this hemispheric continental cap follows the evolution of the paleogeographic center. In doing so, we emphasize on the role of “continentality” in the tidal response, while avoiding the under-sampling of geometric scenarios due to theoretical limitations (Hansen 1982; Tyler 2021) or due to uncertainties in plate tectonic models (Matthews et al. 2016; Daher et al. 2021). To compute this evolution, we adopt the recently developed paleogeographic reconstructions that cover the past billion years (Merdith et al. 2021). A postprocessing of these reconstructions allows us to produce the latitudinal evolution of the center of the continental cap captured in Fig. 1. The tidal frequencies at which oceanic resonances are excited and the amplitudes of these resonances vary with the surface position of the hemispherical ocean (Fig. B.1). Super-continental formations and breakups thus have their mark on the predicted lunar recession rate.
|
[
"Peck et al. 2001"
] |
[
"The existence of an early ocean is supported by",
"and from records of the oxygen isotope composition of seawater"
] |
[
"Background",
"Background"
] |
[
[
836,
852
]
] |
[
[
539,
586
],
[
772,
834
]
] |
2015AandA...574A..62S__Burkepile_et_al._(2004)_Instance_1
|
Quiescent prominences are objects formed by relatively cool plasma in the hot corona. Their cool material occurs mostly in the dipped magnetic field lines. In quiescent prominences, which can persist from several hours to several days, magnetic dips form quasi-vertical structures called threads. Quiescent prominences are often observed as part of magnetic structures composed of three coronal patterns: the prominence itself, surrounded by a low-density cavity and a dense helmet streamer overlying the cavity (Engvold 1989). The disruption of the helmet streamer often signifies the beginning of a coronal mass ejection (CME) that can reflect the three-part structure of the helmet streamer (when observed in the white light): the bright leading shell surrounding a dark cavity in which bright prominence material occurs (Crifo et al. 1983; Hundhausen 1999). As the masses of CMEs are most often inferred from the visible-light observations, the masses given in literature are usually only the masses of the leading shell. However, Burkepile et al. (2004) also saw highly structured material most probably originating from a prominence in 63% of CMEs associated with eruptive prominences at the limb observed in visible light using the broadband filter (λλ 5000−5350 Å). Although such observations are called as white light in literature, it would be better to consider them as observations in visible light, because naturally integral intensities in the visible spectral range are observed including both continuum and absorption lines, while white light is just a continuum caused by Thomson scattering on free electrons without the lines. Thus, the presence of absorption spectral lines in visible light complicates very much the estimation of prominence mass, but even if the influence of absorption lines was eliminated, continuum alone only allows us to estimate the mass of ionised material, thus mass of the prominence would be underestimated. Low (1996) stated that the contribution of an erupting prominence to the total mass of CME is usually one order of magnitude less than that of the shell, but in some cases these contributions can be comparable. Low et al. (2003) proposed theoretically an importance of the prominence mass for the deposit of magnetic energy for driving a CME. Thus, developing methods for the estimation of the total mass of prominences can help to explain the connection between prominences and CMEs and provide an important quantity for more accurate estimation of total mass ejected by a CME.
|
[
"Burkepile et al. (2004)"
] |
[
"However,",
"also saw highly structured material most probably originating from a prominence in 63% of CMEs associated with eruptive prominences at the limb observed in visible light using the broadband filter (λλ 5000−5350 Å).",
"Although such observations are called as white light in literature, it would be better to consider them as observations in visible light, because naturally integral intensities in the visible spectral range are observed including both continuum and absorption lines, while white light is just a continuum caused by Thomson scattering on free electrons without the lines."
] |
[
"Background",
"Background",
"Compare/Contrast"
] |
[
[
1035,
1058
]
] |
[
[
1026,
1034
],
[
1059,
1273
],
[
1274,
1644
]
] |
2016MNRAS.463..696L__Randich_et_al._2006_Instance_1
|
Another possible explanation for the peculiar solar composition is that some of the dust in the pre-solar nebula was radiatively cleansed by luminous hot stars in the solar neighbourhood before the formation of the Sun and its planets. This dust-cleansing scenario is supported by the finding that the open cluster M67 seems to have a chemical composition closer to the solar composition than most solar twins (Önehag et al. 2011, hereafter O11 and Önehag, Gustafsson & Korn 2014, hereafter O14). They suggested that the proto-solar nebula was dust-cleansed by massive stars, similar to what happened for the proto-cluster cloud of M67, while the majority of solar twins in the field would presumably have formed in less massive clusters where no nearby high-mass star (≥15 M⊙) was formed. M67 offers the possibility for studying the solar-type stars in a dense cluster environment. This cluster has about solar metallicity ([Fe/H] in the range −0.04 to +0.03, e.g. Hobbs & Thorburn 1991; Yong, Carney & Teixera de Almeida 2005; Randich et al. 2006; Pasquini et al. 2008). The age of M67 is also comparable with that of the Sun: 3.5–4.8 Gyr (Yadav et al. 2008). Pasquini et al. (2008) listed 10 solar-twin candidates in M67, including M67-1194 which hosts a hot Jupiter with a period of 6.9 d and a minimum mass of 0.34 MJup (Brucalassi et al. 2014). M67-1194 was studied by O11 and re-visited by O14. They found that unlike nearby solar twins, which were systematically analysed by Meléndez et al. (2009) and Ramírez, Meléndez & Asplund (2009), the chemical composition of M67-1194 is more solar-like and therefore suggested that the Sun may have been formed in a similar cluster environment, perhaps even in M67. While this scenario is plausible when considering chemical abundances and ages, the dynamics are problematic: the Sun has an orbit close to the Galactic plane, while M67 is presently some 450 pc above the plane. The probability that the Sun, if formed in M67 with an orbit similar to the present cluster orbit, was scattered or diffused out of the cluster into the present solar orbit, is found to be quite low, if the existence of the outer planetary system is taken into account (Pichardo et al. 2012). Gustafsson et al. (2016) suggested that the high-altitude metal-rich clusters (such as M67) were formed in orbits close to the Galactic plane and later scattered to higher orbits by interaction with giant molecular clouds and spiral arms. Thus, it is possible, though not very probable, that the Sun formed in such a cluster before scattering occurred. Currently, only one solar twin in M67 (M67-1194) was spectroscopically analysed with high precision (∼0.02 dex). It is thus crucial to analyse the chemical composition of additional solar-type stars in M67 of very high precision.
|
[
"Randich et al. 2006"
] |
[
"This cluster has about solar metallicity ([Fe/H] in the range −0.04 to +0.03, e.g."
] |
[
"Motivation"
] |
[
[
1029,
1048
]
] |
[
[
883,
965
]
] |
2016MNRAS.455..449H__Oman_et_al._2015_Instance_1
|
With only six free parameters, the standard Λ cold dark matter (ΛCDM) cosmological model fits no less than 2500 multipoles in the cosmic microwave background (CMB) angular power spectrum (Planck Collaboration XVI 2014), the Hubble diagram of Type Ia supernovae, the large-scale structure matter power spectrum, and even the detailed scale of baryonic acoustic oscillations. It thus provides the current basis for simulations of structure formation, and is extremely successful down to the scale of galaxy clusters and groups. Nevertheless, it still faces numerous challenges on galaxy scales. Among these, the most important ones are the too-big-to-fail problem (Boylan-Kolchin, Bullock & Kaplinghat 2011) and the satellite-plane problem (e.g. Pawlowski, Pflamm-Altenburg & Kroupa 2012; Ibata et al. 2014) for dwarf galaxies, the tightness of the baryonic Tully–Fisher relation (McGaugh 2012; Vogelsberger et al. 2014), or the unexpected diversity of rotation curve shapes at a given mass scale (Oman et al. 2015). The latter problem is actually a subset of a more general problem, i.e. that the shapes of rotation curves indeed do not depend on the Dark Matter (DM) halo mass, contrary to what would be expected in ΛCDM, but rather on the baryonic surface density, as has long been noted (e.g. Zwaan et al. 1995). This makes the problem even worse, since the rotation curve shapes are not only diverse at a given mass scale, but uniform at a given baryonic surface density scale, implying a completely ununderstood fine-tuning of putative feedback mechanisms. On the other hand, this behaviour of rotation curves is an a priori prediction of the formula proposed by Milgrom more than 30 yr ago (Milgrom 1983a,b), relating the total gravitational field to the Newtonian field generated by baryons alone, and which can be interpreted as a modification of Newtonian dynamics on galaxy scales below a characteristic acceleration (Modified Newtonian Dynamics (MOND), for a review see Famaey & McGaugh 2012; Milgrom 2014). With this simple formula, high surface brightness (HSB) galaxies are predicted to have rotation curves that rise steeply before becoming essentially flat, or even falling somewhat to the not-yet-reached asymptotic circular velocity, while low surface brightness (LSB) galaxies are predicted to have rotation curves that rise slowly to the asymptotic velocity. This is precisely what is observed, and was predicted by Milgrom long before LSB galaxies were even known to exist. The formula also predicts the tightness of the baryonic Tully–Fisher relation.
|
[
"Oman et al. 2015"
] |
[
"Nevertheless, it still faces numerous challenges on galaxy scales. Among these, the most important ones are",
"or the unexpected diversity of rotation curve shapes at a given mass scale"
] |
[
"Background",
"Background"
] |
[
[
996,
1012
]
] |
[
[
526,
633
],
[
920,
994
]
] |
2018ApJ...864...49P__Tacconi_et_al._2013_Instance_1
|
Figure 9 summarizes our constraints on the molecular gas mass fraction in the analyzed samples of galaxies. In order to convert CO luminosity to molecular gas mass, we consider two different assumptions for αCO. First, we assume a constant value of αCO = 3.6 M⊙ (K km s−1 pc2)−1 adopted by some previous studies (e.g., Daddi et al. 2010b; Decarli et al. 2014; Walter et al. 2016). We then also consider a metallicity-dependent conversion factor, evaluated by assuming a redshift and stellar mass–metallicity relation (Genzel et al. 2015). Previous studies investigating optically and FIR-selected galaxy samples have estimated the relationship between gas mass fraction, stellar mass, redshift, and SFR offset from the main sequence (e.g., Genzel et al. 2015; Scoville et al. 2017). While the PHIBSS project estimated molecular gas masses by measuring the CO(3–2) line emission (Tacconi et al. 2013; Genzel et al. 2015), Scoville et al. (2016, 2017) used the flux on the Rayleigh–Jeans tail of the dust continuum emission to estimate the total gas masses. We here assume that the samples of galaxies plotted, although not complete to any degree due to their preselection for having a spectroscopic redshift, may be somewhat representative of the star-forming main sequence (Figure 9). Their star formation rates are consistent with scatter around the main sequence and appear to include as many galaxies above and below the main sequence estimated by Speagle et al. (2014). Although our CO detections (both blind and with previous spectroscopic redshifts) are indicative of gas fractions compatible with or above (GN19) expectations for main-sequence galaxies, the individual CO nondetections and the stacked signal appear to be systematically lower than the predicted averages, suggesting lower gas mass fractions than might be expected (Figure 9). We can quantify the apparent deficit in stacked signal relative to expectations for the
sample by calculating expected gas masses for the individual stacked galaxies predicted as a function of their redshift, stellar masses, and star formation rates. The expected sample average molecular gas mass is
adopting the best-fit relation by Genzel et al. (2015) and
according to the relation by Scoville et al. (2017). The constant CO luminosity conversion factor above would imply a ratio between expected and observed stacked CO luminosity of 4.8 ± 2.4 and 6.3 ± 3.1, according to the relations by Genzel et al. (2015) and Scoville et al. (2017), respectively. Applying instead the metallicity-dependent CO conversion factor suggested by Genzel et al. (2015) to individual galaxies would somewhat reduce the tension, implying ratios of 3.0 ± 1.7 and 3.8 ± 2.1, according to the relations by Genzel et al. (2015) and Scoville et al. (2017), respectively. While the constraints for low stellar mass galaxies may be compatible with an evolving CO conversion factor due to low metallicity, this is unlikely to resolve the apparent conflict at the high-mass end and may point to lower-than-expected gas masses.
