adsabs/FOCAL huggingface dataset:
SFR and 




L

CO(1-0)
′



$ L^\prime_{\rm CO(1{-}0)} $



 can be derived from observations, but the molecular gas mass, MH2 = αCO × 




L

CO(1-0)
′



$ L^\prime_{\rm CO(1{-}0)} $



, is the key quantity of interest. The observed SFR/




L

CO(1-0)
′



$ L^\prime_{\rm CO(1{-}0)} $



 ratio is related to both the depletion time (τdep) and the CO-to-H2 conversion factor (αCO) as SFR/




L

CO(1-0)
′



$ L^\prime_{\rm CO(1{-}0)} $



 = SFR/(MH2/αCO) = αCO/τdep. Many studies have shown that αCO increases rapidly at low metallicity theoretically and observationally (Israel 1997; Glover & Mac Low 2011; Leroy et al. 2011; Narayanan et al. 2012; Elmegreen et al. 2013; Shi et al. 2016). If τdep is constant, as is found for the nearby disc galaxies (Leroy et al. 2008; Bigiel et al. 2011), then the dependence of the SFR/




L

CO(1-0)
′



$ L^\prime_{\rm CO(1{-}0)} $



 ratio on metallicity is the direct consequence of the variation of αCO. However, τdep in dwarf galaxies remains uncertain and tends to be shorter at a lower mass and at a higher specific SFR (sSFR; Saintonge et al. 2011; Shi et al. 2014; Hunt et al. 2020). IZw18, as a dwarf galaxy with a relatively high sSFR and low mass, is likely to have low τdep. We speculates that as αCO = τdep × 



SFR


L

CO
(
1
−
0
)

′




$ \frac{\mathrm{SFR}}{L^\prime_{\mathrm{CO(1{-}0)}}} $


, the high SFR/




L

CO(1-0)
′



$ L^\prime_{\rm CO(1{-}0)} $



 ratio of IZw18 would overwhelm the possibly low τdep, and this would result in a high αCO. Moreover, considering the dependence of αCO on both sSFR and metallicity for the metal poor galaxies, IZw18 follows the empirical relation found in Hunt et al. (2020) within the uncertainty of the stellar mass, as shown in the bottom right corner of Fig. 5. This relation was derived based on a recent compilation of ∼400 metal poor galaxies (Ginolfi et al. 2020, MAGMA). We note that the individual star-forming regions of the metal-poor galaxies from Shi et al. (2015, 2016) all fall slightly above the relation, but well within the scatter of the global measurements of the galaxies. This indicates that αCO changes continuously with metallicity and sSFR. A further constraint on the conversion factor (αCO) of IZw18 is beyond the scope of this paper.
Citation: Leroy et al. 2008
Citation function: Uses; Uses
The behaviour of the radio emission from the early-type magnetic stars has been definitively quantified by the empirical scaling relationship between the radio spectral luminosity and the magnetic flux rate: $L_{\nu ,\mathrm{rad}} {\propto } (\Phi / P_{\mathrm{rot}})^2 = {B_{\mathrm{p}}^2 R_{\ast }^{4}} / {P_{\mathrm{rot}}^2}$, found by Leto et al. (2021) and confirmed by Shultz et al. (2022). This empirical relationship is the consequence of the physical mechanism supporting the non-thermal acceleration able to produce the electrons responsible for the magnetospheric radio emission of early-type magnetic stars. It was recently demonstrated by Owocki et al. (2022) that the power provided by centrifugal breakout events, continuously occurring within the stellar magnetosphere, is directly related to the power of the radio emission. From this point of view, the CBOs are non-random events. Breakouts occur continuously in a well-constrained magnetospheric region, where the resulting reconnection of the magnetic fields drives the acceleration of the local electrons. The CBO luminosity is related to the stellar parameters by the relation (Owocki et al. 2022): 
$$\begin{eqnarray*}
L_{\mathrm{CBO}} = \frac{B_{\mathrm{p}}^2 R_{\ast }^3}{P_{\mathrm{rot}}} \times W {\mathrm{~~(erg \, s^{-1}),}}
\end{eqnarray*}$$with Bp in gauss, R* in centimetres, and Prot in seconds. In the above relation $W=2 \pi R_{\ast } / P_{\mathrm{rot}} \sqrt{G M_{\ast }/R_{\ast }}$ is the dimensionless critical rotation parameter calculated as the ratio between the equatorial stellar velocity and the corresponding orbital velocity, defined by the gravitational law at the stellar equator (G = 6.67408 × 10−8 cm3 g−1 s−2 is the gravitational constant). Once the explicit relation of W is substituted within the LCBO definition, we find that $L_{\mathrm{CBO}} \propto {B_{\mathrm{p}}^2 R_{\ast }^{4.5} M_{\ast }^{-0.5}} / {P_{\mathrm{rot}}^2} \propto L_{\nu ,\mathrm{rad}}$, which in practice is almost the same relation empirically found, except for the term related to the mass, that in any case has negligible effect. The maximum possible value of W is W = 1, which, due to magnetic braking, would have to be at the beginning of the star’s life; however no magnetic star has ever been found with W⪆0.5 (Shultz et al. 2019b). In fact, this parameter progressively decreases as the star loses angular momentum throughout its life (Keszthelyi et al. 2019, 2020).
Citation: Owocki et al. (2022)
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Citation: Owocki et al. 2022
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The presence of neutral lines in the IRIS spectra allow us an accurate absolute wavelength calibration, and therefore a more accurate determination of line Doppler shifts, as compared to some other spectral observations; e.g., with EIS. For the post-flare loops, we find that the Fe xii emission is largely unshifted at the loop tops and predominantly redshifted at both loop footpoints, as expected for the draining of plasma in the later phases of the flare, when the heating has ceased. For the moss observations, we find that the distribution of Doppler shifts is peaked at a redshift of a few km s−1, but with significant wings on both the red and the blue sides. In our interpretation, guided by the Bifrost 3D models, Fe xii appears redshifted for hotter coronal temperatures (≳4 MK), while blueshifted when the corona is at cooler temperatures. The distribution of Doppler shift we observe shows a wide range of blue- and redshifts, and it suggests that there may be a continuous mix of heating and cooling in the moss we observed. For cooler TR lines (





