1. Introduction
For nearly a century, scientists have been investigating the mechanisms that cause space weather events such as magnetic storms (e.g. Chapman & Ferraro Reference Chapman and Ferraro1940; Ferraro Reference Ferraro1952). Magnetic reconnection is identified as one of the most important processes in space plasmas that drives various explosive phenomena, such as solar flares (e.g. Masuda et al. Reference Masuda, Kosugi, Hara, Tsuneta and Ogawara1994), coronal mass ejections (e.g. Gosling, Birn & Hesse Reference Gosling, Birn and Hesse1995) and the magnetic substorms in the Earth's magnetotail (e.g. Angelopoulos et al. Reference Angelopoulos, McFadden, Larson, Carlson, Mende, Frey, Phan, Sibeck, Glassmeier and Auster2008). It efficiently converts the magnetic energy in a current sheet to the kinetic and internal energies of the plasma through reconfiguration of the topology of the magnetic field and connectivity of the magnetic field lines. In the laboratory, magnetic reconnection destabilizes the plasma (e.g. Yamada et al. Reference Yamada, Levinton, Pomphrey, Budny, Manickam and Nagayama1994; Dorfman et al. Reference Dorfman, Ji, Yamada, Yoo, Lawrence, Myers and Tharp2013) and is fatal to stable controlled fusion. Understanding of how the reconnection triggers and evolves is crucial for a complete description of how energy is stored and released in different plasma environments.
As the plasma is a multiscale system, various types of waves and instabilities exist on the largest magnetohydrodynamic (MHD) scale, the intermediate ion-kinetic scale and the smallest electron-kinetic scale. Consequently, magnetic reconnection is a multiscale process. In the past two decades, numerical simulations (e.g. Daughton, Scudder & Karimabadi Reference Daughton, Scudder and Karimabadi2006; Guo et al. Reference Guo, Liu, Daughton and Li2015; Cassak et al. Reference Cassak, Genestreti, Burch, Phan, Shay, Swisdak, Drake, Price, Eriksson and Ergun2017; Lu et al. Reference Lu, Angelopoulos, Artemyev, Pritchett, Liu, Runov, Tenerani, Shi and Velli2019) as well as in situ measurements of space plasma (e.g. Burch et al. Reference Burch, Torbert, Phan, Chen, Moore, Ergun, Eastwood, Gershman, Cassak and Argall2016; Torbert et al. Reference Torbert, Burch, Phan, Hesse, Argall, Shuster, Ergun, Alm, Nakamura and Genestreti2018) have greatly enhanced our understanding of the microscopic kinetic physics of reconnection in the ion diffusion region and electron diffusion region. But it still remains unclear as how fast reconnection triggers in the macroscopic current sheets whose dimensions are much larger than any ion kinetic scales, e.g. current sheets in the preflare configurations in the solar corona.
In an ideal-MHD regime, the magnetic field lines are ‘frozen in’ the plasma and change of their connectivity is prohibited. Thus, a certain mechanism that breaks the ideal-MHD condition must play a role for the reconnection to happen. In macroscopic current sheets, this mechanism is either collision-induced resistivity, or some kind of effective resistivity caused by microscopic wave–particle interactions (e.g. Büchner & Elkina Reference Büchner and Elkina2006; Ma et al. Reference Ma, Chen, Zhang and Yu2018). Hence, the triggering problem of reconnection at MHD scales is essentially the stability problem of the resistive current sheet. Since the 1960s, many works have been conducted on the resistive instability, i.e. the so-called ‘tearing instability’, of the current sheet (Furth, Killeen & Rosenbluth Reference Furth, Killeen and Rosenbluth1963; Coppi, Greene & Johnson Reference Coppi, Greene and Johnson1966). The tearing mode grows with the help of resistivity that transfers the magnetic energy stored in the shear magnetic field to the growing perturbations, leading to the formation of a chain of plasmoids. Considering an infinitely long current sheet with thickness $a$
, one can define the dimensionless Lundquist number $S= aV_A/\eta$
; where $V_A = B/\sqrt {\mu _0 \rho }$
is the characteristic Alfvén speed, with $\rho$
and $B$
being the plasma density and asymptotic magnetic field strength, and $\eta$
is the magnetic diffusivity, which is resistivity divided by the permeability $\mu _0$
. For simplicity, we will refer to $\eta$
as ‘resistivity’ hereinafter. Linear theory predicts that the most unstable tearing mode has a growth rate $\gamma \tau _a \sim S^{-1/2}$
where $\tau _a = a/V_A$
is the Alfvén crossing time. This relation implies a faster growth of the instability with larger resistivity. As plasma in most of the space environments and laboratories is weakly collisional (Ji & Daughton Reference Ji and Daughton2011; Pucci, Velli & Tenerani Reference Pucci, Velli and Tenerani2017), the tearing mode seems to grow at a very slow speed.