|
[
"Tacconi et al. 2013"
] |
[
"While the PHIBSS project estimated molecular gas masses by measuring the CO(3–2) line emission",
"Scoville et al. (2016, 2017) used the flux on the Rayleigh–Jeans tail of the dust continuum emission to estimate the total gas masses."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
879,
898
]
] |
[
[
783,
877
],
[
921,
1055
]
] |
2021ApJ...910...86R__Stark_et_al._2017_Instance_1
|
One of the major endeavors of modern observational cosmology is to paint a coherent picture of the history of the universe. To this end, the final frontier remains the identification and characterization of the first sources that appeared in the universe, those which played a significant role in reionizing the intergalactic medium from a neutral state to a fully ionized one over the first billion years (corresponding to redshifts of 6 ≲ z ≲ 12). Extragalactic surveys (of deep fields as well as lensing clusters; Grogin et al. 2011; Koekemoer et al. 2011; Bradley et al. 2012; Ellis et al. 2013; Bradley et al. 2014; Schmidt et al. 2014; Treu et al. 2015; Lotz et al. 2017; Salmon et al. 2018; Coe et al. 2019) with the Hubble Space Telescope (HST) have yielded impressive gains in the number of galaxy candidates at redshifts z = 7–10, with samples reaching over 1000 objects, and revolutionized our understanding of galaxy evolution therein. Complementing these observations, the spectroscopic confirmation (e.g., Finkelstein et al. 2013; Oesch et al. 2015; Zitrin et al. 2015; Roberts-Borsani et al. 2016; Hoag et al. 2017; Stark et al. 2017; Hashimoto et al. 2018) and characterization (e.g., Laporte et al. 2017; Mainali et al. 2018; Endsley et al. 2021) of over a dozen sources has seen impressive advances with ground-based spectroscopy (e.g., probing the rest-frame UV and FIR with Keck/MOSFIRE, VLT/X-Shooter, and ALMA), particularly for the brightest and rarest objects. For the rest-frame optical, however, the Spitzer Space Telescope has, until now, afforded the only realistic means for statistical analyses. However, the Infrared Array Camera’s (IRAC) coarse spatial resolution and the limited depth probed by many surveys make robust and uncontaminated constraints on galaxy properties a challenging feat. Further advances with current facilities are challenging owing to the limited wavelength coverage of HST and the observed faintness of star-forming galaxies as one approaches redshifts of z > 10. The imminent arrival of the James Webb Space Telescope (JWST) has the potential to detect galaxies well beyond the current frontier of z ∼ 12 (e.g., Behroozi et al. 2020) thanks to the unprecedented resolution and sensitivity of its near-IR (NIR) imaging and spectroscopic capabilities, and revolutionize our current understanding of galaxy evolution.
|
[
"Stark et al. 2017"
] |
[
"Complementing these observations, the spectroscopic confirmation",
"of over a dozen sources has seen impressive advances with ground-based spectroscopy (e.g., probing the rest-frame UV and FIR with Keck/MOSFIRE, VLT/X-Shooter, and ALMA), particularly for the brightest and rarest objects."
] |
[
"Background",
"Background"
] |
[
[
1131,
1148
]
] |
[
[
948,
1012
],
[
1264,
1484
]
] |
2018MNRAS.481.3573L__Robotham_et_al._2017_Instance_1
|
Another important test for galaxy formation models, is whether they place the right amount of stellar mass into discs and bulges. Moffett et al. (2016) measured the SMF separating galaxies into different morphological types and also into discs/bulges. With this, they derived the fractional contribution from bulges/discs to the total stellar mass in bins of stellar mass. This is quite a difficult measurement to do in observations as light profile fitting is required, which can be robustly done in very disc- and bulge-dominated galaxies, but when both components contribute similarly, the measurement is less robust (Robotham et al. 2017). In Fig. 11, we compare Shark with these measurements for two measurements of bulge mass.9 The first one is considering all the mass in the central concentration (regardless of whether it was formed due to mergers or disc instabilities; solid line), and the second one assumes that the bulge mass formed through disc instabilities (either through the starburst triggered by the gas being fueled to the centre or the stars that are transferred from the disc to the bulge) is part of the disc (dashed line). The bulge mass formed via mergers include both the stars formed via merger-driven starbursts and stars that were accreted by the bulge as a result of a merger (but that were formed in the disc of the primary and/or in the secondary galaxies). The latter is done as pseudobulges in Moffett et al. (2016) were added up to the disc rather than the bulge, and those are thought to form through secular processes taking place in the discs of galaxies (Kormendy & Kennicutt 2004). The effect of assigning the bulge mass formed via disc instabilities to the disc has the effect of shifting the transition from disc- to bulge-dominated stellar budget to higher stellar masses, much closer to the Moffett et al. (2016) observations. In Shark, we find that the formation of elliptical galaxies (i.e. spheroid dominated, massive galaxies) is dominated by galaxy mergers rather than disc instabilities, as disc instabilities only increase the bulge contribution by ≈10 per cent at $10^{11}\, \rm M_{\odot }$. This is qualitatively similar to the finding in large cosmological hydrodynamical simulations that elliptical galaxies form primarily via galaxy mergers (Wellons et al. 2016; Clauwens et al. 2018; Lagos et al. 2018a).
|
[
"Robotham et al. 2017"
] |
[
"This is quite a difficult measurement to do in observations as light profile fitting is required, which can be robustly done in very disc- and bulge-dominated galaxies, but when both components contribute similarly, the measurement is less robust"
] |
[
"Motivation"
] |
[
[
621,
641
]
] |
[
[
373,
619
]
] |
2019MNRAS.490.2071Y__Riess_et_al._2018_Instance_1
|
Set II: we now focus on the observational constraints on the model parameters after the inclusion of the local measurement of H0 by Riess et al. (2018) with the previous data sets (CMB, Pantheon, and CC) in order to see how the parameters could be improved with the inclusion of this data point. Since for this present UM, the estimation of H0 from CMB alone is compatible with the local estimation of H0 by Riess et al. (2018), thus, we can safely add both the data sets to see whether we could have something interesting. Following this, we perform another couple of tests after the inclusion of R18. The observational results on the model parameters are summarized in Table 4. However, comparing the observational constraints reported in Table 3 (without R18 data) and Table 4 (with R18), one can see that the inclusion of R18 data (Riess et al. 2018) does not seem to improve the constraints on the model parameters. In fact, the estimation of the Hubble constant H0 remains almost similar to what we found in Table 3. In order to be more elaborate in this issue, we have compared the observational constraints on the model parameters before and after the inclusion of R18 to other data sets. In Figs 7 (CMB versus CMB+R18), 8 (CMB+CC versus CMB+CC+R18), 9 (CMB+Pantheon versus CMB+Pantheon+R18), and 10 (CMB+Pantheon+CC versus CMB+Pantheon+CC+R18), we have shown the comparisons which prove our claim. One can further point out that the strong correlation between the parameters μ and H0 as observed in Fig. 5 still remains after the inclusion of R18 [see specifically the (μ, H0) planes in Figs 7–10]. The physical nature of μ does not alter at all. That means the correlation between H0 and μ is still existing after the inclusion of R18 to the previous data sets, such as CMB, Pantheon, and CC. In addition to that since μ ≲ 0.9 according to all the observational data sets, thus, the transition from past decelerating era to current accelerating era occurs to be around z ≲ 0.6, similar to what we have found with previous data sets (Table 3).
|
[
"Riess et al. (2018)"
] |
[
"Set II: we now focus on the observational constraints on the model parameters after the inclusion of the local measurement of H0 by",
"with the previous data sets (CMB, Pantheon, and CC) in order to see how the parameters could be improved with the inclusion of this data point."
] |
[
"Uses",
"Uses"
] |
[
[
133,
152
]
] |
[
[
1,
132
],
[
153,
296
]
] |
2021ApJ...907...47L__Lee_et_al._2019_Instance_1
|
In Figure 8, we also find small differences in the [Na, Al, O/Fe] abundance ratios between the stars in the bright and faint RC groups, although it is not as clear as in the case of [Na, Al, O/H] abundances. In particular, unlike Figure 7, stars in the bRC group are more enhanced in [Na/Fe] but appear to be more depleted in [Al/Fe] and [O/Fe] than those in the fRC group. The mean differences are 0.053 ± 0.021 dex, 0.032 ± 0.018 dex, and 0.071 ± 0.045 dex in [Na/Fe], [Al/Fe], and [O/Fe], respectively, which are marginally significant at p-values of 0.22, 0.18, and 0.23. When the relative fraction of RC stars is taken into account (27%; see Section 4), the difference in [Na/Fe] between the genuine RC stars would correspond to Δ[Na/Fe] ∼ 0.20 dex, which is comparable to that expected from our chemical evolution model for the bulge stars (Δ[Na/Fe] = 0.2 ∼ 0.3 dex; Kim & Lee 2018; Lee et al. 2019).10
10
The previous study by Lee et al. (2019) noted a clear separation of the two groups according to Na abundance among bright RGB stars in the outer bulge. The apparent lack of such a distinct difference between the two groups in this study may be due to a larger uncertainty on abundances of relatively faint sample stars.
The overall chemical patterns, however, are not identical to those observed in typical GCs, where the later-generation stars are more enhanced in [Na, Al/Fe] and more depleted in [O, Mg/Fe] than the first-generation stars at a given metallicity, although the trend of [Na, Al O/Fe] between the two RCs is less clear. Figure 9 shows the comparison of stars in this study with stars in metal-rich GCs ([Fe/H] > −1.0) on the Na–O diagram. The stars used in this study have a different distribution from stars in GCs. Although the bRC group is slightly more enhanced in [Na/Fe] and more depleted in [O/Fe] than the fRC group, the [Na/Fe] variation of RC stars is smaller than that of GC stars. This discrepancy might imply the different chemical evolution between stars in the bulge and typical GCs. We note, however, that even though we employ only metal-rich GC stars for the comparison, the majority of stars are still far more metal-poor ([Fe/H] −0.5) than stars in the bulge. Because the relatively small [Na/Fe] variation is expected from the chemical evolution model for metal-rich bulge stars and the O-depletion is indistinct in some metal-rich GCs, such as NGC 6121 and 47 Tuc (see Kim & Lee 2018; Lee et al. 2019), the direct comparison of bulge stars with similarly metal-rich GCs on the Na–O plane would require further spectroscopic observations for such GCs in the bulge.
|
[
"Lee et al. 2019"
] |
[
"When the relative fraction of RC stars is taken into account (27%; see Section 4), the difference in [Na/Fe] between the genuine RC stars would correspond to Δ[Na/Fe] ∼ 0.20 dex, which is comparable to that expected from our chemical evolution model for the bulge stars (Δ[Na/Fe] = 0.2 ∼ 0.3 dex;"
] |
[
"Similarities"
] |
[
[
889,
904
]
] |
[
[
576,
872
]
] |
2018AandA...616A..11G__Westerhout_1957_Instance_1
|
Gaia DR2 contains unprecedented information about the Galaxy, which should allow us to infer its current structure, its equilibrium state, its evolution, modes of mass growth over time, dark matter distribution (and perhaps nature), to cite a few of the questions of modern Galactic astrophysics. As an example, it has been known for several decades that the Galactic disc contains large-scale non-axisymmetric features, including a central boxy/peanut-shaped bar (Okuda et al. 1977; Maihara et al. 1978; Weiland et al. 1994; Dwek et al. 1995; Binney et al. 1997; Babusiaux & Gilmore 2005; López-Corredoira et al. 2005; Rattenbury et al. 2007; Cao et al. 2013) and its possible in-plane extension (Hammersley et al. 2000; Benjamin et al. 2005; Cabrera-Lavers et al. 2007; Wegg et al. 2015), a warp (Burke 1957; Kerr 1957; Westerhout 1957; Weaver 1974; Djorgovski & Sosin 1989; Evans et al. 1998; Gyuk et al. 1999; Drimmel & Spergel 2001; López-Corredoira et al. 2002; Momany et al. 2006; Robin et al. 2008; Reylé et al. 2009; Amôres et al. 2017), and spiral arms (Georgelin & Georgelin 1976; Taylor & Cordes 1993; Drimmel 2000; Bissantz & Gerhard 2002; Churchwell et al. 2009; Vallée 2014; Reid et al. 2014; Hachisuka et al. 2015; Hou & Han 2015). However, full knowledge of these asymmetric structures, that is, of their spatial extent, pattern speeds, and number (in case of spiral arms) is still lacking. Since asymmetries constitute the driver of the secular evolution in galaxy discs (see e.g. Minchev et al. 2012; Fouvry et al. 2015; Halle et al. 2015; Aumer et al. 2017 and Kormendy 2013, for a review) by redistributing angular momentum between the inner and outer disc and between its baryonic and dark matter content (Debattista & Sellwood 2000; Bournaud & Combes 2002; Athanassoula 2003; Martinez-Valpuesta et al. 2006; Combes 2011), quantifying their characteristics is fundamental for understanding to what extent the Milky Way has “simply” evolved secularly in the last ~9 Gyr (Hammer et al. 2007; Martig et al. 2014), or whether some more complex evolutionary scenarios need to be invoked.
