Mass flows and their behaviors in the SS 433–W 50 system

Inoue, Hajime

doi:10.1093/pasj/psac050

Abstract

We propose a scenario to interpret the overall observational features of the SS 433–W 50 system. The most unique features of SS 433 are the presence of the precessing, mildly relativistic jets and the obscuration of the central engine, which are considered to be due to a supercritical accretion on to the central compact object. The jets are likely to be ejected from the innermost region of the accretion flow. The concept of the accretion ring (Inoue, 2021, PASJ, 73, 795) is applied to the outer boundary of the accretion flow and the ring is supposed to have a precession. The accretion ring is expected to extend a two-layer outflow of a thin excretion disk and a thick excretion flow, as well as the accretion flow. The thin excretion disk is discussed as eventually forming the optically thick excretion belt along the Roche lobe around the compact object, contributing to the obscuration of the central engine. The thick excretion flow is likely to turn to the supersonic wind (disk wind) with the terminal velocity of ∼10⁸ cm s⁻¹ and to collide with the supernova remnant (SNR) matter at the distance of ∼10¹⁸ cm. The interactions of the jets with the disk wind are considered to cause the features of the jets observed at the distances of 10¹⁴–10¹⁵ cm and ∼10¹⁷ cm. Finally, it is discussed that the jets are braked by the SNR matter at the distance of ∼10 pc and the momentum carried by the jet is transferred to the SNR matter shoved by the jet. The SNR matter pushed to the inside of the precession cone is expected to gather along the cone axis and to form the elongated structures in the east and west directions from the main W 50 structure.

1 Introduction

SS 433 is well known to steadily exhibit precessing and mildly relativistic jets. The motion was discovered in the observations of optical emission lines and soon found to be nicely explained by the kinematic model, in which matter is ejected in two opposing jets that are collimated and oppositely aligned to within a few degrees (see Margon 1984 for the early review). The ejection velocity proves to be 0.26c (c: the velocity of light), the axis of the jets rotates with a period of 162.5 d, the central axis of the rotation cone is inclined by 79° from the line of sight, and the half-angle of the cone is 20°.

Further observations revealed the presence of two other periodicities in the SS 433 system than the preccessing period: the binary period of 13.1 d (Crampton et al. 1980; Cherepashchuk 1981) and the nodding period of 6.3 d (Katz et al. 1982).

From observational properties periodically varying with the binary period, a number of studies have been done to determine whether the compact object in the SS 433 system is a neutron star or a black hole [see subsections 7.2 and 7.3 of Fabrika (2004) for the brief review as of early 2000; Cherepashchuk et al. (2018) and references therein for recent progress]. Although recent studies favor the black hole hypothesis, it is not conclusive yet.

The bolometric luminosity of SS 433 is of the order of 10⁴⁰ erg s⁻¹ [see subsection 6.6 of Fabrika (2004) and references therein], indicating that a supercritical accretion on to a compact object takes place in the system. The observed spectrum is, however, approximated by a blackbody spectrum with the temperature of several 10⁴ K and the blackbody radius is as large as a few times 10¹² cm. This emission is understood to be a result of reprocessing the direct emission from the central engine by an obscuring matter on the line of sight.

According to the current observational study of binary X-ray sources [see, e.g., Inoue (2022a) for the review on X-ray observations of accretion disks], the emission from an accretion flow on to a compact object is mainly observed in the soft X-ray (∼1–10 keV) band. The observed soft X-ray luminosity from SS 433 is, however, as low as ∼10³⁶ erg s⁻¹ and is, furthermore, realized to come from the jets ejected from the central engine. It is considered that the soft X-rays from the central engine are completely obscured and that only the hard X-rays above ∼20 keV with the luminosity of several times 10³⁵ erg s⁻¹ are observed from an extended hot corona on an accretion disk (Cherepashchuk et al. 2009, 2013).

Several emission lines are detected in the soft X-ray band and well interpreted as results of thermal emission from adiabatically and radiatively cooling conical jets with distances of the order of 10¹² cm from the center (Kotani et al. 1996).

With the detailed jet model, Brinkmann and Kawai (2000) estimated the kinetic luminosity of the jets ∼6 × 10³⁹ erg s⁻¹.

The X-ray jets end at a distance of 2–3 × 10¹³ cm and the optical jets replace the X-ray jets on the further side. The optical emission from the jets often exhibits time variations on a time scale of days, which are interpreted by appearances of cool clouds and their modulations in the jets as a result of thermal instability induced in the jet where the temperature decrease to ∼0.1 keV. The active region of the optical lines from the clouds begins at ∼10¹⁴ cm, has the activity-maximum around several times 10¹⁴ cm, and ends at ∼10¹⁵ cm. For a review on the properties of the optical jets, see section 2 of Fabrika (2004).

The lifetimes of the moving clouds emitting optical lines are observed to be several days, while the cooling time is estimated to be much shorter. Thus, the matter must be continuously reheated. The interaction of the jets with the accretion-disk wind is currently thought to be the most likely origin for the heating [see subsection 5.3 of Fabrika (2004) and references therein].

Radio blobs are resolved in the optical-jet region with VLBI (very long baseline interferometry) observations. The radio strength appreciably weakens as they move outward but grows again as they pass through the “radio brightening zone” at the distance of ∼4 × 10¹⁵ cm from the center. The radio blobs rapidly fade beyond the brightening zone. For a review on the radio properties of the jets, see sections 3 of Fabrika (2004).

The radio interferometric observations also reveal the existence of matter in the plane perpendicular to the jets (Paragi et al. 1999; Blundell et al. 2001). The structure extends to 3–4 × 10¹⁵ cm on both sides of the central source. Outward proper motion of a radio component is detected and the speed is estimated to be ∼10⁸ cm s⁻¹.

The brightness of the radio jets becomes the maximum at a distance of ∼10¹⁵ cm and gradually decreases to ∼10¹⁷ cm, beyond which the jets are not visible until ∼10²⁰ cm. The X-ray imaging observation with Chandra, however, indicates that the jets are reheated so hot as to emit X-rays at the distance of ∼10¹⁷ cm (Migliari et al. 2002).

The jets terminate at a distance of several 10 pc from the center, in the elongated structures in the east and west directions of the radio nebula W 50. W 50 has a shell-like structure with radius of ∼40 pc except the elongated regions and is thought to be a supernova remnant (SNR) (see Margon 1984). Using high-resolution radio imaging observations spanning a 12-year period, the proper motions of the radio filaments across the W 50 nebula were analyzed and the upper limit of ∼1 × 10⁹ cm s⁻¹ was obtained for them (Goodall et al. 2011).