However, considering a two-dimensional (2-D) current sheet whose length is $L$
, in most prereconnection configurations, its aspect ratio can be very large ($L \gg a$
). In this case, we should use $L$
instead of $a$
to measure the growth rate of a tearing instability. After redefining the Alfvén crossing time and Lundquist number such that $\tau _L = L/V_A$
and $S_L = LV_A/\eta$
, it can be shown that the maximum growth rate of tearing mode is $\gamma \tau _L \sim S_L^{-1/2} \times (a/L)^{-3/2}$
(Pucci & Velli Reference Pucci and Velli2013), implying that the aspect ratio of the current sheet is an important factor in determining how fast the mode grows. The growth rate can be extremely large at the low-resistivity limit ($S_L \rightarrow \infty$
) if the current sheet is thinner than a critical value $a/L \sim S_L^{1/3}$
. Especially, in the classic model for steady reconnection with resistivity, i.e. the Sweet–Parker type current sheet, whose aspect ratio is $a/L\sim S_L^{-1/2}$
(Parker Reference Parker1957; Sweet Reference Sweet1958), the maximum growth rate of tearing is $\gamma \tau _L \sim S_L^{1/4}$
(Loureiro, Schekochihin & Cowley Reference Loureiro, Schekochihin and Cowley2007; Tajima & Shibata Reference Tajima and Shibata2018). The positive power-law index means that the growth rate can be extremely large in the limit $S_L \rightarrow \infty$
(Bhattacharjee et al. Reference Bhattacharjee, Huang, Yang and Rogers2009; Huang & Bhattacharjee Reference Huang and Bhattacharjee2013).
The above analysis leads to a plausible scenario of the explosive energy release of the macroscopic current sheet. Initially, the current sheet is thick with $(a/L) > S_L^{-1/3}$
, and thus is stable to a tearing mode. Then some external forces gradually build up magnetic energy and result in thinning of the current sheet. Once the current sheet thins to the critical aspect ratio $(a/L) \sim S_L^{-1/3}$
, the growth of the tearing mode suddenly becomes very fast, and the current sheet breaks up into many plasmoids and smaller-scale current sheets. This process can happen recursively in the newly formed current sheets and dissipates the magnetic energy rapidly (Shibata & Tanuma Reference Shibata and Tanuma2001; Landi et al. Reference Landi, Del Zanna, Papini, Pucci and Velli2015; Tenerani et al. Reference Tenerani, Velli, Rappazzo and Pucci2015b; Papini, Landi & Del Zanna Reference Papini, Landi and Del Zanna2019), until it is terminated due to the decrease of the Lundquist number (Shi, Velli & Tenerani Reference Shi, Velli and Tenerani2018) or the ion kinetic effect (Shi et al. Reference Shi, Tenerani, Velli and Lu2019; Bora, Bhattacharyya & Smolarkiewicz Reference Bora, Bhattacharyya and Smolarkiewicz2021). Thus, tearing instability is an important and fundamental mechanism that facilitates fast reconnection in the large-scale current sheets. Consequently, it is important to thoroughly study it under different configurations. Recent progress on this topic includes the calculations of its linear growth rate with viscosity (Tenerani et al. Reference Tenerani, Rappazzo, Velli and Pucci2015a), different background magnetic field profiles (Pucci et al. Reference Pucci, Velli, Tenerani and Del Sarto2018), Hall effect (Pucci et al. Reference Pucci, Velli and Tenerani2017), guide field (Shi et al. Reference Shi, Velli, Pucci, Tenerani and Innocenti2020), ion-neutral collisions (Pucci et al. Reference Pucci, Singh, Tenerani and Velli2020) and normal component of magnetic field (Shi et al. Reference Shi, Artemyev, Velli and Tenerani2021).
In space plasma, current sheets are frequently accompanied by plasma flows. In the dayside magnetosheath, reconnection events are often observed within highly turbulent plasma (e.g. Huang et al. Reference Huang, Sahraoui, Retinò, Le Contel, Yuan, Chasapis, Aunai, Breuillard, Deng and Zhou2016), and also for the solar wind (e.g. Osman et al. Reference Osman, Matthaeus, Gosling, Greco, Servidio, Hnat, Chapman and Phan2014). As a result, the reconnecting current sheets are likely to be affected by plasma flows of all directions. Plasma flows are also detected in the preflare corona (e.g. Wallace et al. Reference Wallace, Harra, van Driel-Gesztelyi, Green and Matthews2010) and nightside magnetotail current sheet (e.g. Lane et al. Reference Lane, Grocott, Case and Walach2021). At the tip of the helmet streamer where the heliospheric current sheet forms, growth of the tearing instability accompanied by an outward propagating solar wind stream is observed in MHD simulations (Réville et al. Reference Réville, Velli, Rouillard, Lavraud, Tenerani, Shi and Strugarek2020, Reference Réville, Fargette, Rouillard, Lavraud, Velli, Strugarek, Parenti, Brun, Shi and Kouloumvakos2022). Thus, study of how the tearing mode instability is modified by plasma flows is necessary. Many works have been conducted on the effect of a shear flow, i.e. flow parallel to the shear magnetic field (e.g. Hofman Reference Hofman1975; Paris & Sy Reference Paris and Sy1983; Einaudi & Rubini Reference Einaudi and Rubini1986; Chen & Morrison Reference Chen and Morrison1990; Ofman et al. Reference Ofman, Chen, Morrison and Steinolfson1991; Paris, Wood & Stewart Reference Paris, Wood and Stewart1993; Chen, Otto & Lee Reference Chen, Otto and Lee1997; Dahlburg, Boncinelli & Einaudi Reference Dahlburg, Boncinelli and Einaudi1997; Faganello et al. Reference Faganello, Pegoraro, Califano and Marradi2010). For example, Chen et al. (Reference Chen, Otto and Lee1997), through 2-D MHD simulations, show that a sub-Alfvénic shear flow stabilizes the tearing mode, and with a super-Alfvénic shear flow the instability is dominated by the Kelvin–Helmholtz mode. This result is confirmed by a recent work (Shi et al. Reference Shi, Artemyev, Velli and Tenerani2021) that calculates the linear instability growth rate by a boundary value problem (BVP) approach. They also show that when the flow is exactly Alfvénic, the current sheet is extremely stable and the perturbation only grows at the rate of diffusion. Compared with the shear flow, the case where a plasma jet exists at the centre of the current sheet is more complicated, because the jet itself is susceptible to two types of streaming instabilities, i.e. the sausage (varicose) mode and the kink (sinuous) mode. Wang, Lee & Wei (Reference Wang, Lee and Wei1988a), using an initial value solver of compressible MHD equations, show that a super-Alfvénic plasma jet can increase the growth rate of the tearing mode. Subsequent works (Lee et al. Reference Lee, Wang, Wei and Tsurutani1988; Wang et al. Reference Wang, Lee, Wei and Akasofu1988b) show that under this type of configuration both the kink mode and sausage mode exist and the sausage mode mixes with the tearing mode in the presence of resistivity. Two-dimensional MHD (Bettarini et al. Reference Bettarini, Landi, Rappazzo, Velli and Opher2006) and Hall–MHD simulations (Hoshino & Higashimori Reference Hoshino and Higashimori2015) confirm these early results.