|
[
"Westerhout 1957"
] |
[
"As an example, it has been known for several decades that the Galactic disc contains large-scale non-axisymmetric features, including",
"a warp"
] |
[
"Background",
"Background"
] |
[
[
822,
837
]
] |
[
[
297,
430
],
[
791,
797
]
] |
2020ApJ...891...59R__Feruglio_et_al._2015_Instance_1
|
In order to understand the origin of the termination of the powerful jet in Mrk 231, consider the very low accretion rates in these FRI NLRGs, three to four orders of magnitude less than a quasar (Chiaberge et al. 1999, 2002; Hardcastle et al. 2009). However, there are the occasional FRI broad-line galaxies such as the Seyfert 1 galaxy, 3C 120, which has a very prominent FRI morphology on a 400 kpc scale (Walker et al. 1987). Thus, the low accretion rate of FRI NLRGs is not the full explanation of the discrepancy with Mrk 231, but is likely related. One difference between 3C 120 and Mrk 231 is that Mrk 231 has a low-ionization BAL wind (Lipari et al. 1994; Smith et al. 1995). In low radio states, it has also displayed evidence of a high-ionization X-ray-absorbing wind (Feruglio et al. 2015; Reynolds et al. 2017). There are also extreme amounts of intrinsic optical absorption in the galaxy itself from dusty gas (Lipari et al. 1994; Smith et al. 1995). All three circumstances point to a very dense nuclear environment through which the jet must propagate. Furthermore, in Reynolds et al. (2009), it was argued that the density of the free–free absorbing screen at K1 is consistent with the jet being stopped by the BAL wind at K1. By contrast, the extremely low, undetectable accretion of gas in FRI NLRGs is consistent with a very low-density nuclear environment. The launching of the jet in Mrk 231 does not seem to be stopped by the dense nuclear environment, but its propagation does seem to be thwarted. The case of 3C 120 does not seem to be accommodated by this discussion. It is noted that there is no evidence of either a BAL wind or a high-ionization X-ray-absorbing wind in 3C 120 (Oke & Zimmerman 1979; Ballentyne et al. 2004). Thus, it might be the case that there is not an extremely dense nuclear environment near the source of the jet in 3C 120; hence, jet propagation is not hindered. We cannot rule out the possibility that the BAL wind in Mrk 231 also diminishes the power of the jet-launching mechanism as well as providing a drag on its propagation.
|
[
"Feruglio et al. 2015"
] |
[
"In low radio states, it has also displayed evidence of a high-ionization X-ray-absorbing wind"
] |
[
"Background"
] |
[
[
780,
800
]
] |
[
[
685,
778
]
] |
2021AandA...647A..35B__Laffon_et_al._(2010)_Instance_1
|
When we now compare the photodesorption yields at 541 eV between fluences 5 × 1015 photon cm−2 and at 3 × 1017 photon cm−2 in Fig. 2, the CO2 photodesorption first increases from 2.6 × 10−2 molecule/photon to 7.3 × 10−2 molecule/photon. We also observed this phenomenon for CO photodesorption yield (the data are not shown for more clarity), which increased from 1.9 × 10−2 molecule/photon to 3.5 × 10−2 molecule/photon. Second, the estimated yield for the X-ray photodesorption of CH3OH from pure methanol ice decreased by almost one order of magnitude from 9.0 × 10−3 molecule/photon to 1.3 × 10−3 molecule/photon. This indi- cates that the photodesorption of CH3OH is higher for a lower fluence received by the ice when more intact methanol molecules are present in the ice. This aging process favors the photodesorption of simpler molecules such as CO2 or CO. Laffon et al. (2010) estimated with NEXAFS spectroscopy (at the C K-edge) that X-ray irradiation at 150 eV of pure methanol ice at 20 K leads to a survival rate of 50% for methanol after an absorbed dose of 1.1 MGy. In our fixed-energy experiments, we irradiated pure methanol ice with fluences between 5 × 1015 photon cm−2 and 2 × 1016 photon cm−2. Because we irradiated a volume of 0.1 cm2 × 100 ML, with a meanenergy of ~550 eV, and when we consider a volumic mass of condensed methanol of ~ 0.64 g cm−3 (at 20 K; Luna et al. 2018) and an X-ray absorption cross section of ~ 0.6 Mbarn (Ishii & Hitchcook 1988), the absorbed doses used in our fixed energy experiments change from ~ 2 to ~ 15 MGy, whichis quite similar to the absorbed doses in Laffon et al. (2010). This indicates that we could expect a methanol destruction rate of about 50% for our low-fluence experiments. In similar experiments, when irradiating a H2 O:CH4:NH3 (2:1:1) ice mixture covered by a layer of CO:CH3OH (3:1) with 250–1250 eV X-rays during 120 min with a flux of 7.6 × 1014 photon s−1, higher by almost two order of magnitudes than our experiments, Ciaravella et al. (2020) did not detect a desorption signal on the mass channel 31 (attributed to methanol desorption) and estimated that only ~ 20% of methanol molecules remained intact in the first minutes of the irradiation. The irradiation flux therefore appears to be critical for detecting methanol desorption in X-ray irradiation experiments of methanol-containing ices. A lower X-ray flux appears to favor methanol desorption because the methanol destruction rate is lower. This destruction of methanol molecules could also have a significant effect on the formation and desorption of more complex molecules.
|
[
"Laffon et al. (2010)"
] |
[
"estimated with NEXAFS spectroscopy (at the C K-edge) that X-ray irradiation at 150 eV of pure methanol ice at 20 K leads to a survival rate of 50% for methanol after an absorbed dose of 1.1 MGy."
] |
[
"Uses"
] |
[
[
865,
885
]
] |
[
[
886,
1080
]
] |
2016MNRAS.461.1745E__Goerdt_et_al._2010_Instance_1
|
Given that the CDM paradigm only begins to face significant problems at precisely such scales when complex baryonic physics begins to play an important role, it is natural to inquire whether it is the central culprit behind erroneous theoretical predictions. It was for example realized early on that energy from supernovae may be sufficient for driving gas out of the potential wells of dwarf galaxies, the associated mass deficit resulting in the expansion of the central halo region and the flattening of the density profile. More generally, many hydrodynamical simulations implementing stellar and active galactic nuclei (AGN) baryonic feedback processes in a cosmological context are able to reproduce cores (e.g. Governato et al. 2010, 2012; Macciò et al. 2012b; Martizzi et al. 2012; Di Cintio et al. 2014; Chan et al. 2015). However, the complexity of such simulations obscures the physical mechanisms through which these processes affect the dark matter distribution. These mechanisms normally invoke ‘heating’ of the cold central density cusp through an irreversible process, such as dynamical friction from infalling clumps (El-Zant, Shlosman & Hoffman 2001; El-Zant et al. 2004; Tonini, Lapi & Salucci 2006; Romano-Díaz et al. 2008; Goerdt et al. 2010; Cole, Dehnen & Wilkinson 2011; Del Popolo et al. 2014; Nipoti & Binney 2015). Alternatively, repeated gravitational potential fluctuations induced by stellar winds, supernova explosions and AGN could also dynamically heat the central halo (Read & Gilmore 2005; Mashchenko, Couchman & Wadsley 2006; Mashchenko, Wadsley & Couchman 2008; Peirani, Kay & Silk 2008; Governato et al. 2012; Pontzen & Governato 2012, 2014; Zolotov et al. 2012; Martizzi, Teyssier & Moore 2013; Teyssier et al. 2013; Madau, Shen & Governato 2014; Ogiya & Mori 2014). Although the last mechanism may seem most closely related to the supernovae driven wind outflows discussed above, it is in principle more closely connected to the dynamical friction proposal, in the sense that it involves irreversible stochastic dynamics: one may envisage the potential fluctuations leading to cusp-core transformation as originating from stochastic density variations; the relevant ‘clumps’ would be associated with fluctuation scales, as opposed to physically distinct objects dissipating orbital energy via dynamical friction; nevertheless, the basic physical mechanism through which the energy is transferred to the dark matter is similar. For, as is the case in general with processes involving fluctuation and dissipation, fluctuations in a gravitational system can be approximated as stochastic processes described by power spectra and correlation functions, and they can be accompanied by dissipation in the form of dynamical friction (Chandrasekhar 1943; Nelson & Tremaine 1999).
|
[
"Goerdt et al. 2010"
] |
[
"However, the complexity of such simulations obscures the physical mechanisms through which these processes affect the dark matter distribution. These mechanisms normally invoke ‘heating’ of the cold central density cusp through an irreversible process, such as dynamical friction from infalling clumps"
] |
[
"Background"
] |
[
[
1245,
1263
]
] |
[
[
833,
1134
]
] |
2019ApJ...876L..28D__Lamb_et_al._2018_Instance_2
|
In Figures 1 and 2 we show that the X-ray (1.7 keV5
5
This value corresponds to the geometric mean of the XRT energy band, at which the error of the estimated flux can be reasonably suppressed.
), optical (R-band), and radio (6 GHz) fluxes varied with the time of observation applied to the proper corrections if observed at a distance of 200 Mpc, motivated by the fact that the averaged sensitive range of the Advanced LIGO/Virgo detectors in their full-sensitivity run is about 210 Mpc, for the current samples. Due to the faintness of the SGRB afterglow emission, there are gaps of the data between the previous more distant events and GW170817/GRB 170817A (please note that for the latter we only consider the quick decline phase as the early part is significantly influenced by the beam effect of the off-axis outflow). Therefore we extrapolate the very late (t > 200 day) X-ray and optical afterglow data of GRB 170817A to t ∼ 2 day after the burst and then compare them to other events. The radio to X-ray spectrum of the forward shock afterglow emission of GW170817/GRB 170817A is fν ∝ ν−0.6, which yields a p = 2.2 in the slow-cooling synchrotron radiation scenario (Lamb et al. 2018; Troja et al. 2018). In the jet model, such a p can also reasonably account for the very late flux decline of
(Lamb et al. 2018). The extrapolation function of the forward shock emission of GRB 170817A to early times is thus taken as f ∝ t−2.2. Surprisingly, the forward shock afterglow emission of GW170817/GRB 170817A, the first neutron star merger event detected by Advanced LIGO/Virgo, are among the brightest ones for all SGRBs detected so far. Just a few events have X-ray afterglow emission brighter than that of GRB 170817A, as demonstrated in the right panels of Figure 1. The same conclusion holds for the optical and radio afterglow data, as shown in Figure 2, though these two samples are rather limited. We have also compared the distribution of the isotropic gamma-ray energy Eiso, calculated in the rest-frame energy band of 1–104 keV, for the SGRBs with well-measured spectra, and found no significant difference for the SGRBs with and without “long-lasting” afterglow emission (see Figure 3; where the number of events for the X-ray sample are smaller than that presented in Figure 1 because some bursts lack reliable spectral measurements). GRB170817A and GRB 150101B (Troja et al. 2018), two short events with the weakest detected prompt emission, have “bright” late-time afterglow emission because of their off-axis nature.
|
[
"Lamb et al. 2018"
] |
[
"In the jet model, such a p can also reasonably account for the very late flux decline of"
] |
[
"Uses"
] |
[
[
1312,
1328
]
] |
[
[
1215,
1303
]
] |
2017MNRAS.464.3385W__Tsiganis_et_al._2005_Instance_1
|
Oort Cloud. The conclusions that can be reached from this figure about the formation of the Oort Cloud are well known. For example, T93 showed that the parameter space in which an Oort Cloud forms is quite restricted, and that those Oort Clouds that do form have a narrow range of semimajor axes ∼10 000 au. While this parameter space is inhabited by Uranus and Neptune in the Solar system, which should thus readily supply objects to the Oort Cloud (even accounting for the possibility that these planets may have started closer to the Sun; Tsiganis et al. 2005), it could be that Oort Clouds are relatively rare. Many simulations have confirmed these predictions regarding the ability of planets to implant objects in the Oort Cloud (e.g. Dones et al. 2004) while also showing further subtleties such as the ability of Jupiter and Saturn to place a small fraction of the objects they scatter into the Oort Cloud even if ejection is the predominant outcome in such encounters (Brasser, Duncan & Levison 2008). The time-scale predicted for the scattering process to occur is also borne out in numerical simulations. For example, compare the prediction of Fig. 1 that it should take 0.1–1 Gyr for Uranus and Neptune to implant material in the Oort Cloud with fig. 13 of Dones et al. (2015). The radius at which the Oort Cloud forms in the simulations also agrees with that predicted of ∼10 000 au, with some studies includingthe differentiation between inner and outer Oort Clouds (e.g. Lewis, Quinn & Kaib 2013; Brasser & Schwamb 2015). Inspection of Fig. 1 shows that Oort Clouds could form at smaller orbital radii, but that such an outcome requires both low-mass planets and a large system age; for example, if Neptune and Uranus were each of Earth mass, then the Oort Cloud would be at ∼1000 au, but would take ∼20 Gyr to form. Another way to achieve a small Oort Cloud on a shorter time-scale is to place the planetary system in a dense stellar environment (as discussed in Section 3.2).
|
[
"Tsiganis et al. 2005"
] |
[
"While this parameter space is inhabited by Uranus and Neptune in the Solar system, which should thus readily supply objects to the Oort Cloud (even accounting for the possibility that these planets may have started closer to the Sun;",
"it could be that Oort Clouds are relatively rare."