The elongated structures extending to a distance of ∼100 pc are considered to be formed as a result of interactions between the jets and the interstellar matter. The eastern and western optical filaments lie inside the projection of the jet precession cone and are perpendicular to the jet. Extended X-ray jets are also detected with fringing of the optical filaments (Watson et al. 1983).

In this paper, we try to interpret the observational properties of the SS 433–W 50 system briefly reviewed above as widely as possible using a model situation. The model situation is introduced in section 2, interpretations of the observations are presented in section 3, and the summary is given in section 4.

2 Model situation

We suppose the following situation.

A binary system composed of a normal star and a compact object and having an orbital period of 13.1 d stays at the center of a SNR having a spherical shell with a radius of ∼40 pc.

Although the masses of the two stars have not yet been determined precisely from observations, we assume the compact object to be a black hole with ∼10 |$M_{\odot}$| and adopt ∼20 |$M_{\odot}$| as the typical mass of the normal star here. These assumptions have no large effect on the present study.

The separation of the two components, a, in the binary with the orbital period of 13.1 d is estimated as

$$\begin{equation} a = 4.4 \times 10^{12} (1+q)^{1/3} \left(\frac{M_{\rm n}}{20\ M_{\odot }}\right)^{1/3}\:\mbox{cm}, \end{equation}$$

(1)

where q is the ratio of the compact star mass, M_x, to the normal star mass, M_n. Considering M_n ∼ 20 |$M_{\odot}$| and q < 1, the separation is several times 10¹² cm.

In the binary, matter flows from the normal star through the L1 point to the compact object and forms an accretion ring at the Keplerian circular orbit around the compact object determined by the specific angular momentum of the inflowing matter. According to the approximations done in subsections 4.4 and 4.5 of Frank, King, and Raine (2002), the radius of the accretion ring, R₀, is roughly given as

$$\begin{equation} \frac{R_{0}}{a} \simeq \frac{1+q}{q} (0.5 + 0.227 \log q)^{4}, \end{equation}$$

(2)

and is estimated to be ∼0.10 for q ∼ 0.5. On the other hand, the radius of the Roche lobe on the side of the compact star, R₁, is approximated as

$$\begin{equation} \frac{R_{1}}{a} \simeq \frac{0.49 q^{2/3}}{0.6q^{2/3}+\ln (1+q^{1/3})} \end{equation}$$

(3)

(Eggleton 1983), and is 0.32 for q = 0.5. For a ∼ several times 10¹² cm, R₀ and R₁ are considered to be several times 10¹¹ cm and ∼10¹² cm, respectively.

According to a recent study by Inoue (2021a), it is supposed that the accretion ring consists of a geometrically thick envelope and a geometrically thin core, and that angular momenta are transferred from the inner side facing the compact object to the opposite side in the envelope and the core, respectively. As a result, a thick accretion flow and a thick excretion flow extend from the envelope, and a thin accretion disk and a thin excretion disk do so from the core. The schematic configuration around the accretion ring in SS 433 is shown in figure 1.

Fig. 1.

Schematic cross-section of the central part of the system considered in the present scenario. The four diagrams indicate the situation in which the central part precesses around the angular momentum axis of the accretion ring.

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Let us define |$\dot{M}_{0}$|⁠, |$\dot{M}_{\rm ac, f}$|⁠, |$\dot{M}_{\rm ex, f}$|⁠, |$\dot{M}_{\rm ac, d}$|⁠, and |$\dot{M}_{\rm ex, d}$| as the mass inflow rate to the accretion ring from the normal star, and the mass flow rates of the thick accretion flow, the thick excretion flow, the thin accretion disk, and the thin excretion disk, respectively. Then, we assume that

$$\begin{equation} \dot{M}_{\rm ac, f} = \dot{M}_{\rm ex, f} = \frac{F\,\, \dot{M}_{0}}{2}, \end{equation}$$

(4)

and

$$\begin{equation} \dot{M}_{\rm ac, d} = \dot{M}_{\rm ex, d} = \frac{(1-F)\,\, \dot{M}_{0}}{2}. \end{equation}$$

(5)

Here, F is the fraction of the total flow rate to the thick accretion and excretion flows in the total inflow rate, and is supposed to be several to |$10\%$| when the inflow rate is as large as or larger than the Eddington rate (Inoue 2021a). The equalities of the mass flow rate between the accretion flows and the excretion flows are derived from the study by Inoue (2021a)—see also the appendix in Inoue (2021b).

For the accretion flow on to the central compact object, a situation is expected to happen in which a geometrically thin standard disk (Shakura & Sunaev 1973) is sandwiched by a geometrically thick and optically thin advection dominated accretion flow (ADAF) (Narayan & Yi 1994). The thin standard disk is considered to turn to an optically thick ADAF called as the slim disk (Abramovicz et al. 1988) in the inner region because of the supercritical accretion rate.

Jets are ejected from the innermost region of the slim disk in two opposite directions perpendicular to the equatorial plane of the disk. The jet velocity, v_j, is assumed to be 0.26c as observed.

The flow rate of the specific angular momentum on to the black hole is |$\sim \sqrt{r_{\rm in}GM}$| for the innermost radius of the accretion flow, r_in, and is negligibly small compared with the inflow rate to the accretion ring, j₀, given as

$$\begin{equation} j_{0}= \sqrt{R_{0}GM}, \end{equation}$$

(6)

since r_in ≪ R₀. Thus, almost all the angular momentum of the accreted matter should be transferred to the excreted matter across the boundary layer in the accretion ring irrespectively of whether the flow is thick or thin. As a result, the specific angular momentum carried by the excreted matter, j_ex, should be twice as large as the intrinsic inflow rate, j₀, written as

$$\begin{equation} j_{\rm ex} \simeq 2\,\, j_{0}. \end{equation}$$

(7)

The angular momentum transfer is carried out with stress due to turbulent viscosity in the boundary layer, and the total torque given to the innermost side of the excretion flow is |$\dot{M}_{\rm ac}\,\, j_{0}$|⁠. Hence, the energy transfer rate in association with the angular momentum transfer to the excretion flow should be |$\dot{M}_{\rm ac} j_{0} \Omega _{0}$|⁠, where Ω₀ is the angular velocity of the inflowing matter at R₀. Considering |$\dot{M}_{\rm ex} = \dot{M}_{\rm ac}$|⁠, |$j_{0}=\sqrt{R_{0}GM}$|⁠, and |$\Omega _{0}=\sqrt{GM/R_{0}^{3}}$|⁠, the specific energy given to the excretion flow, ϵ_out, is estimated as

$$\begin{equation} \varepsilon _{\rm out} = j_{0}\,\, \Omega _{0} \simeq \frac{GM}{R_{0}}. \end{equation}$$

(8)

In the case of a thin excretion disk, the matter outflows slowly by adjusting the centrifugal force to balance with the gravitational force and radiating the surplus energy via the blackbody emission. The outflow of the excretion disk should terminate at the Keplerian circular orbit, r_ex, determined by the specific angular momentum j_ex given in equation (7), which is calculated as

$$\begin{equation} r_{\rm ex} = \frac{j_{\rm ex}^{2}}{GM} \simeq 4 \frac{j_{0}^{2}}{GM} = 4 R_{0}. \end{equation}$$

(9)

Since the Roche lobe radius around the compact star, R₁, is calculated to be ∼3R₀ for q ∼ 0.5 from equations (2) and (3), this radius, r_ex, is slightly larger than the Roche lobe radius around the compact star. Thus, the geometrically thin excretion disk would, in practice, significantly suffer the tidal effect from the normal star and could form a belt-like structure along the Roche lobe, eventually merging with the normal star atmosphere, as schematically depicted in figure 2. We call this structure the “excretion belt” hereafter.