In this study, we carry out a comprehensive investigation of the stability problem of the current sheet with a plasma jet in the framework of linear incompressible MHD, using an eigenvalue problem solver. We examine both the streaming sausage mode and streaming kink mode with and without resistivity. We derive the controlling equation set under a generalized configuration such that the plasma jet can have arbitrary angle with respect to the magnetic field and a finite guide field is allowed. The paper is organized as follows. In § 2, we describe the background fields used in this study. In § 3, we derive the equation set for the perturbation field. In § 4 we present the detailed results of our calculation. In § 5 we summarize the results and discuss the possible applications of the results to space plasma.
2. Equilibrium and background fields
We start from the resistive-MHD equation set
where $\rho,\boldsymbol{V},\boldsymbol{B},P$
are density, velocity, magnetic field and pressure; $\kappa$
is the adiabatic index; $S$
is the Lundquist number. Incompressibility is assumed throughout the study, i.e. $\rho (t,\boldsymbol{x}) \equiv \rho _0$
is a constant. In a generalized configuration, the background magnetic field and velocity consist of both $x$
and $z$
components but are functions of $y$
only:
Consequently, the momentum convection term and the magnetic tension force are both zero: $\boldsymbol{V}\boldsymbol {\cdot } \boldsymbol {\nabla } \boldsymbol{V} \equiv 0, \boldsymbol{B}\boldsymbol {\cdot } \boldsymbol {\nabla } \boldsymbol{B} \equiv 0$
, and the divergences of $\boldsymbol{V}$
and $\boldsymbol{B}$
are also zero. The zeroth-order scalar pressure ensures the pressure balance
where $P^T$
is the uniform total pressure. With the above configuration, the background field is in equilibrium without resistivity, but will diffuse with a finite resistivity. However, in most of the space and laboratory plasmas, the resistivity is extremely small, thus the diffusion time is much longer than the growth time of instabilities of interest. Therefore, we are able to neglect the diffusion of background fields.
In this study, we adopt the Harris-type current sheet model for the magnetic field such that $B_x(y)=B_0 \tanh (y/a)$
. We also allow a uniform guide field $B_z(y) = B_g$
. We assume the velocity is of the following form:
i.e. the jet rotates from the $x$
direction by an angle $\alpha$
(figure 1a). We adopt a flow function,
where $d$
is the half-thickness of the jet, and $V_0$
is the flow speed at the centre of the current sheet ($y=0$
) and is a variable parameter. Throughout the study, we fix $d=a$
, i.e. the width of the jet is the same as the width of the current sheet. The profile of $V(y)$
is plotted in figure 1(b) together with the profile of the $x$
-component of the magnetic field $B(y)$
.
Figure 1. (a) Coordinate systems used in this study and the background fields. Coordinate system $\tilde {x}$
–$\tilde {z}$
is $x$
–$z$
rotated by an angle $\theta$
with respect to $y$
axis, so that $\tilde {x}$
is parallel to the wavevector $\boldsymbol{k}$
. (b) The $y$
-profiles of the background flow $V(y)$
(blue solid) and the $x$
-component of the magnetic field $B(y)$
(orange dashed) used in this study.
3. Equation set for the perturbations
For the perturbations, we consider a Fourier mode whose growth rate $\gamma$
is a complex number and wavevector $\boldsymbol{k}$
is in the $x$
–$z$
plane with an arbitrary angle $\theta$
with respect to the $x$
direction (figure 1a):
Hence, the perturbation fields have the form
For simplicity, one can rotate the $x$
–$z$
coordinates with respect to the $y$
-axis by angle $\theta$
and get a new coordinate system $\tilde {x}\unicode{x2013}\tilde {z}$
(figure 1a) such that $\boldsymbol{k} = k \hat {e}_{\tilde {x}}$
, and thus there is $\partial _{\tilde {z}}\equiv 0$
. The background magnetic field and velocity can then be written as $\boldsymbol{B} = B_{\tilde {x}}(y) \hat {e}_{\tilde {x}} + B_{\tilde {z}}(y) \hat {e}_{\tilde {z}}$
and $\boldsymbol{V} = V_{\tilde {x}}(y) \hat {e}_{\tilde {x}} + V_{\tilde {z}}(y) \hat {e}_{\tilde {z}}$
after projection to the new coordinate system.