] |
[
"Compare/Contrast",
"Compare/Contrast"
] |
[
[
542,
562
]
] |
[
[
308,
541
],
[
565,
614
]
] |
2022ApJ...926..206Z__Schwab_et_al._2016_Instance_1
|
The first possible scenario is that all FRBs are generated from magnetars, the population of non-repeating FRBs are from magnetars born in the delay formation channels (Margalit et al. 2019) such as neutron star−neutron star mergers associated with short gamma-ray bursts (Rosswog et al. 2003; Price & Rosswog 2006; Giacomazzo & Perna 2013), neutron star−white dwarf mergers possibly associated with rapidly evolving supernovae (SNe) (Toonen et al. 2018; Zhong & Dai 2020), or white dwarf−white dwarf mergers (Yoon et al.2007; Schwab et al. 2016) possibly associated with SNe Ia, and accretion-induced collapse of white dwarfs (Nomoto & Kondo 1991; Tauris et al. 2013; Schwab et al. 2015); the population of repeating FRBs are from magnetars born in the prompt formation channels such as type I superluminous SNe, long gamma-ray bursts, or core-collapse SNe (Duncan & Thompson 1992; Kouveliotou et al. 1998; Kasen & Bildsten 2010). Nonetheless, it seems to be untrue based on recent observations in host galaxies and offsets of the repeaters FRBs 20180916B (Marcote et al. 2020; Tendulkar et al. 2021) and 20200120E (Bhardwaj et al. 2021a; Kirsten et al. 2021). So it may be that the observed non-repeating and repeating FRBs are respectively from less active and more active magnetars due to different magnetic field strengths. If this is the case, however, it is unclear whether the distribution discrepancies of properties between the one-off events and repeaters can arise from different magnetic field strengths. Moreover, if the one-off events and repeaters are produced by an identical mechanism like that generating FRB 20200428 and its associated X-ray burst, it is conceivable that the more active magnetars yielding repeaters could produce more frequent more violent outbreaks resulting in brighter bursts than the less active magnetars yielding one-off events. This appears to be inconsistent with the statistical results shown in Figure 3, though it is still a complicated issue.
|
[
"Schwab et al. 2016"
] |
[
"The first possible scenario is that all FRBs are generated from magnetars, the population of non-repeating FRBs are from magnetars born in the delay formation channels",
"or white dwarf−white dwarf mergers"
] |
[
"Background",
"Background"
] |
[
[
527,
545
]
] |
[
[
0,
167
],
[
474,
508
]
] |
2018AandA...613A..35K__Leloudas_et_al._2011_Instance_1
|
As shown in Fig. 3, the differences in metallicity between different SN subclasses are not significant. This is in contradiction with what is expected from single-star evolution theory, where metallicity-driven winds are crucial: type Ic SNe, which are the most highly stripped, would show the highest metallicity, followed by type Ib, and finally the H-rich type II SNe. The observations, on the other hand, reveals that this is not the case. Some SNe Ic are even located in the low-metallicity part of the distribution in the current sample. This result strengthens the notion that metallicity may not play an important role in deciding theresulting SN type, in agreement with other works based on SN environments (Anderson et al. 2010, 2015; Leloudas et al. 2011; Galbany et al. 2016a). The environments of broad-lined SNe IcBL are found to be relatively metal poor compared to the normal CCSNe, in agreement with previous studies (Modjaz et al. 2011; Galbany et al. 2016a). However, we note that there are only two such SNe in the current sample. The explosion site of SN 1998bw (the first SN to be associated with a GRB: 980425; Galama et al. 1998; Krühler et al. 2017) in this study shows a lower metallicity of 12 + log(O/H) = 8.30 dex compared to the GRB-less SN 2009bb (Pignata et al. 2011), 12 + log(O/H) = 8.49 dex. Levesque et al. (2010a), using slit spectroscopy of the explosion site, concluded that the high metallicity of SN 2009bb site is consistent with typical GRB-less SNe IcBL and not with GRB hosts. Their metallicity value recalculated on the Marino et al. (2013) N2 scale is 12 + log(O/H) = 8.52 dex. These two different cases illustrate the importance of metallicity in deciding whether an SN IcBL progenitor would also produce GRB or not (Modjaz et al. 2008; Levesque et al. 2010b). Progenitors with higher metallicity are not able to spin fast enough and thus produce high angular momentum essential for GRB jet production, eventually producing a GRB-less SN IcBL (Woosley & Bloom 2006).
|
[
"Leloudas et al. 2011"
] |
[
"This result strengthens the notion that metallicity may not play an important role in deciding theresulting SN type, in agreement with other works based on SN environments"
] |
[
"Similarities"
] |
[
[
745,
765
]
] |
[
[
544,
715
]
] |
2022MNRAS.517.5744G__Caro_et_al._2016_Instance_3
|
The CO photodesorption yield reaches its highest value when this ice is deposited at low temperatures (down to 7 K, the lowest temperature studied experimentally) and decreases gradually at higher deposition temperatures (Öberg et al. 2007; Öberg et al. 2009; Muñoz Caro et al. 2010, 2016; Sie et al. 2022). The explanation for this phenomenon motivated further research. It was found that the columnar structure of CO ice samples, grown at incidence angles larger than 45°, increases the effective ice surface exposed to UV photons and therefore the photodesorption efficiency (González Díaz et al. 2019), but ice surface effects cannot account for the large variations observed in the photodesorption of CO ice samples deposited at different temperatures (Muñoz Caro et al. 2016). Absorption band shifts of CO ice in the UV and IR ranges only occurred at deposition temperatures above 20 K (Lasne et al. 2015; Muñoz Caro et al. 2016), suggesting that CO ice grown at lower temperatures is amorphous below 20 K in our experiments, and therefore, the decreasing photodesorption yield is not related to a transition from amorphous to crystalline ice, instead it might be associated to a different degree of molecular disorder in CO ice samples, depending on their deposition temperature. Photon energy transfer via Wannier-Mott excitons between the first photoexcited molecule in the ice and a molecule on the ice surface capable to desorb was proposed (Chen et al. 2017; McCoustra & Thrower 2018). Molecular disorder seems to enhance this energy transfer between neighbour molecules. The colour temperature variations measured at different deposition temperatures could also be the result of molecular disorder (Carrascosa et al. 2021). Urso et al. (2016), Cazaux et al. (2017), and Carrascosa et al. (2021) did not find significant changes in the desorption behaviour or the colour temperature of pure CO ice during controlled warm-up, which points to a low value of the diffusion in the ice. Finally, Sie et al. (2022) investigated the CO photodesorption yield dependence on ice thickness.
|
[
"Muñoz Caro et al. 2016"
] |
[
"Absorption band shifts of CO ice in the UV and IR ranges only occurred at deposition temperatures above 20 K",
"suggesting that CO ice grown at lower temperatures is amorphous below 20 K in our experiments, and therefore, the decreasing photodesorption yield is not related to a transition from amorphous to crystalline ice, instead it might be associated to a different degree of molecular disorder in CO ice samples, depending on their deposition temperature."
] |
[
"Background",
"Compare/Contrast"
] |
[
[
912,
934
]
] |
[
[
783,
891
],
[
937,
1286
]
] |
2022AandA...658A..35K__Boquien_et_al._2019_Instance_1
|
It is important in our analysis to only keep sources for which the SFR and stellar mass estimates are reliable. To this end, we first excluded sources whose fit had reduced χ2,
χ
red
2
>5
$ \chi ^2_{\rm red}>5 $
. This criterion is based on visual inspection of the SED fits and has been adopted in previous studies (e.g., Masoura et al. 2018; Mountrichas et al. 2019; Buat et al. 2021). Increasing the limit to
χ
red
2
>6
$ \chi ^2_{\rm red}>6 $
adds more sources to our sample, but most of them have bad fits and thus unreliable host galaxy measurements. Reducing the threshold to
χ
red
2
>4
$ \chi ^2_{\rm red}>4 $
would exclude additional sources, the vast majority of which have reliable fits. This criterion eliminates 133 AGN (9% of the sample). For each parameter calculated by the SED fitting process, X-CIGALE estimates two values. One is evaluated from the best-fit model, and one weights all models allowed by the parametric grid, with the best-fit model having the heaviest weight (Boquien et al. 2019). This weight is based on the likelihood, exp (−χ2/2), associated with each model. A large difference between these two values for a specific parameter indicates that the fitting process did not result in a reliable estimation for this parameter (Mountrichas et al. 2021a,c; Buat et al. 2021). Thus, when we compared the stellar masses of type 1 and 2 AGN, we only included in our analysis sources with
1
5
≤
M
∗
,
best
M
∗
,
bayes
≤
5
$ \frac{1}{5}\leq \frac{M_{*,\rm best}}{M_{*, \rm bayes}} \leq 5 $
, where M∗,best and M∗,bayes are the best-fit and Bayesian-fit values of M*, respectively. This method to exclude unreliable estimations has been applied in recent studies (Mountrichas et al. 2021a,c; Buat et al. 2021). Using different values for the boundaries of the criterion (i.e. 0.1–0.33 for the lower limit and 3–10 for the upper limit) does not affect the results of our analysis. This criterion reduces the sample to 944 X-ray AGN (93% of the initial dataset). Of these systems, 729 are type 1 and 215 are type 2 (Table 2). Similarly, for the comparison of the SFR of the two AGN types, we require
1
5
≤
SFR
best
SFR
bayes
≤
5
$ \frac{1}{5}\leq \frac{\mathrm{SFR_{best}}}{\mathrm{SFR_{bayes}}} \leq 5 $
, where SFRbest and SFRbayes are the best and Bayesian values of the SFR, respectively, estimated by X-CIGALE. This reduces our sample to 860 X-ray AGN (85% of the initial dataset). Of these sources, 673 are type 1 and 187 are type 2. We note that throughout our analysis we used the Bayes calculations of X-CIGALE for the various parameters.
|
[
"Boquien et al. 2019"
] |
[
"One is evaluated from the best-fit model, and one weights all models allowed by the parametric grid, with the best-fit model having the heaviest weight"
] |
[
"Uses"
] |
[
[
1030,
1049
]
] |
[
[
877,
1028
]
] |
2015AandA...584A.103S__Baym_et_al._1971b_Instance_1
|
Also plotted in Fig. 2 is the pressure in the outer crust from some popular EoSs that model the complete structure of the NS. The figure is drawn up to nb = 3 × 10-4 fm-3, thus comprising the change from the outer crust to the inner crust in order to allow comparison of the EoSs also in this region (notice, however, that inner crust results are not available for the FRDM). We show in Fig. 2 the EoS from the recent BSk21 Skyrme nuclear effective force (Pearson et al. 2012; Fantina et al. 2013; Potekhin et al. 2013; Goriely et al. 2010) tabulated in (Potekhin et al. 2013). The parameters of this force were fitted (Goriely et al. 2010) to reproduce with high accuracy almost all known nuclear masses, and to various physical conditions including the neutron matter EoS from microscopic calculations. We see in Fig. 2 that after the experimentally constrained region, the BSk21 pressure is similar to the BCPM and FRDM results, with just some displacement around the densities where the composition changes from a nucleus to the next one. In the seminal work of BPS (Baym et al. 1971b) the nuclear masses for the outer crust were provided by an early semi-empirical mass table. The corresponding EoS is seen to display a similar pattern with the BCPM, FRDM, and BSk21 results in Fig. 2. The EoS by Lattimer & Swesty (1991), taken here in its Ska version (Lattimer 2015; LS-Ska), and the EoS by (Shen et al. 1998b,a; Sumiyoshi 2015; Shen-TM1) were computed with, respectively, the Skyrme force Ska and the relativistic mean-field model TM1. In the two cases the calculations of masses are of semiclassical type and A and Z vary in a continuous way. Therefore, these EoSs do not present jumps at the densities associated with a change of nucleus in the crust. Beyond this feature, the influence of shell effects in the EoS is rather moderate because to the extent that the pressure at the densities of interest is largely determined by the electrons, small changes of the atomic number Z compared with its semiclassical estimate modify only marginally the electron density and, consequently, the pressure. The LS-Ska EoS shows good agreement with the previously discussed models, with some departure from them in the transition to the inner crust. The largest discrepancies in Fig. 2 are observed with the Shen-TM1 EoS that in this region predicts a softer crustal pressure with density than the other models.
|
[
"Baym et al. 1971b"
] |
[
"In the seminal work of BPS",
"the nuclear masses for the outer crust were provided by an early semi-empirical mass table.",
"The corresponding EoS is seen to display a similar pattern with the BCPM, FRDM, and BSk21 results in Fig. 2."