Fig. 2.

Schematic view of the excretion belt from top (upper) and on the meridian cross-section (bottom).

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In the case of a thick excretion flow, the flow is considered to be advection-dominated. Then, the specific energy can be approximated to be conserved and expressed as

$$\begin{equation} \frac{v^{2}}{2} + w - \frac{GM}{r} = \varepsilon _{\rm out}, \end{equation}$$

(10)

where v is the flow velocity, w is the specific enthalpy, and ϵ_out is the constant specific energy of the thick excretion flow. This specific energy is the quantity given in equation (8) on the assumption that the specific energy of the inflowing matter to the accretion ring is zero. The flow is expected to become supersonic and the terminal velocity, v_∞, is estimated as

$$\begin{eqnarray} v_{\infty } &\simeq& \sqrt{2GM/R_{0}}\nonumber\\ &&= 9 \times 10^{7} \left(\frac{M}{10\,\, M_{\odot }}\right) \left(\frac{R_{0}}{10^{11.5}\:\mbox{cm}}\right)^{-1}\:\mbox{cm}\:\mbox{s}^{-1}.\nonumber \\ \end{eqnarray}$$

(11)

The accretion ring precesses with the period T_p = 162.5 d around the rotational axis of the binary and the half-opening angle of the precessing cone, θ_p, is 20°. Because of the precession of the accretion ring, all the flows—the excretion flows, accretion flows, and the jets—can be considered to precess simultaneously.

3 Interpretations of observations

3.1 Precession of the accretion ring

The observations of the jet motions clearly indicate the precession of the jet axis. Considering that the jets are ejected in the directions perpendicular to the accretion disk, Whitemire and Mates (1980), van den Heuvel, Ostriker, and Petterson (1980), and Katz (1980) proposed the slaved disk model for SS 433, in which the disk precesses slaved to the precession of the companion star but the stability of the precession is questionable. Inoue (2012, 2019) discusses that the precession of the accretion ring induced by the tidal force from the companion star is rather stable.

The power spectral analysis of the Doppler shifts observed from the SS 433 jets by Katz et al. (1982) shows the presence of two periods at 5.8 d and 6.3 d. Katz et al. (1982) explain the 6.3 d period but fail the 5.8 d period in the frame of the slaved disk model, while Inoue (2012) gives a simpler interpretation of both the periods through the nodding motion of the jet axis.

3.2 Hard X-ray emission

Cherepashchuk et al. (2013) present the hard X-ray (above 20 keV) light curve folded with the precession period. It looks similar to the X-ray light curve of Her X-1 folded with the super-orbital period of 35 d (see, e.g., Inoue 2019). The super-orbital light curves of Her X-1 and two other X-ray pulsars are shown to be well reproduced by a model in which X-rays from a compact object towards us are periodically obscured by a precessing ring at the outermost part of an accretion disk around the central object, from Inoue (2019). These suggest that the periodic obscuration of the central X-ray source by the precessing accretion ring takes place even in SS 433. If so, the hard X-rays should come from a region further in than the accretion ring. Indeed, the analysis of the shape and width of the primary hard X-ray eclipses indicates that the hard X-ray emission (20–100 keV) is formed in an extended, hot, quiasi-isothermal corona on an accretion disk (Cherepashchuk et al. 2009).

The hard X-ray emission above ∼20 keV has a luminosity of several times 10³⁵ erg s⁻¹ (Cherepashchuk et al. 2003) and a power-law spectrum with a photon index ∼3 or larger (Cherepashchuk et al. 2009, 2013). These characteristics are consistent with a picture that the accretion disk in SS 433 is in the slim disk configuration (Inoue 2022a) and that the soft X-rays from the disk are completely obscured by the photo-electric absorption, while the hard X-rays are attenuated by electron scattering. Kubota and Makishima (2004) analyzed the data of the black hole binary XTE J1550-564 during the outburst and showed that the spectrum of the source in the slim disk configuration has a steep hard tail above the main soft component and that the luminosity of the power-law component is about |$1\%$| of that of the soft component apparently saturating at the Eddington luminosity. If we assume the mass of the compact object in SS 433 is ∼10 |$M_{\odot}$|⁠, the soft X-ray luminosity should be ∼10³⁹ erg s⁻¹ and thus the hard X-ray luminosity is expected to be ∼10³⁷ erg s⁻¹. Hence, the observed hard X-ray luminosity of several times 10³⁵ erg s⁻¹ requires an intensity attenuation by a factor of slightly more than 10. A part of the attenuation would be given by the accretion ring and the other part would be by the excretion belt, which is discussed is subsection 3.4.

3.3 Jet ejection

Recently, Inoue (2022b) proposed a mechanism for the innermost region of a slim disk to eject the bi-polar jets steadily, and the model can be applied to the SS 433 jets.

Most observations indicate that the slim disk intermittently appears in the limit cycle activities in the very high state [see section 5 of Inoue (2022a), while the jet activity of SS 433 seems to be basically steady. Thus, if the jets really come from the innermost region of the slim disk, it is required that the slim disk steadily exists in the SS 433 system. Indeed, the activities of Type II bursts from the Rapid Burster, which are understood to be the result of the limit cycle activity [see, e.g., subsection 5.3 in Inoue (2022a)], appear in the decay portions of the outbursts and the cyclic period becomes shorter as the source intensity decreases (Guerriero et al. 1999). These are naturally interpreted by a picture that the recurrent period of the limit cycle becomes so long as for the source to be practically in the steady slim disk phase when the accretion rate is very high.

3.4 Obscuration of the central engine

X-ray observations of the jets clearly indicate that the observed soft X-rays below ∼10 keV come from the jets and the bright soft X-rays expected from the central part of the accretion disk should be almost completely absorbed via the photoelectric absorption. Even the X-rays observed from the jets are limited to those from the region with the plasma temperature below ∼20 keV (Kotani et al. 1996) and then the central part with the higher temperature should be obscured. It is further pointed out by Kotani et al. (1996) that the receding jet is more obscured than the preceding jet and that the matter responsible for the obscuration should have a size significantly larger than the accretion disk.