The next step is to derive a closed linear equation set for the eigenvalue problem. We start from the linearized momentum equation (with uniform density $\rho$
)
where we have normalized the magnetic field by $\sqrt {\mu _0\rho }$
so that it is in the units of speed. To get rid of the first-order pressure, we can take the curl of the equation and get
Using $\boldsymbol {\nabla } \boldsymbol {\cdot } \boldsymbol{u} = 0$
and $\boldsymbol {\nabla } \boldsymbol {\cdot } \boldsymbol{b} = 0$
to eliminate $u_{\tilde {x}}$
and $b_{\tilde {x}}$
, the $\tilde {z}$
component of the above equation can be rearranged in the following form:
where the prime indicates $\partial _y$
. We note that the above equation contains only the $y$
-component of $\boldsymbol{u}$
and $\boldsymbol{b}$
. The linearized induction equation is
where $S = a V_{A}/\eta$
is defined with the half-thickness of the current sheet $a$
and the upstream Alfvén speed $V_A=B_0/\sqrt {\mu _0 \rho }$
. The $y$
-component of the linearized induction equation is
which also contains only the $y$
-component of $\boldsymbol{u}$
and $\boldsymbol{b}$
. Thus, (3.5) and (3.7) form a closed equation set for $u_y$
and $b_y$
:
Here we have assimilated a ${\rm \pi} /2$
phase difference between $u_y$
and $b_y$
, i.e. we have replaced $ib_y$
with $b_y$
.
One can immediately find that, if $V_{\tilde {x}} = 0$
, i.e. if $\boldsymbol{k} \boldsymbol {\cdot } \boldsymbol{V} = 0$
, the system is purely determined by $B_{\tilde {x}}$
, similar to the classic tearing case, and the growth rate is purely real. As an example, consider an antiparallel magnetic field $\boldsymbol{B}=B_x(y) \hat {e}_x$
and an out-of-plane flow $\boldsymbol{V}=V_z(\,y) \hat {e}_z$
. In this case, if $\boldsymbol{k}=k \hat {e}_x$
, the system reduces to the classic tearing case, i.e. the flow has no effect on the solution. But if the wavevector is not along $x$
and has a finite $z$
-component, the flow will alter the growth rate and introduce an oscillation (a non-zero frequency) to the solution. Similarly, if $\boldsymbol{k} \boldsymbol {\cdot } \boldsymbol{B} = 0$
, the system is determined purely by the flow $V_{\tilde {x}}$
and there are only stream-induced instabilities.
As a final remark, we note that (3.8) is in generalized form, and works for any functions $\boldsymbol{V}(\,y)$
and $\boldsymbol{B}(\,y)$
once they have the form of (2.2a,b) and the system is incompressible.
4. Results
Equation (3.8) is a BVP with the boundary conditions $u_y(y\rightarrow \pm \infty )=0, b_y(y\rightarrow \pm \infty )=0$
. Far from the centre of the current sheet ($y\rightarrow \pm \infty$
), the derivatives of the background fields reduce to zero. Consequently, one can see that $u_y,b_y \propto \exp (-k|y|)$
satisfy (3.8a), and (3.8b) just gives the ratio $u_y/b_y$
at $y\rightarrow \pm \infty$
. In this study, we use the numerical BVP solver implemented in the Python package SciPy (Virtanen et al. Reference Virtanen, Gommers, Oliphant, Haberland, Reddy, Cournapeau, Burovski, Peterson, Weckesser and Bright2020) to solve (3.8). In practice, we set the boundaries at $y=\pm 15a$
, which are large enough to acquire accurate solutions.
Before solving the equation set, we need to define the parities of $u_y$
and $b_y$
first. Given that $V_{\tilde {x}}(y)$
is an even function and $B_{\tilde {x}}(y)$
is an odd function, one can see from (3.8) that there are two possible combinations of the parities of $u_y$
and $b_y$
: (1) $u_y(y)$
is odd and $b_{y}(y)$
is even; (2) $u_y(y)$
is even and $b_{y}(y)$
is odd. The first case is the so-called ‘sausage’ (varicose) mode, which leads to the formation of a chain of blobs and plasmoids (if resistivity is non-zero). Tearing instability is categorized to the sausage mode. The second case is the ‘kink’ (sinuous) mode, which results in wavy distortion of the current sheet and stream. In solving the problem, we carefully search for both of the two modes.
4.1. Instabilities of a plasma jet inside a non-resistive current sheet
In this section, we analyse the simplest 2-D case, where the jet, magnetic field and wavevector are all aligned along the $x$
direction, i.e. there is no out-of-plane flow or guide field. Besides, the current sheet is non-resistive. We will show how the pure streaming instabilities are modified by the interaction between the jet and the current sheet.