] |
[
"Background",
"Background",
"Similarities"
] |
[
[
1073,
1090
]
] |
[
[
1045,
1071
],
[
1092,
1183
],
[
1184,
1292
]
] |
2019ApJ...881...38E__Pinto_et_al._2016_Instance_1
|
We fitted the XMM-Newton spectra of ULX-1 and ULX-2 from observation 0794581201 with an absorbed power-law model, using two tbabs absorption components, one frozen to the Galactic value of NH = 1.84 × 1021 cm−2 and the other allowed to vary. In the case of ULX-1, a power-law model was not sufficient to produce a good fit, with χ2/dof = 131.9/90 and the fit showing significant soft residuals at ∼1 keV. These soft residuals are known to be a common feature in the spectra of ULXs and bright X-ray binaries (e.g., Bauer & Brandt 2004; Carpano et al. 2007; Middleton et al. 2015b), and found in other ULXs to be a combination of emission and absorption features related to powerful outflowing winds (including in NGC 6946 ULX-3; Pinto et al. 2016). We used an additional Gaussian component to empirically fit these soft residuals, which resulted in a very large statistical improvement and provided an acceptable fit (χ2/dof = 84.5/87). Using the best-fitting model for each source, we calculated the absorbed 0.3–10 keV flux of both objects. We do not correct for absorption since extending steep power laws, which are empirical rather than physical models, to low energies is likely to overestimate the true flux of the object beneath the absorption—unabsorbed fluxes using such models can be a factor of 2 or 3 higher than those for more physically motivated models, and the amount of absorption is itself model-dependent, so we make fewer assumptions by just considering the observed, absorbed flux. For comparison, we also fitted both spectra with an absorbed multicolor disk model. Although this was able to produce a statistically acceptable fit, the Gaussian representing the soft residuals in ULX-1 contributes to a physically unreasonable portion of the spectrum, broadening to fit the majority of the soft emission, and the residuals for ULX-2 are even less well characterized than for a power-law fit. Therefore, a steep power law seems to be the best empirical model for both spectra. We present the spectral fit results for these two sources in Table 3 and the best-fitting model plots in Figure 2.
|
[
"Pinto et al. 2016"
] |
[
"These soft residuals are known to be a common feature in the spectra of ULXs and bright X-ray binaries",
"and found in other ULXs to be a combination of emission and absorption features related to powerful outflowing winds (including in NGC 6946 ULX-3;"
] |
[
"Similarities",
"Similarities"
] |
[
[
729,
746
]
] |
[
[
405,
507
],
[
582,
728
]
] |
2019MNRAS.487.5666S__Wolfe_et_al._1986_Instance_1
|
The observations of the characteristic double-horned line profiles (Stewart et al. 2014) are all restricted to very low redshifts (z 0.5). Our understanding of the kinematics of H i at high redshifts (z > 0.5) majorly depends on the study of the QSO absorption spectra where the absorption features are mainly dominated by the Lyman−α systems (Zafar et al. 2013). The systems with the largest column densities, the DLAs, are proposed to be the progenitors of the present-day spiral galaxies (Wolfe et al. 2005). Modelling of DLAs observations suggests that DLAs resemble rotating disc galaxies with circular velocities typically of the order of $100\!-\!200\, {\rm km\, s}^{-1}$ (Wolfe et al. 1986; Kauffmann & Charlot 1994; Klypin et al. 1995; Lanzetta, Wolfe & Turnshek 1995; Wolfe 1995; Jedamzik & Prochaska 1998). A number of theoretical studies indicate that the DLAs can have circular velocities as low as ${\sim } 50 \, {\rm km\, s}^{-1}$ (e.g. Kauffmann 1996) which is supported by a number of numerical simulations (Pontzen et al. 2008; Cen 2012; Bird et al. 2014, 2015). QSO absorption spectra also suggest that the gas inside the DLAs has a velocity dispersion of $\sigma _v \approx 5\!-\!10\, {\rm km\, s}^{-1}$ (Wolfe et al. 2005) which is comparable with the nearby galaxies. Other observations also reveal that the high-redshift galaxies show rotational motion (e.g. see Pettini 2009 and references therein). Based on these considerations we have assumed that across the entire range z ≤ 6 the H i resides in rotating disc galaxies which exhibit a double-horned line profile similar to those seen for local galaxies. The value of the parameters α, hf, σv are found to vary from galaxy to galaxy in the local Universe. The statistics of these parameters and their redshift evolution are largely unknown. In our analysis each simulation corresponds to fixed values of these parameters which are held constant over the entire z range. We have carried out simulations covering the entire range of parameter values in order to estimate how this variation affects the RSD. Considering vcirc which determines the overall width of the line profile, it may be noted that the values are determined by the halo mass distribution which evolves with z.
|
[
"Wolfe et al. 1986"
] |
[
"Modelling of DLAs observations suggests that DLAs resemble rotating disc galaxies with circular velocities typically of the order of $100\\!-\\!200\\, {\\rm km\\, s}^{-1}$"
] |
[
"Background"
] |
[
[
681,
698
]
] |
[
[
513,
679
]
] |
2020AandA...639A..46B__Štverák_et_al._(2009)_Instance_2
|
The linear relationship that we observe between breakpoint energy and core temperature is in line with previous measurements (e.g. McComas et al. 1992; Štverák et al. 2009), for both the halo and strahl. According to Scudder & Olbert (1979), a linear trend in the halo relation also follows under the assumption that binary Coulomb collisions dominate electron dynamics in the solar wind. However, in order to align with available experimental data, Scudder & Olbert (1979) set a scaling factor of Ebp/kBTc = 7, which differs from our scaling factor of Ebp/kBTc = 5.5 ± 0.1. With a scaling factor of Ebp/kBTc = 7, Scudder & Olbert (1979) predict that a transformation of thermal electrons into the suprathermal population occurs as the solar wind flows out from the Sun. Findings by Štverák et al. (2009), on the other hand, show that the (nh + ns)/nc ratio remains roughly constant with heliocentric distance in the slow wind, suggesting a lack of interchange between the thermal and suprathermal populations. However Štverák et al. (2009) observes some variability in the (nh + ns)/nc ratio in the fast wind, which they attribute to either statistical effects due to a lack of samples or a possible “interplay” between thermal and suprathermal electrons. Scudder & Olbert (1979) also predict that the halo Ebp/kBTc ratio remains constant with heliocentric distance, whereas Štverák et al. (2009) find that the halo Ebp/kBTc ratio decreases with heliocentric distance. These findings by Štverák et al. (2009), along with the discrepancy between our calculated ratio of Ebp/kBTc = 5.5 ± 0.1 and the prediction of Ebp/kBTc = 7, suggest that the model of Scudder & Olbert (1979) requires a minor update to either the theory or to the input parameters. The discrepancy, however, may also be indicative of other processes, such as wave-particle scattering (e.g. Gary et al. 1994), that possibly modifies the ratio between breakpoint energy and core temperature while preserving its linear relationship.
|
[
"Štverák et al. (2009)"
] |
[
"Findings by",
", on the other hand, show that the (nh + ns)/nc ratio remains roughly constant with heliocentric distance in the slow wind, suggesting a lack of interchange between the thermal and suprathermal populations."
] |
[
"Differences",
"Differences"
] |
[
[
783,
804
]
] |
[
[
771,
782
],
[
804,
1010
]
] |
2017ApJ...834..178Y__Tachihara_et_al._2007_Instance_2
|
In order to investigate the gas kinematics at an early evolutionary stage and the formation of Keplerian disks, we conduct ALMA observations toward three candidate young protostars, Lupus 3 MMS, IRAS 15398−3559, and IRAS 16253−2429. They are selected from our SMA sample (Yen et al. 2015a). These three protostars all have relatively low protostellar masses (0.1 M⊙), inferred from the infalling motions in their protostellar envelopes, and they do not show clear signs of a spin-up rotation on a 1000 au scale; i.e., no signatures of Keplerian disks are seen in our SMA observations (Yen et al. 2015a). Lupus 3 MMS is a Class 0 protostar with a bolometric luminosity (Lbol) of 0.41 L⊙ and a bolometric temperature (Tbol) of 39 K in the Lupus 3 cloud at a distance of 200 pc (Tachihara et al. 2007; Comerón 2008; Dunham et al. 2013). Our SMA results suggest that the protostellar mass in Lupus 3 MMS can be as low as 0.1 M⊙ (Yen et al. 2015a). IRAS 15398−3559 is a Class 0/I protostar with an Lbol of 1.2 L⊙ and a Tbol of 61 K in the Lupus 1 cloud at a distance of 150 pc (Froebrich 2005; Comerón 2008). Early single-dish observations of its CO outflow suggest that IRAS 15398−3559 is close to face on (van Kempen et al. 2009). Recent SMA and ALMA observations show that it is actually closer to edge on (Oya et al. 2014; Bjerkeli et al. 2016). With this new estimated inclination angle (∼70°), our SMA data suggest a low protostellar mass (0.1 M⊙) and a low specific angular momentum in the protostellar envelope (∼1 × 10−4 km s−1 pc; Yen et al. 2015a). IRAS 16253−2429 is a Class 0 protostar with an Lbol of 0.24 L⊙ and a Tbol of 36 K in the ρ Ophiuchus star-forming region at a distance of 125 pc (Dunham et al. 2013). Both CARMA and our SMA results suggest that its protostellar mass is 0.1 M⊙ (Tobin et al. 2012a; Yen et al. 2015a). These three protostars are all embedded in dense cores with masses ≳0.5 M⊙ (Froebrich 2005; Tachihara et al. 2007; Enoch et al. 2009). Therefore, they are excellent targets by which to study the gas motions on a 100 au scale at an early evolutionary stage.
|
[
"Tachihara et al. 2007"
] |
[
"These three protostars are all embedded in dense cores with masses ≳0.5 M⊙",
"Therefore, they are excellent targets by which to study the gas motions on a 100 au scale at an early evolutionary stage."