The matter outflowing through the thin excretion disk is likely to contribute significantly to the obscuration of the central engine and the central part of the jets. As discussed earlier, the matter is considered to form the excretion belt along the Roche lobe eventually and to merge with the atmosphere of the optical star finally (see figure 2). Since the excretion disk precesses with the precessing angle, θ_p ≃ 20°, the excretion belt would have the width of ∼2R₁sin θ_p and the total flow length of ∼2πR₁, and thus the total surface area would be |$\sim 4\pi R_{1}^{2} \sin \theta _{\rm p}$|⁠. Then, if the staying time of the matter in the excretion belt is assumed to be the Keplerian circular period around the compact star given as |$2\pi R_{1}^{3/2}/(GM_{\rm x})^{1/2}$|⁠, the average optical depth for the Thomson scattering of the excretion belt, τ_eb, is approximately estimated to be

$$\begin{equation} \tau _{\rm eb} \simeq \frac{\dot{M}_{\rm ex, d}}{2 (GM_{\rm x})^{1/2} R_{\rm 1}^{1/2} \sin \theta _{\rm p}} \frac{\sigma _{\rm T}}{m_{\rm p}}, \end{equation}$$

(12)

where σ_T is the cross-section for the Thomson scattering. The τ_eb value for |$\dot{M}_{\rm ex, d} \simeq 10^{20}\:$|g s⁻¹, M_x ≃ 10 |$M_{\odot}$|⁠, and R₁ ≃ 10¹² cm is 1.6, which is consistent with the expectation for the obscuration.

Note that the net photoelectric cross-section per hydrogen atom of matter with the cosmic abundance steeply decreases as the photon energy increases and is about the same as the Thomson cross-section at ∼10 keV (see, e.g., Morrison & McCammon 1983). Hence, the situation can reasonably be understood as that the soft X-rays below ∼10 keV from the central engine are completely absorbed, while the hard X-rays above ∼10 keV are only attenuated.

3.5 Interaction between the jet and the disk wind

As shown in section 1, the observed properties of jets around the 10¹⁴ to 10¹⁵ cm distance are currently explained by the interaction of the jets with the disk wind. The radio structures in the plane perpendicular to the jets (Paragi et al. 1999; Blundell et al. 2001) support the presence of the disk wind. In the frame of the accretion ring concept, the disk wind can naturally be explained by the thick excretion flow. Hereafter, we call the thick excretion flow as the disk wind, following the current terminology.

Let us consider a situation in which an element of the jet is ejected with the velocity, v_j, at a moment of t = 0 in a radial direction on the surface of the precessing cone with the opening half-angle of θ_p, and where the rotational axis of the accretion ring yielding the disk wind coincides with the jet direction at t = 0. The time for the jet element to reach a distance r from the center, t_je,r, is given as t_je,r = r/v_j. The disk-wind matter around the jet element at r is considered to have started from the center at t_dw,r,0 = t_je,r − r/v_dw = −[1 − (v_dw/v_j)]r/v_dw. Hence, a difference of the precession phase angle of the disk-wind matter from that at t = 0, ϕ, is calculated as

$$\begin{equation} \frac{\varphi }{2 \pi } = \frac{t_{\rm dw, r, 0}}{T_{\rm p}} = -\frac{r/v_{\rm dw}}{T^{\prime }_{\rm p}}, \end{equation}$$

(13)

where

$$\begin{equation} T^{\prime }_{\rm p} = \left( 1 - \frac{v_{\rm dw}}{v_{\rm j}} \right)^{-1} T_{\rm p} \simeq T_{\rm p}. \end{equation}$$

(14)

Then the angle θ_jw between the jet-element direction and the normal axis of the disk wind at r can approximately be estimated from the isosceles spherical triangle drawn in figure 3 as

$$\begin{equation} \cos \theta _{\rm jw} = \cos ^{2} \theta _{\rm p} + \sin ^{2} \theta _{\rm p} \cos \varphi . \end{equation}$$

(15)

As seen from this equation, θ_jw increases from zero to the maximum of 2θ_p ≃ 40° as ϕ increases from zero to π/2 in association with the increase of r from zero to r_opt, given as

$$\begin{equation} r_{\rm opt} \simeq v_{\rm w} \frac{T_{\rm p}}{2} = 7.0 \times 10^{14} \left(\frac{v_{\rm w}}{10^{8}\:\mbox{cm}\:\mbox{s}^{-1}}\right)\:\mbox{cm}. \end{equation}$$

(16)

Since the elevation angle of the matter in the disk wind from the equatorial plane is (π/2) − θ_jw, the density of the wind matter interacting with the jet matter is expected to increase as θ_jw increases. This can explain well the optical line activity strengthened at the distance around several times 10¹⁴ cm as briefly introduced in section 1.

Fig. 3.

Isosceles spherical triangle formed by the z-axis, the jet direction at t = 0 which is on the x–z plane, and the disk axis at t = −t′ which rotates an angle ϕ around the z-axis from the x–z plane. Another spherical triangle formed by the z-axis, the disk axis at t = −t′, and the crossing point of the small circle depicted by the disk wind edge around the disk axis with the x–z plane is also drawn for the calculations in appendix 1. The inclination angle of the precession axis from the z-axis is θ_p, the angular radius of the disk edge circle is θ_d and the angle of the crossing point from the z-axis is θ.

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The above configurations are schematically drawn in figure 4. This figure exhibits the shape of the disk wind edge on the meridian cross-section along the jet axis at a certain moment, where the elevation angle of the disk wind edge is assumed to be 60° (the opening angle of the wind ∼120°). The typical scale of this figure is v_wT_p and the jets run almost instantaneously over this scale since the jet speed is much faster than v_w. Hence, this figure approximately shows how the jets intersect with the disk wind. See appendix 1 for the calculation of the shape of the disk wind edge.

Fig. 4.

Cross-section of the disk wind on the meridian plane including the angular-momentum axis of the precession and the direction of the jet ejection at a moment. The opening angle of the disk wind is assumed to be 2 × 60°. A bullet of the jet starts from the center, and first interacts with the disk-wind matter around the distance of v_w T_p/2, which could cause the enhancement of the optical lines.