[K]





), the dominance of redshift is well established (e.g., Doschek et al. 1976; Dere et al. 1984; Achour et al. 1995; Chae et al. 1998b; Peter & Judge 1999). The physical processes causing the observed redshifts are still debated, and several models have been proposed (e.g., Hansteen 1993; Patsourakos & Klimchuk 2006; Peter et al. 2006; Hansteen et al. 2010). Hansteen et al. (2010) use 3D radiative MHD models of the solar atmosphere to investigate the TR redshifts, and find that in their model the heating is dominated by rapid intermittent events at low heights which heat the plasma locally to coronal temperature and produces downflows at TR temperature due to the local overpressure. The model analyzed by Hansteen et al. (2010) reached modest coronal temperatures (≲2.5 MK), therefore with significant Fe xii emission in loop structures rather than at the loop footpoints, as in the case for moss. Here we used a different Bifrost 3D MHD simulation, which reaches higher coronal temperatures and represents a better model for moss observations. We find that the moss Fe xii (mostly) redshifted emission we observe with IRIS is reproduced in the model when the corona gets hot enough and most of the Fe xii is confined in the TR loops. The Fe xii then behaves like a TR line, and is redshifted, in the model, as a consequence of the local heating at low heights which causes upflows in the hotter lines and downflows in the cooler (TR) lines.
Citation: Dere et al. 1984
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After crossing the bow shock, neutral atoms can interact with the shocked protons. The interaction length is given by

(2)

\begin{equation}
\lambda = \frac{V_{\rm NS}}{X_{\rm ion} n_{\rm ISM} r_{\rm c} \langle \sigma (v_{\rm rel}) v_{\rm rel} \rangle },
\end{equation}

where Xion is the ionization fraction of the ISM, rc is the compression ratio of the bow shock, σ(vrel) is the relevant cross-section of the process under consideration and 〈σvrel〉 is the collision rate averaged over the ion distribution function. When the ion distribution is a Maxwellian, 〈σvrel〉 is well approximated (within 20 per cent) by the expression (Zank et al. 1996; Blasi et al. 2012)

(3)

\begin{equation}
\langle \sigma v_{{\rm rel}} \rangle \approx \sigma (U^*) U^*,
\end{equation}

where

(4)

\begin{equation}
U^*= \sqrt{\frac{8}{\pi } \frac{2 k_{\rm B} T}{m_{\rm p}} }
\end{equation}

is the average, relative speed between the incoming hydrogen atom and ions (T is the temperature of the shocked ISM determined by the Rankine–Hugoniot jump conditions, assumed to be ≫TISM). Using the fiducial values nISM = 0.1 cm−3 and VNS = 300 km s−1, together with an ionization fraction of 90 per cent and rc = 4 (the typical value for strong shocks), leads to the following estimates for the mean free paths:

(5)

\begin{eqnarray}
\lambda _{{\rm ion},{\rm p}} &amp; \approx 3.0 \times 10^{20} \, \rm cm,
\end{eqnarray}


(6)

\begin{eqnarray}
\lambda _{{\rm ion},{\rm e}} &amp; \approx 2.2 \times 10^{16} \, \rm cm,
\end{eqnarray}


(7)

\begin{eqnarray}
\lambda _{\rm CE} &amp; \approx 1.5 \times 10^{15} \, \rm cm
\end{eqnarray}

for the ionization due to collisions with protons, electrons and CE, respectively. Note that λion, e has been calculated under the assumption that the electrons downstream of the shock equilibrate rapidly with protons, thereby acquiring the same temperature. If this assumption does not hold, the collisional length scale for ionization due to electrons can become much larger than the value reported in equation (6). In addition, the values (5)–(7) are to be taken as lower limits as they are valid just ahead of the nebula, where the compression ratio and the temperature obtain their maximum values.
Citation: Zank et al. 1996
Citation function: Uses; Uses
The driving mechanism and physical nature of large-scale EUV waves have been hotly debated over the past 20 years (Chen 2017). Because chromospheric Moreton waves were believed to be driven by the flare pressure pulses for a long time after Uchida (1968), solar physicists naturally thought at the beginning that EUV waves also have the same driving mechanism (e.g., Khan & Aurass 2002; Hudson et al. 2003; Warmuth et al. 2004). However, more and more recent high-resolution observational studies have shown that EUV waves are in fact bow shocks driven by CMEs instead of flare pressure pulses (e.g., Chen 2006; Kienreich et al. 2009; Patsourakos & Vourlidas 2009; Patsourakos et al. 2010; Liu et al. 2011b; Shen & Liu 2012a, 2012c; Xue et al. 2013; Long et al. 2017a; Shen et al. 2017a; Cunha-Silva et al. 2018). Further investigations indicated that EUV waves can also be excited by other kinds of slightly different processes, such as newly formed expanding loops, coronal jets, and mini-filament eruptions on small scales (e.g., Podladchikova et al. 2010; Zheng et al. 2012, 2013; Su et al. 2015; Shen et al. 2017a), and the sudden expansion of transequatorial loops driven by coronal jets on large scales (Shen et al. 2018e, 2018d). In these scenarios, those authors found that such EUV waves often have a shorter lifetime (a few minutes) compared to those driven by CMEs (tens of minutes). Moreover, Shen et al. (2018a) reported an interesting non-CME-associated homologous EUV wave event in which the initial speeds of the waves are about 1000 km s−1, and the authors proposed that the observed EUV waves were large-amplitude nonlinear fast-mode magnetosonic waves or shocks directly driven by the associated recurrent coronal jets, resembling the generation of a piston shock in a tube. These new observations suggest that EUV waves can be excited by different kinds of coronal disturbances, e.g., CMEs, coronal jets, and mini-filament eruptions.
Citation: Patsourakos & Vourlidas 2009
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Citation: Shen et al. 2018e
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Citation: Shen et al. (2018a)
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Due to recent major advances in the observational studies of active galactic nuclei (AGN), evidence is growing that massive black holes (MBHs) heavier than $10^5\, \rm {M_\odot }$ form in nature and power AGN activity at the centres of galaxies through gas accretion (Schmidt 1963; Hopkins, Richards & Hernquist 2007; Merloni & Heinz 2008; Ueda et al. 2014; Aird et al. 2015). The demographic study of AGN and the dynamics of stars and gas around the centre of nearby galaxies, further provided evidence that most (if not all) massive galaxies in the Universe host MBHs in their nuclei (Genzel & Townes 1987; Dressler & Richstone 1988; Kormendy 1988; Kormendy & Richstone 1992; Genzel, Hollenbach & Townes 1994; Salucci et al. 1999; Peterson et al. 2004; Vestergaard & Peterson 2006). Even more, the existing correlations between the mass of MBHs and key properties of their host galaxies hint for their co-evolution (Haehnelt & Rees 1993; Faber 1999; O’Dowd, Urry & Scarpa 2002; Häring & Rix 2004; Kormendy & Ho 2013; Savorgnan et al. 2016). Even though these findings sharpened our knowledge on the role of MBHs in the formation and evolution of galaxies, there is a need to contextualize galaxies and MBHs within the broad cosmological context. It is commonly accepted that the Universe behaves in a hierarchical way. Cosmic structure formed through the hierarchical assembly of dark matter (DM) haloes, and the galaxies observed nowadays assembled through mergers with smaller companions and accretion of matter from the cosmic filaments (White & Rees 1978; White & Frenk 1991; Haehnelt & Rees 1993; Kauffmann et al. 1999; Guo et al. 2011; Vogelsberger et al. 2014a,b; Schaye et al. 2015; Nelson et al. 2018; Pillepich et al. 2018). Consequently, the existence of MBHs at the centre of galaxies and the main role of mergers in the Universe, hint for the existence of massive black hole binary systems (MBHBs) that might have formed and coalesced throughout the whole Universe lifetime.
Citation: Kormendy 1988; O’Dowd, Urry & Scarpa 2002
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Citation: Ueda et al. 2014
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Citation: Kauffmann et al. 1999
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The two components in V593 Cen are early-type stars that contain the convective core and the radiative envelope. This suggests that the period oscillation cannot be explained by the magnetic activity cycle mechanism, which is usually adopted to explain the cycle period change of later-type binary stars (e.g., Applegate 1992; Lanza et al. 1998). Therefore, the light-travel-time effect via the presence of a tertiary body is adopted to explain the cyclic change of the orbital period (e.g., Irwin 1952; Chambliss 1992; Borkovits & Hegedüs 1996; Liao & Qian 2010). Using this method, tertiary bodies have been detected orbiting massive close binary stars. Some examples are V701 Sco (Qian et al. 2006), V382 Cyg, TU Mus (Qian et al. 2007), and AI Cru (Zhao et al. 2010). Though it seems that binaries favor a more uniform distribution of eccentricities (Tokovinin 1997, 2008; Raghavan et al. 2010; Moe & Di Stefano 2017), we still assume that the tertiary wanders in the circular orbit. This is because the period change (in Figure 2) displays a periodic oscillation, and this oscillation is fitted either by a sine curve without eccentricities, or by using the value of the outer orbital period and the inner binary period (Pouter and Pinner). The period ratio for this multiple system is about Pouter/Pinner = 2.5 × 104. For this kind of multiple star system with a very large period ratio, the outer orbit usually has very low or no eccentricities (Shatsky 2001; Tokovinin 2004, 2008). Moreover, in large Pouter/Pinner ratio multiple system (commonly found in triple systems) the Kozai effect becomes too weak (Tokovinin 2008). For brevity, considering that the third body in the V593 Cen system is moving in a circular orbit, the projected radius of the orbit of the eclipsing pair rotating around the mass central of the triple system, the mass function, and the masses of the third component could be computed with these equations as follows:
5