In figure 2(a,b), we show the dispersion relations $\gamma (k)$
and $\omega (k)$
for the sausage mode (solid lines) and kink mode (dashed lines). The wavenumber is normalized to $d$
, and $\gamma$
and $\omega$
are normalized to $V_0/d$
. The lines are colour-coded with the $B_0/V_0$
ratio, as marked in figure 2(a). Both modes have larger growth rate with smaller magnetic field, as it is well known that the magnetic field suppresses the stream-shear instability. The kink mode has larger growth rate than the sausage mode for $B_0/V_0 \lesssim 0.4$
. But when $B_0/V_0 \gtrsim 0.4$
, the sausage mode becomes more unstable than the kink mode, i.e. the magnetic field stabilizes the kink mode more effectively. The wavenumber of the most unstable kink mode is roughly twice the wavenumber of the most unstable sausage mode. In figure 2(c,d), we show 2-D profiles of $V_x$
and $J_z$
, respectively, for the sausage mode with $B_0/V_0=0.4$
and $kd=0.46$
, which is the fastest growing mode for $B_0/V_0=0.4$
. Here, the 2-D profiles are calculated by summing the background fields and the solved eigenfunctions of the perturbations. The solid curves in figure 2(c,d) are the streamlines and the magnetic field lines, respectively, and one can see the formation of a chain of sausage-like structures in both the velocity and the magnetic field. The dashed lines in the two figure panels mark the resonance surfaces where $\omega = k V(y)$
and the fields undergo sharp transitions. Figure 2(e, f) are similar to figure 2(c,d) but for kink mode with $B_0/V_0=0.4$
and $kd=0.83$
, which is the fastest growing mode for $B_0/V_0=0.4$
. We can see the growth of kink mode deforms the jet and the magnetic field in a sinuous way.
Figure 2. Streaming instabilities of a plasma jet inside a current sheet. The jet and the wave vector are both parallel to the magnetic field, and there is no guide field. (a,b) Growth rate ($\gamma$
) and oscillation frequency ($\omega$
) as functions of wavenumber $k$
for the sausage mode (solid lines) and the kink mode (dashed lines) with different magnetic field and jet speed ratios $B_0/V_0$
. Black lines represent the non-magneto fluid case ($B_0/V_0 = 0$
). Here, the wavenumber is normalized by the half-thickness of the jet, which is equal to the half-thickness of the current sheet, and the growth rate and frequency are normalized to $d/V_0$
. Panels (c) and (d) are plotted based on the sausage mode with $B_0/V_0=0.4$
and $kd=0.46$
(the fastest growing mode). (c) Two-dimensional profiles of $V_x$
, solid lines are the streamlines, and the two dashed lines mark the resonance surfaces where $\omega = k V(y)$
. (d) 2-D profiles of $J_z$
(out-of-plane current density), solid lines are the magnetic field lines, and the two dashed lines mark the resonance surfaces. Panels (e) and ( f) are similar to panels (c) and (d) but for the kink mode with $B_0/V_0=0.4$
and $kd=0.83$
(the fastest growing mode). We note that in panels (c)–( f) all the physical quantities are the sums of the linear eigenfunctions and the background fields.
Here we must point out that, in the non-resistive case, growth of the streaming sausage mode does not result in reconnection of the magnetic field since the system is ideal-MHD. The sausage-like structures are not plasmoids but are merely antiphased deformation of the magnetic field lines on the two sides of the current sheet (figure 2d). Actually, from (3.8b), one can see that $b_y(y=0) = 0$
in the limit $S\rightarrow \infty$
.
4.2. Instabilities of a plasma jet inside a resistive current sheet
Based on the results of the previous section, we explore how a finite resistivity will modify the streaming instabilities and how the tearing mode couples with the streaming sausage mode.
Figure 3(a,b) show the dispersion relation $\gamma (k)$
and $\omega (k)$
for the sausage mode (solid lines) and kink mode (dashed lines) with $V_0/B_0=2.5$
and different Lundquist numbers. Different from figure 2, hereinafter we normalize $\gamma$
and $\omega$
by the Alfvén crossing time $a/V_A$
, i.e. $a/B_0$
, which is conventional for the analysis of the tearing mode. Although not very significantly at this $V_0/B_0$
ratio, resistivity enhances the maximum growth rates of both modes. It is clear that the kink mode is less affected by resistivity compared with the sausage mode. This is a reasonable result because the streaming sausage mode is expected to couple with the tearing mode whose growth rate is determined by the resistivity. To better illustrate this point, figure 3(c) shows the maximum growth rates max($\gamma$
) as functions of the ratio $V_0/B_0$
for the sausage and kink modes. The blue line is the sausage mode with $S=1000$
, and the orange line is the kink mode with $S=1000$
. The two black lines are the cases without resistivity ($S\rightarrow \infty$
). As already shown in § 4.1, in the non-resistive case, the kink mode (black squares) is more unstable than the sausage mode (black circles) with large $V_0/B_0$
but its growth rate decreases fast towards zero as $V_0/B_0$
approaches $\sim 2$
from above. The growth rate of sausage mode decreases to zero at $V_0/B_0 \sim 1$
. With $S=1000$
, the maximum growth rate of the kink mode almost does not change from the non-resistive case. For the sausage mode, the growth rate does not change for $V_0/B_0 \geq 2$
but becomes larger than the non-resistive case for $V_0/B_0 <2$
. Especially, even for $V_0/B_0 < 1$
, its maximum growth rate is larger than zero. This is because of the transition of the streaming sausage mode to the tearing mode at small $V_0/B_0$
ratio. Thus, in the finite (but not too large) resistivity case, we can roughly divide the sausage mode into two regimes according to $V_0/B_0$
. For $V_0/B_0 \gtrsim 2$
, the mode is almost dominated by the jet, hence it is the ‘streaming sausage’ mode. For $V_0/B_0 \lesssim 2$
, especially for $V_0/B_0 \lesssim 1$
, the mode becomes heavily affected by the resistivity, thus it is the ‘streaming tearing’ mode. A similar result was obtained by Wang et al. (Reference Wang, Lee and Wei1988a), who showed that there is a sharp increase of the maximum growth rate of the sausage mode at $V_0/B_0 \approx 1.2$
, implying a transition of the tearing mode to streaming mode.