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When the jet matter approaches to the distance of

$$\begin{equation} v_{\rm w} T_{\rm p} = 1.4 \times 10^{15} \left(\frac{v_{\rm w}}{10^{8}\,\mbox{cm}\:\mbox{s}^{-1}}\right)\:\mbox{cm}, \end{equation}$$

(17)

its position becomes perpendicular to the equatorial plane of the disk wind at that position and the ambient density of the disk wind is likely to be very low. When it further advances to the distance of (3/2)v_wT_p ≃ 2.1 × 10¹⁵ cm, θ_jw again increases to 40°. However, the density of the ambient disk wind in the cylinder-like region through which the jet has once passed is expected to be kept very low after sweeping up of the disk wind matter by the jet as discussed in subsection 5.2 of Fabrika (2004).

The matter in a particular region of the disk wind meets the jet matter every precession period as it advances outward. It should be shoved away at the first encounter with the jet and a cylinder-like space with very low density should appear and co-move with the ambient disk-wind matter. Since the radius of the cylinder-like region is rΘ_j at a distance r, the time scale for the ambient matter to refill the empty region, t_rf, could be approximated as t_rf ∼ (rΘ_j)/c_s, where c_s is the sound velocity of the ambient disk-wind matter. When the refilling time scale is much longer than the precession period, the situation of very low density in the cylinder-like region can be considered to be kept at the successive encounters after the first one.

Considering that the cooling time of the disk-wind matter is much shorter than the traveling time to the distance of ∼10¹⁵ cm, the temperature of the wind matter should be no higher than 10⁴ K and so the sound velocity, c_s, should be no larger than 10⁶ cm s⁻¹. Hence, we can say that the swept-up situation is kept in a region of

$$\begin{equation} r \gg c_{\rm s}\,\, T_{\rm p} = 1.4 \times 10^{15} \left(\frac{c_{\rm s}}{10^{6}\:\mbox{cm}\:\mbox{s}^{-1}}\right). \end{equation}$$

(18)

For c_s ≲ 10⁶ cm s⁻¹, the swept-up situation in the disk wind can be expected to be kept after the first encounter with the jet. This also supports the argument for the radio brightening zone because of the free expansion of the clouds in the jet, as in subsection 5.2 of Fabrika (2004).

3.6 Interaction of the disk wind with the SNR matter

It is known that SS 433 is in the radio nebula W 50 which has the spherical structure with the radius of ∼40 pc except the elongated structures in the east and west directions. This nebula is currently considered to be a SNR.

The disk wind radially flows within a boundary elevation angle from the equatorial plane. Since the ram pressure at the front of the disk wind decreases as it advances pushing the SNR matter, the wind front should significantly be decelerated by the SNR matter at some distance. As discussed in appendix 2, the quasi-steady situation is expected to be realized when the position of the backward shock becomes stationary. The distance of the stationary backward shock, r_bs,dw, is determined from A = 1/9, and A is given in equation (A15). Here, ρ_st in appendix 2 should be replaced with the density of the SNR matter in W 50, ρ_snr, and ρ_sf,bs should be the density of the disk wind on the upstream side of the backward shock, ρ_dw,bs, which is given as

$$\begin{equation} \rho _{\rm dw, bs} \simeq \frac{\dot{M}_{\rm dw}}{4\pi r_{\rm bs, dw}^{2} \cos \alpha _{\rm dw} v_{\rm w}}, \end{equation}$$

(19)

where |$\dot{M}_{\rm dw} (= \dot{M}_{\rm ex, f}$|⁠) is the mass flow rate through the disk wind, and α_dw is the elevation angle of the upper boundary of the disk wind from the equatorial plane. ξ is set at unity for the disk wind. From these equations, the position of the backward shock, r_dw,bs, is approximately calculated as

$$\begin{eqnarray} r_{\rm dw, bs} &\simeq & 6.6 \times 10^{17} \left(\frac{\dot{M}_{\rm dw}}{10^{19}\:\mbox{g}\:\mbox{s}^{-1}}\right)^{1/2} \left(\frac{v_{\rm w}}{10^{8}\:\mbox{cm}\:\mbox{s}^{-1}}\right)^{-1/2}\nonumber \\ &&\times \left(\frac{\rho _{\rm snr}}{10^{-24}\:\mbox{g}\:\mbox{cm}^{-3}}\right)^{-1/2}\:\mbox{cm}, \end{eqnarray}$$

(20)

where α_dw = 60° is adopted. Hence, the stationary backward shock is expected to appear at the distance of 7 × 10¹⁷ cm for |$\dot{M}_{\rm dw} \sim 10^{19}\:$|g s⁻¹, v_w ∼ 10⁸ cm s⁻¹, and ρ_snr ∼ 10⁻²⁴ g cm⁻³.

The heating of the disk wind matter should take place behind the backward shock and could induce refilling of the matter-swept-up region in the jet-path during a cycle of the jet precession, as discussed in the next paragraph. This could cause reheating of the jet matter through the interaction with the refilled disk matter, possibly providing the origin of the jet structures observed in the X-ray band around ∼10¹⁷ cm distance (Migliari et al. 2002). This distance of ∼10¹⁷ cm seems smaller than the above estimated distance of the backward shock for the reasonable parameters. However, since the interaction should takes place in a region with an elevation angle above 50° from the disk-wind equatorial plane, it is expected that the ram pressure of the disk wind is significantly weaker than the typical value there and that the front shock distance is significantly smaller than the typical value.

As discussed in the previous subsection, the time scale for the ambient disk-wind matter to refill the empty region along the jet path at a distance, r, is approximated as rΘ_j/c_s. Then, the refilling could be effective when the refilling time-scale is no longer than the precession period, and we can get the following condition for the effective refilling to happen:

$$\begin{equation} r \lesssim \frac{c_{\rm s} T_{\rm p}}{\Theta _{\rm j}} = 7 \times 10^{16} \left(\frac{c_{\rm s}}{10^{8}\:\mbox{cm}\:\mbox{s}^{-1}}\right)\:\mbox{cm}. \end{equation}$$

(21)

Considering c_s ∼ v_dw ∼ 10⁸ cm s⁻¹ on the subsonic side of the backward shock, this is roughly consistent with the reheated position to emit X-rays within the accuracy of the present estimation.

The schematic view of the interaction between the disk wind and the SNR matter and that between the jet and the shocked disk wind is shown in figure 5.

Fig. 5.

Schematic cross-section of the interaction between the disk wind and the SNR matter, and the position of the reheated zone emitting X-rays as the result of the interaction of the jet with the shocked disk-wind matter.