6





where A3 is the amplitude of the O − C oscillation, c is the light speed, P3 is the period of the O − C oscillation, and i′ is the inclination of the tertiary orbital plane. Finally, when the inclination i′ = 90°, it is calculated that the lowest mass of the tertiary companion is 4.3 (0.3) M⊙ and the separation between the binary and the tertiary companion is about 25.5(2.2) au. For the other inclinations, the third body should have a larger mass and a shorter separation. When the inclination i′ = 826, which shows that the third body is in the same plane with the inner binary, its mass is calculated to be 4.7 (0.4) M⊙ and the separation is 25.2(2.8) au.
Citation: Zhao et al. 2010
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Citation: Shatsky 2001
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Citation: Applegate 1992
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The binary characteristics of early-type magnetic stars may provide crucial clues, allowing one to test alternative fossil field hypotheses. The non-magnetic chemically peculiar stars of Am (A-type stars with enhanced lines of Fe-peak elements) and HgMn (late-B stars identified by strong lines of Hg and/or Mn) types are frequently found in close binaries (Gerbaldi, Floquet & Hauck 1985; Ryabchikova 1998; Carquillat & Prieur 2007), including eclipsing systems (Nordstrom & Johansen 1994; Strassmeier et al. 2017; Takeda et al. 2019). In contrast, only about ten close (Porb  20 d) spectroscopic binaries containing at least one magnetic ApBp star are known (Landstreet et al. 2017). The overall incidence rate of magnetic upper main sequence stars in close binaries is less than 2 per cent (Alecian et al. 2015), although this fraction is significantly higher if one includes wide long-period systems (Mathys 2017). This low incidence of magnetic ApBp stars in close binaries is frequently considered as an argument in favour of the stellar merger origin of fossil fields (de Mink et al. 2014; Schneider et al. 2016). In this context, confirmation of magnetic ApBp stars in short-period binary systems gives support to alternative theories or, at least, demonstrates that early-type stars may acquire magnetic fields through different channels. In addition, detached close binary stars, particularly those showing eclipses, are valuable astrophysical laboratories that provide model-independent stellar parameters and allow one to study pairs of co-evolving stars formed in the same environment. Until recently, no early-type magnetic stars in eclipsing binaries were known. The first such system, HD 66051, was identified by Kochukhov et al. (2018). The second system, HD 62658 containing twin components of which only one is magnetic, was found by Shultz et al. (2019). Several other eclipsing binaries containing candidate ApBp stars were proposed (Hensberge et al. 2007; González, Hubrig & Castelli 2010; Skarka et al. 2019), but the magnetic nature of these stars has not been verified by direct detections of their fields using the Zeeman effect. In this paper, we put a spotlight on another candidate eclipsing magnetic Bp star, which received little attention prior to our work despite being significantly brighter than the confirmed magnetic eclipsing systems HD 62658 and HD 66051.
Citation: Alecian et al. 2015
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Citation: Kochukhov et al. (2018)
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Citation: Shultz et al. (2019)
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Citation: Gerbaldi, Floquet & Hauck 1985
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Citation: Schneider et al. 2016
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In our model atom, singlet and triplet terms are often separated, and thus we would
            like to estimate how this total cross-section is distributed among final spin states. We
            considered a representation where the spin of the Rydberg electron is coupled to the
            spin of the other electrons in the target atom Sc (assuming
            this to be a good quantum number) to give the total electronic spin of the Rydberg atom,
              S.
            Recoupling of angular momenta gives the required spin-changing scattering amplitude in
            terms of the singlet and triplet scattering amplitudes (i.e. in the representation in
            which the Rydberg electron spin and the hydrogen atom spin are coupled), and the
            differential cross-section for change of spin S → S′, S ≠ S′,
            is found to be \begin{eqnarray} |f_{\rm e}(p,\theta,S \rightarrow S')|^2 &amp; = &amp; \frac{(2S'+1)}{8(2S_{\rm c}+1)} |f^+ - f^-|^2 \\[3mm]  &amp; = &amp;\frac{(2S'+1)}{2(2S_{\rm c}+1)} \sigma_{\rm exch} \end{eqnarray}




|

f

e


(
p,θ,S
→

S

′


)

|

2






=









(
2

S

′


+
1
)


8
(
2

S

c


+
1
)





|

f

+


−

f

−



|

2












=









(
2

S

′


+
1
)


2
(
2

S

c


+
1
)






σ

exch






where σexch is the
            differential cross-section for electron exchange in elastic e−+H  collisions (e.g. Field 1958; Burke
              & Schey 1962). The derivation essentially follows Dalgarno & Rudge (1965), except that in our case the electron
            spin is coupled to the core electronic spin and not to the nuclear spin. We note that
              S′ =
                Sc ± 1/2 and the factor
              (2S′ + 1) /
                2(2Sc + 1) is always in the range 0 to
            1 and represents the probability that an electron exchange leads to a change in the
            total electronic spin of the Rydberg atom. In the limit p → 0, − f± →
                a±, where a± are the
            scattering lengths, which have values of a+ = 5.965 and a− = 1.769
            atomic units (Schwartz 1961). In this limit, for
            the case of Sc =
              1/2, via comparison of the expressions for the cross sections we
            find that \begin{eqnarray} \sigma_{nl,S=0 \rightarrow n'l',S'=0} = 0.706 \: \sigma_{nl \rightarrow n'l'}, \\ \sigma_{nl,S=0 \rightarrow n'l',S'=1} = 0.294 \: \sigma_{nl \rightarrow n'l'}, \\ \sigma_{nl,S=1 \rightarrow n'l',S'=0} = 0.098 \: \sigma_{nl \rightarrow n'l'}, \\ \sigma_{nl,S=1 \rightarrow n'l',S'=1} = 0.902 \: \sigma_{nl \rightarrow n'l'}. \end{eqnarray}





σ

nl,S
=
0
→

n

′



l

′


,

S

′


=
0


=
0.706

σ

nl
→

n

′



l

′




,














σ

nl,S
=
0
→

n

′



l

′


,

S

′


=
1


=
0.294

σ

nl
→

n

′



l

′




,














σ

nl,S
=
1
→

n

′



l

′


,

S

′


=
0


=
0.098

σ

nl
→

n

′



l

′




,














σ

nl,S
=
1
→

n

′



l

′


,

S

′


=
1


=
0.902

σ

nl
→

n

′



l

′




.
Citation: Dalgarno & Rudge (1965)
Citation function: Uses; Compare/Contrast