Figure 3. (a,b) Dispersion relation $\gamma (k)$
and $\omega (k)$
for sausage mode (solid lines) and kink mode (dashed lines) with $V_0/B_0=2.5$
. Colours of the curves correspond to the Lundquist numbers, such that yellow is $S=10$
, light purple is $S=100$
and dark purple is $S=1000$
. Black curves are non-resistive cases ($S\rightarrow \infty$
). Different from figure 2, here $\gamma$
and $\omega$
are normalized by the Alfvén crossing time $a/V_A$
(or $a/B_0$
). (c) Maximum growth rate max($\gamma (k)$
) as a function of $V_0/B_0$
. Blue and orange curves are the sausage and kink modes, respectively, with $S=1000$
. The two black dashed curves are the non-resistive cases ($S\rightarrow \infty$
).
Figure 4(a) shows how the maximum growth rate of the sausage mode scales with the Lundquist number for different $V_0/B_0$
, and figure 4(b) shows the corresponding wavenumbers. The black dashed and dotted lines in panel (a) show $\gamma \propto S^{-1/2}$
and $\gamma \propto S^{-0.4}$
for references, and those in panel (b) show $k \propto S^{-1/4}$
and $k\propto S^{-0.15}$
for references. Classic tearing mode theory (Furth et al. Reference Furth, Killeen and Rosenbluth1963; Coppi et al. Reference Coppi, Greene and Johnson1966) shows that the maximum growth rate and corresponding wavenumber have the scaling relations $\gamma \propto S^{-1/2}$
and $k \propto S^{-1/4}$
in the limit of large $S$
. This is confirmed by the result for $V_0/B_0 = 0$
in figure 4. As $V_0/B_0$
increases, both the maximum growth rate and the wavenumber increase, but the slopes of these lines remain unchanged for $V_0/B_0 <1$
, implying that the mode is still tearing-like. But at $V_0/B_0 = 1$
, the scaling relations change such that the two lines are less steep, meaning that the dependence of the instability on the resistivity becomes weaker. At $V_0/B_0=1.25$
, the two lines are flat for $S \geq 10^5$
, indicating that the instability becomes weakly dependent on resistivity and start to transit to pure streaming sausage mode.
Figure 4. (a) Maximum growth rate of the sausage mode as a function of the Lundquist number $S$
for different $V_0/B_0$
. The black dashed line shows $\gamma \propto S^{-1/2}$
and the black dotted line shows $\gamma \propto S^{-0.4}$
for reference. (b) Corresponding wavenumber of the most unstable mode as a function of $S$
. The black dashed line shows $k \propto S^{-1/4}$
and the black dotted line shows $k \propto S^{-0.15}$
for reference.
In figure 5, we show the solved eigenfunctions $u_y$
(figure 5a1–a3) and $b_y$
(figure 5b1–b3) for the most unstable sausage modes with $S=10^6$
. Figures 5(a1,b1), 5(a2,b2) and 5(a3,b3) are $V_0/B_0=$
0, 0.75 and 1.25, respectively. In each figure panel, the solid curve is the real part, and the dashed curve is the imaginary part of the function. Figure 5(a1) and 5(b1) correspond to the classic tearing case, where the solution shows a very thin singular layer, or ‘inner’ layer, around $y=0$
, which is dominated by the resistivity. Outside the singular layer, the solution is determined by the non-resistive parts of (3.8). As the ratio $V_0/B_0$
increases (figure 5a2–b2 and figure 5a3–b3), the outer solution is altered by the jet, while the inner singular layer persists, as can be seen in the embedded blow-up plots. In addition, as $V_0/B_0$
increases, the relative amplitude of $b_y$
and $u_y$
drops. For the pure tearing mode, $|b_y| \gg |u_y|$
, implying the mode is magnetic field dominated. As the mode transits to streaming sausage mode, magnetic field perturbation gradually weakens.
Figure 5. Eigenfunctions $u_y$
(a1–a3) and $b_y$
(b1–b3) for the most unstable sausage modes with $S=10^6$
and varying $V_0/B_0$
. Panels (a1,b1), (a2,b2) and (a3,b3) are $V_0/B_0=0, 0.75$
and 1.25, respectively. In each panel, the solid and dashed curves are the real and imaginary parts of the eigenfunctions. In panels (a2) and (b3), the embedded plots show the close-ups of the eigenfunctions.
4.3. Jet along the guide field and the oblique tearing mode
In this section, we only discuss the sausage mode because the kink mode is not affected much by the resistivity and not directly related to the reconnection process. In space environments or laboratory plasma, the jet is not necessarily parallel to the reconnecting magnetic field component, and it is possible that the jet has a finite component along the guide field direction (generally speaking, the $z$
direction no matter whether a guide field exists). In this case, the most unstable mode may be oblique whose wavevector also has a component along the guide field. In this section, we consider the case where the jet is along the $z$
axis. As the instability is determined fully by the background fields projected on the wavevector direction, one can imagine that, when we rotate $\boldsymbol{k}$
from the $x$
direction to the $z$
direction, the instability transits from pure tearing mode to pure streaming mode.