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3.7 Interaction of the jets with the SNR matter

Similarly to the case of the interaction of the disk wind with the SNR matter, the front of the jet should be significantly decelerated by the SNR matter at some distance. According to appendix 2 again, the distance of the stationary backward shock in the jet, r_bs, j, is determined from A = 1/9, and A is given in equation (A15). In this case, ρ_st should again be the density of the SNR matter, ρ_snr, but ρ_sf,bs should be the density of the jet on the upstream side of the backward shock, ρ_j,bs, which is given as

$$\begin{equation} \rho _{\rm j, bs} \simeq \frac{\dot{M}_{\rm j}}{2\pi r_{\rm bs, dw}^{2} \Theta _{rm j}^{2} v_{\rm j}}, \end{equation}$$

(22)

where |$\dot{M}_{\rm j}$| is the total mass flow rate through the two opposite jets, and Θ_j is the half-opening angle of the jet. ξ should be significantly smaller than unity here. From these equations, the position of the backward shock, r_j,bs, is approximately calculated as

$$\begin{eqnarray} r_{\rm j, bs} &\simeq & 11 \left(\frac{\xi }{0.1}\right)^{-1/2} \left(\frac{\dot{M}_{\rm j}}{10^{20}\:\mbox{g}\:\mbox{s}^{-1}}\right)^{1/2} \left(\frac{\Theta _{\rm j}}{{1{_{.}^{\circ}}2}}\right)^{-1/2}\nonumber \\ &&\times \left(\frac{\rho _{\rm snr}}{10^{-24}\:\mbox{g}\:\mbox{cm}^{-3}}\right)^{-1/2}\:\mbox{pc}. \end{eqnarray}$$

(23)

Hence, the stationary backward shock is expected to appear at the distance of ∼10 pc for ξ ∼ 0.1, |$\dot{M}_{\rm j} \sim 10^{20}\:$|g s⁻¹, Θ_j ≃ |$1{_{.}^{\circ}} 2$|⁠, and ρ_snr ∼ 10⁻²⁴ g cm⁻³.

If the position of the backward shock in the jet is located around 10 pc, most of the momentum carried by the jet is likely to be transferred to the SNR matter which is shoved by the front of the jet, as the front advances further in the SNR matter with the velocity of V_cd ≃ v_j/4 ≃ 2 × 10⁹ cm s⁻¹. The SNR matter pushed into the precessing cone of the jet is expected to gather along the cone axis and to advance toward the interstellar space with the speed of ∼10⁹ cm s⁻¹. This could be the origin of the elongated structures seen in radio (Dubner et al. 1998) and X-rays (Watson et al. 1983) in the east and west directions from the main structure in W 50.

The above arguments are schematically visualized in figure 6.

Fig. 6.

Schematic cross-section of the interaction between the jets and the SNR matter.

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4 Summary and discussion

The present study is summarized in figure 7 as a block diagram of the elements discussed above.

Fig. 7.

Block diagram of elements and events in the SS 433–W 50 system.

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The present study starts from adoption of the “accretion ring” concept introduced by Inoue (2021a) to the SS 433 system. It predicts that the accretion ring bears a thick excretion flow and a thin excretion disk. The thick excretion flow is considered to become a supersonic flow eventually with the terminal velocity of 10⁸ cm s⁻¹, and it plays an important role as the disk wind in the present scenario. The thin excretion disk tends to form a circular ring with a radius four times as large as that of the accretion ring, if there is no gravitational effect from the normal star. In the present scenario for the SS 433 system, the circular ring appears in the form of the excretion belt, which gives a significant contribution to the obscuration of the central engine.

The accretion ring is also discussed to be a place where a precessing motion is naturally excited (Inoue 2012, 2019). The precession seen in the jet motion of SS 433 can be considered to originate at the accretion ring. Indeed, the hard X-ray light curve folded with the precession period resembles the X-ray light curve of Her X-1 folded with the super-orbital period of ∼35 d, while Inoue (2019) reveals that the X-ray light curve of Her X-1 is well reproduced by a model in which X-rays towards us from a compact object are periodically obscured by a precessing accretion ring around a central object. Since the elevation angles of the line of sight to the binary plane are as small as about 10° or lower for both SS 433 and Her X-1, the resemblance of the light curve between SS 433 and Her X-1 suggests that the hard X-rays on the line of sight from SS 433 are periodically obscured by the precessing accretion ring.

In spite of such similarities between SS 433 and Her X-1 as mentioned above, distinct differences of SS 433 from Her X-1 are the presence of the jets and the obscuration of the central engine. These are likely to be due to the different situation of the accretion rate: The supercritical accretion of ∼10²⁰ g s⁻¹ is considered to happen on to the compact object with several solar mass in SS 433, while the accretion rate in Her X-1 is no larger than 10¹⁸ g s⁻¹ against the neutron star of ∼1 solar mass. Such a high accretion rate is expected to cause the jet ejection from the central part of the slim disk (Inoue 2022b). The estimation of the optical depth for Thomson scattering of the excretion belt in equation (12) indicates τ_eb > 1 for |$\dot{M} \sim 10^{20}\:$|g s⁻¹ in the case of SS 433, while τ_eb is estimated to be 0.09 for |$\dot{M} \sim 10^{18}$|⁠, M_x ∼ 1 |$M_{\odot}$|⁠, and R₁ ∼ 3 × 10¹¹ cm in the case of Her X-1.

The jets begin to be observed in soft X-rays after they go out of the central obscured region, and the diagnostics of the X-ray lines are consistent with the adiabatically and radiatively cooling model of the conical jets (Kotani et al. 1996).

As the jet goes out from ∼10¹³ to ∼10¹⁴ cm, the temperature decreases and the mainly observed wave-band shifts from the X-ray band to the optical band. The time variabilities of the optical lines on a time scale of days begin at ∼10¹⁴ cm, have the activity maximum at several times 10¹⁴ cm and end around 10¹⁵ cm. These behaviors are consistent with being due to the first interaction of the jet with the disk wind, as discussed in subsection 3.5.

The matter in the certain region in the disk wind which first encountered the jet at several times 10¹⁴ cm should have been shoved away from the jet path by the jet matter, and the region is likely to get almost empty. The empty region should co-move with the ambient matter in the disk wind and repeatedly encounter the jet every precession period after the first encounter. The time scale for the ambient matter to refill the empty region is, however, estimated to be longer than the precession period in the region given in equation (18), and thus the empty situation of the jet path in the disk wind is kept until the disk wind reaches the distance of ∼10¹⁷ cm.

At the ∼10¹⁷ cm distance, the jet path in the disk wind is expected to enter the subsonic region behind the backward shock induced by the interaction with the SNR matter and to be refilled with the ambient disk-wind matter. The interaction of the jet element with the refilled disk-wind matter could cause reheating the jet element to emit X-rays, as discussed in subsection 3.6.

The jet-element gets into the SNR region after passing through the backward shock region and the jet-path in the SNR matter is likely to be swept up again. The ram pressure of the jet decreases as it goes further and is finally braked by the SNR matter, forming the stationary backward shock at the distance of ∼10 pc. The momentum carried by the jet is expected to transfer to the SNR matter in front of the jet. About half of the SNR matter shoved by the jet could be collimated to the outward direction along the central axis of the precession cone and advance into the interstellar space, forming the front shock and the backward shock. This movement could be the origin of the elongated structure on each of the east and west sides.