] |
[
"Motivation",
"Motivation"
] |
[
[
1930,
1951
]
] |
[
[
1838,
1912
],
[
1973,
2094
]
] |
2017AandA...601A.130R__Hernández_et_al._2013_Instance_1
|
Various strategies have been devised to identify modes. One can, for instance, search for frequency patterns appropriate for rapid rotation. The background for this search is the discovery of asymptotically uniform frequency spacings in the numerically computed spectra of uniformly rotating polytropic models (Lignières et al. 2006; Reese et al. 2008) and differentially rotating realistic self-consistent field (SCF) models (Reese et al. 2009a). These uniform spacings have also been modelled through asymptotic semi-analytical formulas (Pasek et al. 2012). In observed spectra, recurrent frequency spacings that may correspond to the large separation or half its value have been found in some stars (García Hernández et al. 2009; García Hernández et al. 2013; Paparó et al. 2016). Moreover, García Hernández et al. (2015) show that mean density estimates based on this type of spacings (obtained via a scaling relation similar to the one in Reese et al. 2008, but based on SCF models) are compatible with independent mass and radii measurements obtained for δ Scuti stars in binary systems. Nonetheless, it is expected that various effects may contribute to hide these regular frequency patterns. First, as mentioned before, the full spectrum is a superposition of sub-spectra corresponding to different classes of modes and some of the uniform spacings only concern one class. This complicates their detection in the full spectrum. Also, owing to their asymptotic nature, these spacings might not be relevant to analyse the low to moderate (up to radial order n ~ 10) frequency domain, typical of most rapidly rotating pulsators. A third effect that may come into play is the presence of mixed modes in evolved stars and/or sharp sound speed gradients since they can potentially modify the regular spacings. Finally, mode selection effects that are due to the non-linearly determined intrinsic mode amplitudes could affect the detectability of the regular patterns. As a first attempt, Reese et al. (2009b) developed a strategy to find these frequency spacings but ran into difficulties when including chaotic modes, which come from another class of modes. Lignières et al. (2010) addressed the same question with encouraging results but their analysis was restricted to the asymptotic regime and relied on simplifying assumptions regarding the spectrum of chaotic modes and the mode visibilities. In this paper, our first goal is to search for regular frequency spacings in the most realistic synthetic spectra available, using relevant frequency ranges and accurate visibility calculations. While they can provide guidance to a similar search in real data, we already know that these results must be taken with caution since the intrinsic mode amplitudes used in this paper are not realistic, but based on ad-hoc prescriptions.
|
[
"García Hernández et al. 2013"
] |
[
"In observed spectra, recurrent frequency spacings that may correspond to the large separation or half its value have been found in some stars"
] |
[
"Background"
] |
[
[
733,
761
]
] |
[
[
560,
701
]
] |
2018MNRAS.473.3810Y__Mitrushchenkov_et_al._2017_Instance_2
|
The lack of data on inelastic processes due to collisions with neutral hydrogen atoms has been a major limitation on modelling of F-, G- and K-star spectra in statistical equilibrium, and thus to reliably proceeding beyond the assumption of local thermodynamic equilibrium (LTE) in analysis of stellar spectra and the determination of elemental abundances. This problem has been well documented, e.g. see Lambert (1993); Barklem (2016a) and references therein. Significant progress has been made in recent times through detailed full-quantum scattering calculations, based on quantum chemical data, for the cases of simple atoms such as Li, Na, Mg and Ca (Belyaev & Barklem 2003; Barklem, Belyaev & Asplund 2003; Belyaev et al. 2010; Barklem et al. 2010; Belyaev et al. 2012; Barklem et al. 2012; Mitrushchenkov et al. 2017). These calculations have demonstrated the importance of the ionic-covalent curve crossing mechanism leading naturally to charge transfer processes (mutual neutralization and ion-pair production), in addition to excitation and de-excitation processes. The importance of this mechanism has allowed various simplified model approaches to be developed, which may be used in cases where suitable quantum chemistry data are not been available. In particular a semi-empirical model has been employed for Al, Si, Be and Ca (Belyaev 2013a,b; Belyaev, Yakovleva & Barklem 2014b; Yakovleva, Voronov & Belyaev 2016; Belyaev et al. 2016), and a theoretical model based on a two-electron asymptotic linear combinations of atomic orbitals (LCAO) approach, has also been employed for Ca (Barklem 2016b, 2017). Comparisons of the two methods show quite good agreement and reasonable agreement with the full quantum calculations is found, particularly for the most important processes with the largest rates (Barklem 2016b, 2017; Mashonkina, Sitnova & Belyaev 2017; Mitrushchenkov et al. 2017). Thus, the model approaches provide a useful route for obtaining estimates of the rates for these processes for many elements of astrophysical interest.
|
[
"Mitrushchenkov et al. 2017"
] |
[
"Comparisons of the two methods show quite good agreement and reasonable agreement with the full quantum calculations is found, particularly for the most important processes with the largest rates"
] |
[
"Similarities"
] |
[
[
1873,
1899
]
] |
[
[
1619,
1814
]
] |
2018ApJ...863..162M__Liu_et_al._2013_Instance_2
|
NLFFF extrapolation provides the reconstructed coronal magnetic field for AR 11158 from 2011 February 13 − 2011 February 15 (Figures 1(d)–(f)). The field lines (yellow lines) within the core of the AR have arcade-like structure with a relatively strong twist mainly near the PIL. These figures show that the magnetic field evolved during this period. Although we did not quantitatively compare the field lines with the observation, in general, the reconstructed coronal field morphologies match with the observations in Figures 1(a)–(c). The general morphologies and the locations of the high-twist fields are also in agreement with many previous studies (Jing et al. 2012; Sun et al. 2012; Dalmasse et al. 2013; Inoue et al. 2013, 2014a; Liu et al. 2013; Wang et al. 2013; Aschwanden et al. 2014; Malanushenko et al. 2014; Zhao et al. 2014). Unlike Zhao et al. (2014) who could identify the twisted flux rope from the topology of the reconstructed coronal field, we could not find an obvious topological signature of a flux rope existing in our NLFFF during our analysis time window. It might be due to the fact that there was little magnetic flux with twist higher than one turn in our NLFFF and it is difficult to topologically define it as a flux rope. However, our result is consistent with other NLFFF results (Jing et al. 2012; Sun et al. 2012; Liu et al. 2013; Wang et al. 2013; Inoue et al. 2014a; Malanushenko et al. 2014). The high-twist region in our result is also in agreement with the region with high helicity flux (Dalmasse et al. 2013) and the location of the flare ribbons (Bamba et al. 2013; Liu et al. 2013), as well as the high current density region (Janvier et al. 2014). Figures 1(g)–(i) show the evolution of the twist distribution map, with the magnetic twist of the field lines plotted at the footpoints of field lines according to a color scale. This shows that the high-twist (strongly right-handed twist corresponding to Tw > 0.5) areas are concentrated in only a limited part of the AR. The high-twist area grew and became even more twisted just before the X2.2 flare (Figure 1(i)). Most parts of the AR have twist values less than 0.25, but near the PIL the twist can reach more than 0.5, even up to about a full turn. This is consistent with the results of Sun et al. (2012) and Inoue et al. (2014a). A high-twist (strong negative/left-handed twist) area also developed in the eastern part of the AR, which did not exist initially on February 13. Both of these high-twist areas produced several flares. However, here we focus on the flares that resulted from the high-twist core region near the center of the AR, where the M6.6 and X2.2 flares occurred.
|
[
"Liu et al. 2013"
] |
[
"However, our result is consistent with other NLFFF results"
] |
[
"Similarities"
] |
[
[
1352,
1367
]
] |
[
[
1257,
1315
]
] |
2015AandA...582A..42K__Cardelli_et_al._(1989)_Instance_1
|
Spectral energy distributions in accordance with the physical parameters of the best-fit spectra were generated with FASTWIND, to provide the flux density per surface unit through our studied bands. The composite flux measured at Earth from a binary at a distance d, at a wavelength λ, reddened to extinction A(λ), is given by \begin{eqnarray*} f_{\lambda}=\dfrac{1}{d^2}\left(R_{1}^{2}F_{1,\lambda}+R_{2}^{2}F_{2,\lambda}\right) \times 10^{-0.4A(\lambda)} \end{eqnarray*}fλ=1d2(R12F1,λ+R22F2,λ)×10−0.4A(λ)where R1,2 and F1,2 are the radii and the surface fluxes of the components, respectively. The composite SED was reddened according to the new family of optical and near-infrared extinction laws for O-type stars provided by Maíz Apellániz et al. (2014), which constitute an improvement of the widely used extinction laws by Cardelli et al. (1989). We ran a fitting algorithm over a wide range of distances with a step of 0.05 kpc, setting free the monochromatic parameters R5495 and E(4405 − 5495) for the type and amount of extinction, respectively, and the best-fit values were considered to be those that minimized the weighted-χ2. To estimate the uncertainty of our measurements, we used a Monte Carlo approach. In particular, we ran the fitting procedure 1000 times using sets of randomly selected parameters (photometry and radii) within their uncertainties assuming a Gaussian distribution is used. The corresponding values of distance are shown in Table 6 for every set of radii and temperatures determined from the four different fit models. Our adopted radii measured with ELC yielded d = 3.52 ± 0.08 kpc, E(4405 − 5495) = 3.66 ± 0.06 mag, R5495 = 3.26 ± 0.04 and A5495 = 11.9 ± 0.1 mag, based on a photometric Teff2 ~ 36 kK. Assuming a spectroscopic Teff2 = 37 kK, the distance changed slightly to d = 3.55 ± 0.08 kpc. Having three degrees of freedom, the reduced-χ2 of our resulting fits was measured to be \hbox{$\chi^2_\textrm{red}\sim11$}χred2~11.
|
[
"Cardelli et al. (1989)"
] |
[
"The composite SED was reddened according to the new family of optical and near-infrared extinction laws for O-type stars provided by Maíz Apellániz et al. (2014), which constitute an improvement of the widely used extinction laws by"
] |
[
"Uses"
] |
[
[
829,
851
]
] |
[
[
596,
828
]
] |
2022AandA...661A.129S__Rodríguez-Almeida_et_al._2021_Instance_3
|
Radio astronomy is recognized as one of the most effective techniques to search for interstellar molecules. By comparing the spectra of candidate molecules in the laboratory with the spectra observed in astronomical surveys, we can determine whether these molecules exist in interstellar space. Therefore, it is necessary to provide rotational spectra of candidates for astronomical detection. Radio astronomy has helped to detect several sulfur-containing molecules in the ISM in recent years: in particular, thiols, the sulfur analogs of alcohols. Methanethiol (or methyl mercaptan, CH3SH) was detected in the Sagittarius B2 (Sgr B2) region of the center of our Galaxy (Linke et al. 1979; Gibb et al. 2000; Müller et al. 2016; Rodríguez-Almeida et al. 2021) and in the protostar IRAS 16293-2422 (Majumdar et al. 2016). Two groups reported to have detected several signs of ethanethiol (C2H5SH) in Sgr B2 (Müller et al. 2016) and Orion (Kolesniková et al. 2014). Most recently, Rodriguez-Almeida reported the first unambiguous detection of ethanethiol in the ISM, toward the G+0.603-0.027 molecular cloud (Rodríguez-Almeida et al. 2021). Moreover, several sulfur-containing species have been observed in comets (Altwegg et al. 2017). Some recent efforts, both from spectroscopy and astronomical searches, to detect S-stitutes of other classes of compounds have also been reported. For instance, thioformic acid (HC(O)SH) was very recently detected in G+0.693–0.027. Its trans-isomer has an abundance of ~1 × 10–10 (Rodríguez-Almeida et al. 2021). Conversely, thioformamide (NH2CHS), the counterpart of for-mamide (NH2CHO), was characterized in the laboratory up to 660 GHz, and its transitions were searched for toward the hot cores Sgr B2(N1S) and Sgr B2(N2), but it was not detected (Motiyenko et al. 2020). The rotational spectrum of thioac-etamide was recently analyzed in the 59.6–110.0 GHz frequency region (5.03–2.72 mm). Its emission was searched for in regions associated with star formation using the IRAM 30 m ASAI observations toward the prestellar core L1544 and the outflow shock L1157–B1. The molecule was not detected, but the study allowed placing constraints on the thioacetamide abundances (Maris et al. 2019; Remijan et al. 2022).