Figure 6(a) shows the maximum growth rate of the sausage mode as a function of $\theta (\boldsymbol{k},\boldsymbol {B_0})$
, which is the angle between $\boldsymbol{k}$
and $x$
-axis, for $S=10^4$
, $B_g=0$
, and different $V_0/B_0$
ratios. If there is no jet ($V_0/B_0 = 0$
), increasing $\theta$
leads to a monotonic decrease of the maximum growth rate. But as the jet speed increases, there is a turning point from which the maximum growth rate starts to rise. Even for a small ratio $V_0/B_0=0.1$
, a turning point exists at large $\theta$
($\approx 75^\circ$
). However, in this case the perpendicular ($\theta =90^\circ$
) mode still grows slower than the parallel ($\theta =0^\circ$
) mode. For large jet speed ($V_0/B_0=$
1.00), the curve is monotonically increasing. The turning point is due to the transition from the pure tearing mode to the pure streaming sausage mode as $\boldsymbol{k}$
rotates. Figure 6(b) shows the wavenumbers corresponding to the modes shown in figure 6(a). In general, as $\theta$
increases, the wavenumber also increases, especially for large $V_0/B_0$
ratios, because the most unstable streaming mode has larger wavenumber than the most unstable tearing mode. In figure 6(c), we plot the maximum growth rate of pure tearing mode as a function of $S$
in blue line with square markers. This curve corresponds to the $\theta =0$
case with $\boldsymbol {V_0} \parallel \hat {e}_z$
. In this figure panel, the horizontal dashed lines mark the maximum growth rate of the pure streaming sausage mode, i.e. the $\theta =90^\circ$
case, for different $V_0/B_0$
ratios. We note that the growth rates of the pure streaming modes are independent of $S$
because these modes decouple with the magnetic field. Consequently, the growth rate is simply proportional to $V_0/B_0$
. From this plot, we see that at any fixed Lundquist number $S$
, there is a critical $V_0/B_0$
above which the perpendicular mode (pure streaming sausage) has larger growth rate than the parallel mode (pure tearing). In figure 6(d) we plot this critical value $(V_0/B_0)_c$
as a function of $S$
. One can read that, for example, for $S=10^3$
, the critical value is $V_0/B_0\approx 0.4$
while for $S=5\times 10^5$
the critical value is $V_0/B_0\approx 0.02$
. For a $V_0/B_0$
that exceeds the critical value, the $\max (\gamma )\unicode{x2013}\theta$
curve (such as those shown in figure 6a) will raise with $\theta$
at some point and eventually reach a value at $\theta =90^\circ$
higher than that at $\theta = 0^\circ$
.
Figure 6. (a) Maximum growth rate of the sausage mode as a function of $\theta$
(angle between $\boldsymbol{k}$
and $\boldsymbol {B_0}$
), for $\alpha =90^\circ$
(angle between $\boldsymbol {V_0}$
and $\boldsymbol {B_0}$
), $S=10^4$
, $B_g=0$
and varying $V_0/B_0$
. (b) The corresponding wavenumbers. (c) Blue curve with square markers: maximum growth rate of pure tearing mode, i.e. $\theta = 0$
, as a function of $S$
. Horizontal dashed lines mark the maximum growth rate of the pure streaming sausage mode, i.e. $\theta =90^\circ$
, with varying $V_0/B_0$
. Note that the growth rates of pure streaming modes (modes decoupled from the magnetic field) are independent of $S$
. (d) Critical value $(V_0/B_0)_c$
, above which the pure streaming sausage mode has larger maximum growth rate than the pure tearing mode, as a function of $S$
.
Then we consider the case with a uniform guide field $B_g$
. In figure 7, each figure panel displays the maximum growth rate of the sausage mode as a function of the guide field strength $B_g/B_0$
for different $\theta$
. Figure 7(a)–(c) correspond to $V_0/B_0=$
0, 0.5 and 1.0, respectively. We note that $B_0$
is the asymptotic amplitude of the $x$
-component of the magnetic field. For small $\theta$
($\theta \lesssim 15^\circ$
), the maximum growth rate is not significantly modified by either $B_g$
or $V_0$
since $\boldsymbol{k}$
is quasi-perpendicular to the guide field direction. As $\theta$
increases, for small and intermediate jet speeds ($V_0/B_0 \leq 0.5$
), the maximum growth rate drops with the guide field. As already shown by Shi et al. (Reference Shi, Velli, Pucci, Tenerani and Innocenti2020), in the no-flow case, the guide field raises the growth rate only at large-$k$
(the so-called constant-$\psi$
) regime, but overall the maximum growth rate of the oblique mode ($\theta>0$
) decreases with an increasing guide field strength. However, figure 7(c) ($V_0/B_0 = 1$
) shows a very different result. For small guide field $B_g/B_0 \leq 0.5$
, the maximum growth rate increases with $\theta$
, similar to the result shown by figure 6, because the streaming mode of the jet has larger growth rate than the tearing mode. As $B_g/B_0$
continues to increase ($B_g/B_0=0.75$
and 1.0), $\max (\gamma )$
does not monotonically increase with $\theta$
but may start to drop with $\theta$
. Clearly, there is a competition between the jet and guide field. The jet tends to increase $\max (\gamma )$
as $\boldsymbol{k}$
rotates from the $x$
direction towards the $z$
direction, while the guide field overall plays a counter role but at certain $\theta$
it may raise $\max (\gamma )$
first before declining it (see curves for $\theta =30^\circ$
, $40^\circ$
and $45^\circ$
). Figure 8 displays the maximum growth rate of the sausage mode as a function of $\theta$
in the case $S=10^4$
, $V_0/B_0=1$
, and $\alpha =90^\circ$
, with different $B_g/B_0$
ratios. Without the guide field, max($\gamma$
) monotonically increases with $\theta$
as already shown in figure 6. But as $B_g/B_0$
increases, peaks appear in the $\max (\gamma )$
–$\theta$
curves, because the guide field effectively stabilizes both the oblique tearing mode and the streaming sausage mode, and hence a finite $B_g$
can significantly decrease the growth rate of the perpendicular mode ($\theta =90^\circ$
).