As summarized above, the scenario in this paper can basically explain the overall features of the SS 433–W 50 system. However, the arguments are yet semi-quantitative. Further study is obviously needed

Appendix 1 Shape of the disk wind edge

Let us introduce a three-dimensional orthogonal coordinate system (x, y, z) and consider a situation in which the rotationally symmetric axis of the mass ejection (the jets and the disk wind) system precesses along a small circle with angular radius θ_p around the z-axis, as shown in figure 3. The rotationally symmetric axis is called the disk axis here.

We define t as a time from a moment when the disk axis is right on the x–z plane, and calculate the shape of the disk wind edge on the x–z plane at the moment of t = 0.

The disk axis of the disk wind ejected at t = −t′ rotates backward by an angle ϕ from the x–z plane, written as

$$\begin{equation} \varphi = \frac{2\pi }{T_{\rm p}} \ t^{\prime }. \end{equation}$$

(A1)

If the zenith angle of the disk wind edge from the disk axis is θ_d, the edge of the disk wind ejected t′ ago should form the small circle with the angular radius of θ_d around the disk axis at that time. Then, the angle of the direction of the disk wind edge on the x–z plane at t = 0 from the z-axis, θ, is calculated with the help of the laws of sine and cosine for the non-isosceles spherical triangle as shown in figure 3 as

$$\begin{eqnarray} &&\cos \theta = \nonumber\\ &&\frac{\cos \theta _{\rm p} \cos \theta _{\rm d}\! +\! \sin \theta _{\rm p} \sin \theta _{\rm d} \cos \varphi \sqrt{1\! -\! \sin ^{2} \theta _{\rm p} \sin ^{2} \varphi /\sin ^{2} \theta _{\rm d}}}{1\!-\!\sin ^{2} \theta _{\rm p} \sin ^{2} \varphi }. \nonumber\\ \end{eqnarray}$$

(A2)

On the other hand, the matter in the disk wind ejected t′ ago is considered to advance radially over the distance, r, by t = 0, given as

$$\begin{equation} r = v_{\rm dw} t^{\prime }. \end{equation}$$

(A3)

As the results, the position of the edge of the disk wind ejected t′ ago on the x–z plane at the moment, x(t′) and z(t′), are calculated as

$$\begin{eqnarray} x(t^{\prime }) &=& r \sin \theta , \end{eqnarray}$$

(A4)

$$\begin{eqnarray} z(t^{\prime }) &=& r \cos \theta . \end{eqnarray}$$

(A5)

Figure 4 plots the calculated results of (x, z) for θ_p = 20° and θ_d = 30° over two precession periods for t′.

Appendix 2 Interaction of supersonic flow with the ambient matter

First, we consider the interaction of a highly supersonic flow with stationary matter in front of the flow, one-dimensionally. This is the case in which the disk wind collides with the ambient SNR matter.

Let ρ_st, ρ_sf, and v_sf be the density of the stationary matter and the density and velocity of the supersonic flow.

A front shock, a contact discontinuity, and a backward shock should appear, and their velocities are designated as V_fs, V_cd, and V_bs, respectively. The strong shock condition is applied to the shocks and the specific heat ratio is set to be 5/3. It is also assumed that the regions behind the front shock and the backward shock have the uniform pressure P_cd and the uniform velocity of V_cd, sandwiching the contact discontinuity surface.

Then, from the equations for the mass continuity and the momentum continuity across the shock surface, we get

$$\begin{equation} V_{\rm fs} = \frac{4}{3} V_{\rm cd}, \end{equation}$$

(A6)

and

$$\begin{equation} P_{\rm cd} = \frac{4}{9} \rho _{\rm st} V_{\rm cd}^{2}, \end{equation}$$

(A7)

for the front shock, and

$$\begin{equation} V_{\rm bs} = \frac{4}{3} V_{\rm cd} - \frac{1}{3} v_{\rm sf}, \end{equation}$$

(A8)

and

$$\begin{equation} P_{\rm cd} = \frac{4}{3} \rho _{\rm sf, bs} (v_{\rm sf}-V_{\rm cd})^{2}, \end{equation}$$

(A9)

for the backward shock. Here, ρ_sf,bs is the density of the supersonic flow on the upstream side to the backward shock. Equating equations (A7) and (A9), we obtain

$$\begin{equation} \frac{4}{9} \rho _{\rm st} V_{\rm cd}^{2} = \frac{4}{3} \rho _{\rm sf} (v_{\rm sf}-V_{\rm cd})^{2}. \end{equation}$$

(A10)

From equations (A6), (A8), and (A10), V_fs, V_cd, and V_bs can be determined for given ρ_st, ρ_sf,bs, and v_sf.

Next, we consider a case in which the jet plunges into the SNR matter. In this case, differently from the first case, the stationary matter should not accumulate in front of the jet but flow out to the side of the jet in the jet rest frame. As a result, the pressure, P_cd, loaded to the jet matter by the stationary matter should be significantly reduced from the value in equation (A7), and can be expressed as

$$\begin{equation} P_{\rm cd} = \xi \frac{4}{9} \rho _{\rm st} V_{\rm cd}^{2}, \end{equation}$$

(A11)

where ξ is the reduction factor. Simultaneously, V_fs is considered to equal V_cd roughly. Equations (A6) and (A9) hold even in this case.

Now, equation (A10) can be generalized for both the above cases as

$$\begin{equation} \xi \frac{4}{9} \rho _{\rm st} V_{\rm cd}^{2} = \frac{4}{3} \rho _{\rm sf, bs} (v_{\rm sf}-V_{\rm cd})^{2}, \end{equation}$$

(A12)

where ξ is 1 in the first case and is significantly smaller the unity in the second case. This equation can be rewritten as

$$\begin{equation} x^{2} - \frac{2A}{A-1}x+\frac{A}{A-1}=0, \end{equation}$$

(A13)

where

$$\begin{equation} x=\frac{V_{\rm cd}}{v_{\rm sf}}, \end{equation}$$

(A14)

and

$$\begin{equation} A = \frac{3 \rho _{\rm sf, bs}}{\xi \rho _{\rm st}}. \end{equation}$$

(A15)

The solution of this quadratic equation for x to pass smoothly through the point of x = 1/2 (the solution for A = 1) is

$$\begin{equation} x = \frac{\sqrt{A}}{\sqrt{A}+1}. \end{equation}$$

(A16)

Since the supersonic flow is the jet or the disk wind here, the density of the supersonic flow at a distance r from the center can be approximated to be proportional to r⁻². On the other hand, the stationary matter is regarded to be the SNR matter here and the density can be set to be constant independently of the place. Hence, we can consider that A ∝ r⁻². When r is sufficiently small, A is much larger than unity and then x ≃ 1, i.e., V_cd ≃ v_sf. However, if the front of the supersonic flow advances so much as for A to get as small as or smaller than unity, x comes to decrease from unity and the front of the sonic flow is decelerated. Finally, the quasi-steady state keeping A constant is likely to be realized when the position of the backward shock stops, i.e., V_bs = 0. Then, from equation (A6) and V_bs = 0, we get V_cd = v_sf/4, i.e., x = 1/4. We see from equation (A16) that A = 1/9 when x = 1/4.