|
[
"Rodríguez-Almeida et al. 2021"
] |
[
"For instance, thioformic acid (HC(O)SH) was very recently detected in G+0.693–0.027. Its trans-isomer has an abundance of ~1 × 10–10"
] |
[
"Background"
] |
[
[
1516,
1545
]
] |
[
[
1382,
1514
]
] |
2015MNRAS.454.1468K__Winckel_2003_Instance_2
|
Owing to their dusty circumstellar environments, a large mid-infrared (mid-IR) excess is a characteristic feature of post-AGB stars and a detection of cold circumstellar material using mid-IR photometry can be used to identify these objects. The first extensive search for these objects was initiated in the mid-80's using results from the Infrared Astronomical Satellite (Neugebauer et al. 1984) which enabled the identification of post-AGB stars in our Galaxy (Kwok 1993). The Toru$\acute{\rm n}$ catalogue (Szczerba et al. 2007) for Galactic post-AGB stars lists around 391 very likely post-AGB objects. The Galactic sample of post-AGB stars have been found to be a very diverse group of objects (Van Winckel 2003). Studies showed that the majority of the optically visible Galactic post-AGB stars could be classified based on their spectral energy distributions (SEDs) into two groups: shell-sources and disc-sources (Van Winckel 2003). The shell-sources show a double-peaked SED with the hot central star peaking at shorter wavelengths while the cold, detached, expanding dust shell peaks at longer wavelengths. This type of SED is considered to be characteristic of objects that follow the single-star evolution scenario mentioned above. The disc-sources do not show two distinct flux peaks in the mid-IR but they do display a clear near-infrared (near-IR) excess indicating that circumstellar dust must be close to the central star, near sublimation temperature. It is now well established that this feature in the SED indicates the presence of a stable compact circumbinary disc, and therefore these sources are referred to as disc-sources (de Ruyter et al. 2006; Deroo et al. 2007; Gielen et al. 2011a; Hillen et al. 2013). The rotation of the disc was resolved with the ALMA array (Bujarrabal et al. 2013a) in one object and using single dish observations Bujarrabal et al. (2013b) confirmed that disc rotation is indeed widespread. Moreover, these disc-sources are confirmed to be binaries and show orbital periods between 100 and 2000 d (Van Winckel et al. 2009; Gorlova et al. 2014). In contrast, for the Galactic shell-sources long-term radial velocity monitoring efforts have not yet resulted in any clear detected binary orbit (Hrivnak et al. 2011), which either confirms the single-star nature of these objects or introduces a possibility that these systems can have companions on very wide orbits.
|
[
"Van Winckel 2003"
] |
[
"Studies showed that the majority of the optically visible Galactic post-AGB stars could be classified based on their spectral energy distributions (SEDs) into two groups: shell-sources and disc-sources"
] |
[
"Background"
] |
[
[
922,
938
]
] |
[
[
719,
920
]
] |
2019MNRAS.487...24G__Rogers_2015_Instance_2
|
NASA’s Kepler mission has unveiled a wealth of new planetary systems (e.g. Borucki et al. 2010). These systems offer new insights into the process of planet formation and evolution. One of Kepler’s key findings is that the most common planets in our Galaxy, observed to date, are between 1 and 4 R⊕, i.e. larger than Earth but smaller than Neptune (Fressin et al. 2013; Petigura, Marcy & Howard 2013). Further observations revealed a transition in average densities at planet sizes ∼1.5 R⊕ (Marcy et al. 2014; Rogers 2015), with smaller planets having densities consistent with rocky compositions while larger planets having lower densities indicating significant H/He envelopes. In addition, Owen & Wu (2013) noticed a bimodal distribution of observed planet radii. Since then, refined measurements have provided strong observational evidence for the sparseness of short-period planets in the size range of ∼1.5–2.0 R⊕ relative to the smaller and larger planets, yielding a valley in the small exoplanet radius distribution (e.g. Fulton et al. 2017; Fulton & Petigura 2018). For example, the California-Kepler Survey reported measurements from a large sample of 2025 planets, detecting a factor of ∼2 deficit in the relative occurrence of planets with sizes ∼1.5–2.0 R⊕ (Fulton et al. 2017). Studies suggest that this valley likely marks the transition from the smaller rocky planets: ‘super-Earths’, to planets with significant H/He envelopes typically containing a few per cent of the planet’s total mass: ‘sub-Neptunes’ (e.g. Lopez & Fortney 2013, 2014; Owen & Wu 2013; Rogers 2015; Ginzburg, Schlichting & Sari 2016). Furthermore, the location of this valley is observed to decrease to smaller planet radii, Rp, with increasing orbital period, P. In a recent study involving asteroseismology-based high precision stellar parameter measurements for a sample of 117 planets, a slope $\text{d log} R_\mathrm{ p}/ \text{d log} P = -0.09^{+0.02}_{-0.04}$ was reported for the radius valley by Van Eylen et al. (2018). A similar value for the slope of $-0.11^{+0.03}_{-0.03}$ was reported by Martinez et al. (2019).
|
[
"Rogers 2015"
] |
[
"Studies suggest that this valley likely marks the transition from the smaller rocky planets: ‘super-Earths’, to planets with significant H/He envelopes typically containing a few per cent of the planet’s total mass: ‘sub-Neptunes’ (e.g."
] |
[
"Background"
] |
[
[
1574,
1585
]
] |
[
[
1293,
1529
]
] |
2022ApJ...936..102A__Williams_et_al._2006_Instance_1
|
Following the original BGK kinetic-theoretical scheme (Bernstein et al. 1957), we transform the above equations to the energy frame, defined as
5
w=12v2+ϕ.
In the energy frame, the ion distribution transforms into
6
fi(w)=Γ(κi)πκi−3/2Γ(κi−1/2)1+2wκi−3/2−κi,
where
7
f(x,v)dv=f(w)dw/2w−ϕ.
Here, the total energy (w) is normalized with
mvth,i2
, i.e., 2k
B
T
i
. As the ions encounter a negative potential well (a pulse), depending on their respective velocities, some of them will become trapped and some of them will pass through. Hence two types of population exist: a trapped population, and a passing population. We assume the passing populations to follow the initial distribution function - a kappa distribution function. In addition, we also assume the form of the potential in which particles are trapped to be prescribed. Spacecraft observations show that wave potential structures of Gaussian form are common in space and astrophysical plasmas (Matsumoto et al. 1994; Williams et al. 2006). A negative wave potential well acts as a perturbation capable of trapping ions in it. We assume this potential well to have Gaussian form, given by
8
ϕ(x)=−ψexp−x22δ2,
where ψ (>0) denotes the amplitude, and δ is the width of the perturbation, respectively. More precisely, δ is actually the distance where the potential decreases to 0.6065 times the maximum amplitude of ψ. The FWHM of the perturbation is actually given by Δ = 2.35δ. The net charge density can thus be expressed as
9
d2ϕdx2=1−Hne−ni,p−ni,tr+H,
where n
i,p
is the passing ion density, and
ni,tr
is the trapped ion density. Particles with suitable velocities falling in the potential range
−−ϕ,+−ϕ
will become trapped, and the rest will pass through. Thus, the range of integration for both passing and trapped ion distributions is given by
10
ni,p=∫−∞−−ϕfp(x,v)dv+∫+−ϕ∞fp(x,v)dv.
As the passing ions follow the kappa distribution, we obtain the passing ion density as
11
ni,p=1−2AB−ϕ2F1[κi,1/2,3/2;ϕ/B],
where 2
F
1 is the hypergeometric function of the first kind. Now that we have obtained the passing ion density, we move on to the derivation of the trapped ion density. In terms of distribution function, the trapped ion density is given by
12
ni,tr=∫−−ϕ+−ϕftr(x,v)dv.
We may now derive the trapped ion density by rearranging Equation (9) as
13
ni,tr=1−Hne−ni,p−d2ϕdx2+H.
Substituting from Equation (11) and differentiating the potential in Equation (8) twice, in Equation (13), we obtain the trapped electron density as
14
ni,tr=1−H1−Trϕ(κe−3/2)0.5−κe+2ϕlog−ϕψδ2+ϕδ2−1+2AB−ϕ2F1[κi,1/2,3/2;+ϕ/B]+H.
For convenience, we make a transformation − ϕ = ∣ϕ∣ → ρ, (>0), and we obtain the trapped ion density as
15
ni,tr=1−HρTrκe−32+10.5−κe−2ρlogρψδ2−ρδ2−1+2ABρ2F1[κi,1/2,3/2;−ρ/B]+H.
|
[
"Williams et al. 2006"
] |
[
"Spacecraft observations show that wave potential structures of Gaussian form are common in space and astrophysical plasmas",
"We assume this potential well to have Gaussian form"
] |
[
"Uses",
"Uses"
] |
[
[
997,
1017
]
] |
[
[
850,
972
],
[
1106,
1157
]
] |
2016ApJ...821...74J__Díaz_et_al._2016_Instance_1
|
Recent theoretical work has suggested that the presence, or lack thereof, of long-period giant planets could affect the formation of such systems. Batygin & Laughlin (2015) argued that the migration of Jupiter within our own solar system might have disrupted a massive primordial inner protoplanetary disk that could have formed multiple short-period super-Earths; they predicted that systems like the Kepler short-period multiple systems should typically lack long-period giant planets. A related question is, how common are planetary systems broadly similar in architecture to our solar system, with small close-in planets and more distant giant planets? We can begin to answer these questions in the near future through the combination of searches for short-period super-Earths and data from the long-term RV programs that have been monitoring many bright FGK stars for well over a decade. Super-Earths can be found with either high-precision RV observations or space-based transit searches. Such high-precision RV surveys include those being undertaken currently with HARPS (e.g., Díaz et al. 2016), HARPS-N (M15), APF (Vogt et al. 2014), and CHIRON (Tokovinin et al. 2013), and in the near future with MINERVA (Swift et al. 2015), CARMENES (Quirrenbach et al. 2014), ESPRESSO (Mégevand et al. 2014), and SPIRou (Artigau et al. 2014). The major upcoming space-based transit survey is that of TESS (Ricker et al. 2015). Long-term RV programs include the McDonald Observatory Planet Search (e.g., Endl et al. 2016), the Anglo-Australian Planet Search (e.g., Jones et al. 2010), the Lick-Carnegie Exoplanet Survey (e.g., Rowan et al. 2016), the CORALIE planet search (Marmier et al. 2013), and the planet search at ESO (e.g., Zechmeister et al. 2013). Long-period giant planets will also be found by Gaia, which will produce a huge sample of astrometrically detected planets (Perryman et al. 2014). While most of the Kepler sample is too faint to have been observed previously by long-term RV surveys (e.g., Coughlin et al. 2015), Gaia will be able to astrometrically detect long-period planets around many of these stars. Our own McDonald Observatory Planet Search program now has a baseline of 12–15 years for ∼200 FGKM stars, and a handful of stars also have lower precision observations dating back more than 25 years. HD 219134 is one of these stars, and here we present an analysis of our RV observations of this star, as well as our data on the stellar activity.
|
[
"Díaz et al. 2016"
] |
[
"Super-Earths can be found with either high-precision RV observations or space-based transit searches. Such high-precision RV surveys include those being undertaken currently with HARPS (e.g.,"
] |
[
"Background"
] |
[
[
1085,
1101
]
] |
[
[
893,
1084
]
] |
2021ApJ...921..179L__Hayes_et_al._2019_Instance_1
|
Quasi-periodic pulsations (QPPs) often refer to the quasi-periodic intensity variations during solar/stellar flares (see Zimovets et al. 2021, for a recent review). In many observations, the flare QPPs were found to show a nonstationary property in the time series integrated over the whole Sun/star or over the oscillation region, for instance, each pulsation has an anharmonic and symmetric triangular profile shape (e.g., Kolotkov et al. 2015; Nakariakov et al. 2019). The signature of flare QPPs can be detected in flare light curves across a broad band of the electromagnetic spectrum, i.e., radio/microwave emissions (Ning et al. 2005; Reznikova & Shibasaki 2011; Nakariakov et al. 2018; Yu & Chen 2019), UV/EUV wavelengths (Shen et al. 2018; Hayes et al. 2019; Reeves et al. 2020; Miao et al. 2021), SXR/HXR and γ-ray channels (Nakariakov et al. 2010; Ning 2017; Hayes et al. 2020; Li et al. 2020c), and the Hα (Srivastava et al. 2008; Kashapova et al. 2020; Li et al. 2020b) or Lyα (Van Doorsselaere et al. 2011; Milligan et al. 2017; Li 2021) emissions. The quasi-periods of these QPPs were reported from subseconds to tens of minutes (e.g., Tan et al. 2010; Shen et al. 2013, 2019; Kolotkov et al. 2018; Karlický & Rybák 2020; Clarke et al. 2021). It should be stated that the observed periods are generally related to the specific channels or flare phases (Tian et al. 2016; Dennis et al. 2017; Pugh et al. 2019), suggesting that the various classes of QPPs could be produced by different generation mechanisms (e.g., Kupriyanova et al. 2020). In the literature, the flare-related QPPs were most often explained by magnetohydrodynamic (MHD) waves, more specifically sausage waves, kink waves, and slow waves (Li et al. 2020a; Nakariakov & Kolotkov 2020; Wang et al. 2021), or by a repetitive regime of magnetic reconnection that could be spontaneous (i.e., self-oscillatory process) or triggered owing to external MHD oscillations (Thurgood et al. 2017; Yuan et al. 2019; Clarke et al. 2021). They can also be interpreted in terms of the LRC-circuit oscillation in current-carrying loops (Tan et al. 2016; Li et al. 2020b) or caused by the interaction between supra-arcade downflows and flare loops (Xue et al. 2020; Samanta et al. 2021).
|
[
"Hayes et al. 2019"
] |
[
"The signature of flare QPPs can be detected in flare light curves across a broad band of the electromagnetic spectrum, i.e., radio/microwave emissions"
] |
[
"Background"
] |
[
[
749,
766
]
] |
[
[
472,
622
]
] |
2019AandA...632A.104G__Hirabayashi_et_al._2016_Instance_2
|
Finally, our observations are consistent with the bilobate shape of the nucleus of comet 8P/Tuttle. As noted in Sect. 1, this shape is likely common among comets because it was found for four out of the six comets for which we have spatially resolved images. This is also the case of the trans-Neptunian object 2014 MU69 (Ultima Thule) observed by the New Horizon spacecraft (Stern et al. 2019). This binary configuration has some implications for the formation and evolution of 8P/Tuttle. A contact binary could result from (i) the accretion at low velocity of two primordial objects (Massironi et al. 2015; Davidsson et al. 2016), (ii) the disruption of a monolithic object due to excessive spin-up resulting from non-gravitational forces or YORP5 effect followed by a reaccretion (Boehnhardt 2004; Ćuk 2007; Hirabayashi et al. 2016), or (iii) the catastrophic disruption of a monolithic object by a collision followed by a re-accretion (Jutzi & Benz 2017; Schwartz et al. 2018). On the one hand, with a low thermal inertia compared with NEAs, the YORP effect is low for comets, in particular for NIC, which have an elongated orbit and spend most of their time far from the Sun, and it may not be sufficient to increase the spin rate of the nucleus to the point where centrifugal exceed gravitational forces. On the other hand, comet 8P/Tuttle has been on a very stable orbit for centuries, and it is likely an evolved comet, as suggested by its low activity, so that it could have been much more active in the past. For cometary nuclei, the primary cause for spin-up is torques caused by outgassing, therefore it is possible that 8P/Tuttle formed as a monolithic body and became a contact binary after its injection into the inner Solar System as a result of excessive spin-up resulting from non-gravitational forces. This scenario has been proposed for comet 67P/Churyumov-Gerasimenko by Hirabayashi et al. (2016). Alternatively, if the binary nature of comet 8P/Tuttle is the result of a primordial accretion or a catastrophic collision in the early Solar Sytem, it could have persisted until now. Similar examples are offered by some binary asteroids that can be stable over the age of the Solar System (Chauvineau et al. 1991), or as proposed by Davidsson et al. (2016) for comet 67P/Churyumov-Gerasimenko. For comet 8P/Tuttle, it is however not possible to distinguish the solution of a binary nucleus that formed in the first billion years of our Solar System (e.g., Matonti et al. 2019) from a more recent origin following its injection into the inner Solar System (e.g., Hirabayashi et al. 2016).
|
[
"Hirabayashi et al. (2016)"
] |
[
"For cometary nuclei, the primary cause for spin-up is torques caused by outgassing, therefore it is possible that 8P/Tuttle formed as a monolithic body and became a contact binary after its injection into the inner Solar System as a result of excessive spin-up resulting from non-gravitational forces. This scenario has been proposed for comet 67P/Churyumov-Gerasimenko by"
] |
[
"Compare/Contrast"
] |
[
[
1892,
1917
]
] |
[
[
1519,
1891
]
] |
2020AandA...641A.155V__Gómez-Guijarro_et_al._2019_Instance_2
|
The scenario presented above has been formulated in various flavors to individually explain several of the properties reported here. The main addition of this work, namely the excitation of CO in distant main-sequence and starburst galaxies, fits in the general picture that we sketched. The ensemble of properties and correlations that we reported here can be also used to revisit the definition of what a starburst is. A standard operational classification is based on the distance from the observed empirical M⋆-SFR correlation, the main sequence. This proved to be a useful distinction and an excellent predictor of several trends (e.g., Sargent et al. 2014), but recent results, including our present and previous analysis (Puglisi et al. 2019), show that the demarcation between starburst and main-sequence galaxies is more blurred that we previously considered. We do detect starburst-like behaviors in galaxies on the main sequence (Elbaz et al. 2018), likely linked to the existence of transitional objects (Popping et al. 2017; Barro et al. 2017b; Gómez-Guijarro et al. 2019; Puglisi et al. 2019, and in prep. to limit the references to recent works based on submillimeter observations). Such transition might well imply an imminent increase of the SFR, driving the object in the realm of starbursts (e.g., Barro et al. 2017b), or its cessation, bringing the system back onto or even below the main sequence (Gómez-Guijarro et al. 2019; Puglisi et al. 2019), with the CO properties potentially able to distinguish between these two scenarios. Regardless of these transitional objects, a definition of starburst based on ΣSFR, rather than ΔMS, would naturally better account for the observed molecular gas excitation properties, dust temperatures and opacities, or SFE (see also Elbaz et al. 2011; Rujopakarn et al. 2011; Jiménez-Andrade et al. 2018; Tacconi et al. 2020). As an example, in Fig. 8 we show the mean SLED of the subsample of galaxies with both CO (2 − 1) and CO (5 − 4) coverage, split at its median ΣSFR. While only tentative at this stage, this suggests a trend of increasing CO excitation with ΣSFR, consistently with Fig. 7 and what mentioned above.
|
[
"Gómez-Guijarro et al. 2019"
] |
[
"Such transition might well imply an imminent increase of the SFR, driving the object in the realm of starbursts",
"or its cessation, bringing the system back onto or even below the main sequence",
"with the CO properties potentially able to distinguish between these two scenarios."
] |
[
"Compare/Contrast",
"Compare/Contrast",
"Future Work"
] |
[
[
1419,
1445
]
] |
[
[
1198,
1309
],
[
1338,
1417
],
[
1469,
1552
]
] |
2018ApJ...866L...1S__Pecharromán_et_al._1999_Instance_3
|
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"
] |
[
"The complex dielectric functions of the samples obtained by heating bayerite and boehmite to various temperatures",
"were derived by modeling the reflectance spectra of pellets obtained by pressing powders of these materials under great pressure"
] |
[
"Uses",
"Uses"
] |
[
[
891,
914
]
] |
[
[
776,
889
],
[
916,
1044
]
] |
2017MNRAS.471.4256V__King_et_al._2000_Instance_1
|
Our ideas about which systems may survive spiral-in and produce WR X-ray binaries were triggered by the realization that the peculiar X-ray binary SS433 has avoided going into CE evolution, and that the donor star in this system is transferring mass to the compact star by stable Roche lobe overflow (RLOF; King & Begelman 1999; King, Taam & Begelman 2000). By analysing the properties of this system we realized that it is the high mass of the compact star in this system [4.3(±0.8) M⊙] in combination with the relatively low mass of its donor star [12.3(±3.3) M⊙] (Hillwig & Gies 2008), which allowed it to avoid CE evolution and enables it to gently spiral in without ever coalescing with its donor. As we consider SS433 to be a ‘keystone’ for understanding the formation of the WR X-ray binaries, we give in Section 2 a brief overview of its properties and evolutionary state. The avoidance of its going into CE evolution is – as we will argue – a consequence of the donor star having a radiative envelope (King et al. 2000), in combination with the donor star and accretor having a mass ratio less than 3.5. In Section 3, we then examine for which donor masses, mass ratios and orbital periods HMXBs will, when they start Roche lobe overflow, avoid going into CE evolution and may survive as WR X-ray binaries with short orbital periods. We also examine under which conditions they may still survive after having gone into CE evolution. In this section, some examples are given on how a number of well-known observed WR+O binaries with relatively short orbital periods are expected to evolve in the future, and are expected to produce WR X-ray binaries and, as a final evolutionary state, close double BHs. This model for producing double BHs is different from the ones proposed by Belczynski et al. (2016), Marchant et al. (2016) and de Mink & Mandel (2016). In Section 4, we attempt to estimate the birth rate of WR X-ray binaries in the Galaxy on the basis of our model, and find it to be still higher than observed and discuss possible ways to minimise this discrepancy. In Section 5, we discuss the results and estimate the possible birth rate of double BHs based on our model.
|
[
"King et al. 2000"
] |
[
"The avoidance of its going into CE evolution is – as we will argue – a consequence of the donor star having a radiative envelope"
] |
[
"Uses"
] |
[
[
1012,
1028
]
] |
[
[
882,
1010
]
] |
2016AandA...586A..92P__Maciesiak_et_al._2012_Instance_1
|
We calculated the dependence of the width of the profiles on the pulse period, considering the different frequencies separately. We note that the pulse width is not a direct reflection of the beam size or diameter (i.e. 2ρ, where ρ is beam radius). For a visual representation of the geometry see for instance Maciesiak et al. (2011) and Bilous et al. (2014). In fact, only if the observer’s line of sight cuts the emission centrally for magnetic inclination angles, α, that are not too small (i.e. α> ~ 60°), w ≈ 2ρ. In such a case, when the emission beam is confined by dipolar open field lines, we would expect a P− 1 / 2 dependence, which has indeed been observed when correcting for geometrical effects by transforming the pulse width into a beam radius measurement (see Rankin 1993; Gil & Krawczyk 1996; Maciesiak et al. 2012). For circular beams, profile width and beam radius are related by the relation first derived by Gil et al. (1984): (5)\begin{equation} {\rho_{10}} = 2\sin^{-1} \left[\sin{\alpha}\sin{(\alpha+\beta)}\sin^2\left(\frac{w_{10}}{4}\right)+\sin^2\left(\frac{\beta}{2}\right)\right]^{1/2}\cdot \end{equation}ρ10=2sin-1sinαsin(α+β)sin2w104+sin2β21/2·The angle β is the impact angle, measured at the fiducial phase, φ, which describes the closest approach of our line of sight to the magnetic axis. This equation is derived under the assumption that the beam is symmetric relative to the fiducial phase. Typically, widths are measured at a certain intensity level (e.g. 50% or 10%, as here), and ρ values are derived accordingly. In many cases, profiles are indeed often asymmetric relative to the chosen midpoint, or become so as they evolve with frequency. We note that for a central cut of the beam (β = 0) and for an orthogonal rotator (α ~ 90°) the equation reduces to ρ = w/ 2 as expected, while in a more general case, where β = 0 and α ≫ ρ the relation reduces to ρ = (w/ 2)sinα. In principle, it is possible to determine α and β with polarisation measurements. However, in reality the duty cycle of the pulse is often too small to obtain reliable estimates (see Lorimer & Kramer 2004). Alternatively, at least for α, the relation reported by Rankin (1993) can be used: (6)\begin{equation} \label{eqn:w50} w_{50,{ \rm core}}(1~{\rm GHz}) = 2.45^\circ \cdot P^{-0.5\pm 0.2}/\sin(\alpha), \end{equation}w50,core(1GHz)=2.45◦·P−0.5±0.2/sin(α),calculated from the observed width dependency on period for the core components of pulsars (see Sect. 5.1), which is intrinsically related to the polar cap geometry. Equation (6) is valid at 1 GHz, but can be applied at LOFAR frequencies, maintaining the same dependence, if the impact angle β ≪ ρcore; sinα should be ignored for orthogonal rotators. Additionally, Rankin (1993), Gil et al. (1993), Kramer et al. (1994), Gould & Lyne (1998) suggested that “parallel” ρ − P relations are found if the radio emission of the pulsar can be classified and separated into emission from “inner” and “outer” cones, which seem to show different spectral properties (see Sect. 5.1 for details).
|
[
"Maciesiak et al. 2012"
] |
[
"In such a case, when the emission beam is confined by dipolar open field lines, we would expect a P− 1 / 2 dependence, which has indeed been observed when correcting for geometrical effects by transforming the pulse width into a beam radius measurement (see"
] |
[
"Compare/Contrast"
] |
[
[
811,
832
]
] |
[
[
519,
776
]
] |
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