References

Abramovicz

M. A.

,

Czerny

B.

,

Lasota

J. P.

,

Szuszkiewicz

E.

1988

,

ApJ

,

332

,

646

Crossref

Search ADS

Blundell

K. M.

,

Mioduszewski

A. J.

,

Muxlow

T. W. B.

,

Podsiadlowski

P.

,

Rupen

M. P.

2001

,

ApJ

,

562

,

L79

Crossref

Search ADS

Brinkmann

W.

,

Kawai

N.

2000

,

A&A

,

363

,

640

Cherepashchuk

A. M.

1981

,

MNRAS

,

194

,

761

Crossref

Search ADS

Cherepashchuk

A. M.

,

Postnov

K. A.

,

Belinski

A. A.

2018

,

MNRAS

,

479

,

4844

Crossref

Search ADS

Cherepashchuk

A. M.

,

Sunyaev

R. A.

,

Molkov

S. V.

,

Antokhina

E. A.

,

Postnov

K. A.

,

Bogomazov

A. I.

2013

,

MNRAS

,

436

,

2004

Crossref

Search ADS

Cherepashchuk

A. M.

,

Sunyaev

R. A.

,

Postnov

K. A.

,

Antokhina

E. A.

,

Molkov

S. V.

2009

,

MNRAS

,

397

,

479

Crossref

Search ADS

Cherepashchuk

A. M.

,

Sunyaev

R. A.

,

Sefina

E. V.

,

Panchenko

I. E.

,

Molkov

S. V.

,

Postnov

K. A.

2003

,

MNRAS

,

411

,

L441

Crampton

D.

,

Cowley

A. P.

,

Hutchings

J. B.

1980

,

ApJ

,

235

,

L131

Crossref

Search ADS

Dubner

G. M.

,

Holdaway

M.

,

Goss

W. M.

,

Mirabel

I. F.

1998

,

ApJ

,

116

,

1842

Crossref

Search ADS

Eggelton

P. P.

1983

,

ApJ

,

268

,

368

Crossref

Search ADS

Fabrika

S.

2004

,

Astrophys. Space Phys. Rev.

,

12

,

1

Frank

J.

,

King

A.

,

Raine

D.

2002

,

Accretion Power in Astrophysics

(

Cambridge

:

Cambridge University Press

)

Goodall

P. T.

,

Blundell

K. M.

,

Bell Burnell

S. J.

2011

,

MNRAS

,

414

,

2828

Crossref

Search ADS

Guerriero

R.

et al.

1999

,

MNRAS

,

307

,

179

Crossref

Search ADS

Inoue

H.

2012

,

PASJ

,

64

,

40

Crossref

Search ADS

Inoue

H.

2019

,

PASJ

,

71

,

36

Crossref

Search ADS

Inoue

H.

2021a

,

PASJ

,

73

,

795

Crossref

Search ADS

Inoue

H.

2021b

,

PASJ

,

73

,

1429

Crossref

Search ADS

Inoue

H.

2022a

,

PASJ

,

74

,

R1

Crossref

Search ADS

Inoue

H.

2022b

,

PASJ

,

submitted

Katz

J. I.

1980

,

ApJ

,

236

,

L127

Crossref

Search ADS

Katz

J. I.

,

Anderson

S. F.

,

Margon

B.

,

Grandi

S. A.

1982

,

ApJ

,

260

,

780

Crossref

Search ADS

Kotani

T.

,

Kawai

N.

,

Matsuoka

M.

,

Brinkmann

W.

1996

,

PASJ

,

48

,

619

Crossref

Search ADS

Kubota

A.

,

Makishima

K.

2004

,

ApJ

,

601

,

428

Crossref

Search ADS

Margon

B.

1984

,

ARA&A

,

22

,

507

Crossref

Search ADS

Migliari

S.

,

Fender

R.

,

Mendez

M.

2002

,

Science

,

297

,

1673

Crossref

Search ADS

PubMed

Morrison

R.

,

McCammon

D.

1983

,

ApJ

,

278

,

119

Narayan

R.

,

Yi

I.

1994

,

ApJ

,

428

,

L13

Crossref

Search ADS

Paragi

Z.

,

Vermeulen

R. C.

,

Fejes

I.

,

Schilizzi

R. T.

,

Spencer

R. E.

,

Stirling

A. M.

1999

,

A&A

,

348

,

910

Shakura

N. L.

,

Sunyaev

R. A.

1973

,

A&A

,

24

,

337

van den Heuvel

E. P. J.

,

Ostriker

J. P.

,

Petterson

J. A.

1980

,

A&A

,

81

,

L7

Watson

M. G.

,

Willingale

R.

,

Grindlay

J. E.

,

Seward

F. D.

1983

,

ApJ

,

273

,

688

Crossref

Search ADS

Whitemire

D. P.

,

Matese

J. J.

1980

,

MNRAS

,

193

,

707

Crossref

Search ADS

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Article Contents

Mass flows and their behaviors in the SS 433–W 50 system

Abstract

1 Introduction

2 Model situation

3 Interpretations of observations

3.1 Precession of the accretion ring

3.2 Hard X-ray emission

3.3 Jet ejection

3.4 Obscuration of the central engine

3.5 Interaction between the jet and the disk wind

3.6 Interaction of the disk wind with the SNR matter

3.7 Interaction of the jets with the SNR matter

4 Summary and discussion

Appendix 1 Shape of the disk wind edge

Appendix 2 Interaction of supersonic flow with the ambient matter

References

Citations

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Article Contents

Mass flows and their behaviors in the SS 433–W 50 system

Abstract

1 Introduction

2 Model situation

3 Interpretations of observations

3.1 Precession of the accretion ring

3.2 Hard X-ray emission

3.3 Jet ejection

3.4 Obscuration of the central engine

3.5 Interaction between the jet and the disk wind

3.6 Interaction of the disk wind with the SNR matter

3.7 Interaction of the jets with the SNR matter

4 Summary and discussion

Appendix 1 Shape of the disk wind edge

Appendix 2 Interaction of supersonic flow with the ambient matter

References

Citations

Views

Altmetric

Email alerts

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