Figure 7. Maximum growth rate of the sausage mode as a function of the guide field strength $B_g/B_0$
for different $\theta$
(angle between $\boldsymbol{k}$
and the $x$
-axis). The jet is along the guide field ($\alpha =90^\circ$
). The Lundquist number is $S=10^4$
. The three panels are results for different flow speeds.
Figure 8. Maximum growth rate of the sausage mode as a function of $\theta$
for $S=10^4$
, $V_0/B_0=1$
, $\alpha =90^\circ$
and different $B_g/B_0$
.
5. Summary
In this study, we adopt a BVP solver to study the instabilities inside a current sheet with the presence of a plasma jet. When the jet is collimated with the antiparallel component of the magnetic field, both of the sausage mode and kink mode can be stabilized by the magnetic field. Without resistivity, the stability thresholds for the kink mode and sausage mode are $V_0/B_0\approx 2$
and $V_0/B_0 \approx 1$
, respectively (figure 3). With a finite resistivity, the streaming sausage mode couples with the tearing mode, but the streaming kink mode is not modified by the resistivity significantly unless the resistivity is very large ($S<100$
). Thus, in most of space and laboratory current sheets where $S$
is extremely large, the kink mode can be excited only if the jet speed is large ($V_0/B_0 \gtrsim 2$
). For $V_0/B_0 \lesssim 1$
, the sausage mode is tearing-like, with a power-law relation between the maximum growth rate and the Lundquist number $\max (\gamma )\propto S^{-1/2}$
in the large $S$
limit, same as the tearing mode without flow, while the values of the maximum growth rate increase with the ratio $V_0/B_0$
. For $V_0/B_0 \gtrsim 1$
, the sausage mode gradually transits to more streaming-like, and the maximum growth rate becomes less dependent on $S$
(figure 4). In the case of a jet flowing along the direction perpendicular to the antiparallel component of the magnetic field, our result reveals that, once the jet speed exceeds a threshold which is determined by the Lundquist number, the maximum growth rate of the sausage mode may increase with the angle between the wavevector and the reconnecting magnetic field component (figure 6). This is because the mode transits from the pure tearing to pure streaming as the wavevector rotates from the antiparallel magnetic field direction to the jet direction. Last, the out-of-plane jet combined with a finite guide field leads to a complex behaviour of the maximum growth rate of the sausage mode. With certain $V_0/B_0$
and $\theta$
values (figures 7c and 8), the maximum growth rate increases with the guide field strength. But the increase is not very large and is non-monotonically dependent on $\theta$
. For example, figure 7(c) shows that the increase in $\max (\gamma )$
with $B_g$
from $B_g/B_0=0$
to $B_g/B_0=0.5$
is larger for $\theta =30^\circ$
and 45$^\circ$
than that for $\theta =40^\circ$
. More importantly, overall the guide field quenches both the oblique tearing mode and the streaming sausage mode. As a result, increasing $B_g$
will gradually turn the monotonically increasing $\max (\gamma )$
–$\theta$
curve to a curve that increases at first and then drops (figure 8).
These results indicate that plasma flow plays an important role in destabilizing the current sheets in space and laboratory plasma. A jet whose width is comparable to that of the current sheet and peak speed similar to the upstream Alfvén speed can enhance the maximum growth rate of the tearing mode to more than twice of that in the no-flow case (figure 3). When the jet has a finite component along the direction perpendicular to the antiparallel component of the magnetic field, even if the component is much smaller than the upstream Alfvén speed, the oblique sausage mode ($\theta >0$
) may have comparable or even larger growth rate than the parallel sausage mode ($\theta = 0$
), and the most unstable mode may be perpendicular ($\boldsymbol{k}=k \hat {e}_z$
) (figure 6). The reason is that the out-of-plane flow (along the $z$
direction) does not feel the stabilization effect by the magnetic field along $x$
, and the growth rate of the pure streaming sausage mode is usually much larger than the pure tearing mode in the large $S$
limit. When the out-of-plane jet and guide field coexist, the most unstable mode may be oblique rather than parallel or perpendicular (figure 8).
We note that several factors which are absent in this study may have non-negligible effects on the analysed instabilities. Here we assume a uniform density profile and incompressibility. However, compressible MHD simulations show that a non-uniform background plasma density such as in the magnetotail can modify the growth rate of both tearing and streaming modes (Hoshino & Higashimori Reference Hoshino and Higashimori2015). In addition, if the Hall effect is included, out-of-plane components of the magnetic field and velocity perturbations are generated even for the parallel mode ($\boldsymbol{k}=k \hat {e}_x$
). Therefore, the out-of-plane jet will modify both the oblique and the parallel modes. Moreover, different widths of the jet and current sheet will change the results (Einaudi & Rubini Reference Einaudi and Rubini1986; Hoshino & Higashimori Reference Hoshino and Higashimori2015). As a final remark, it is worth noting that in the collisionless regime where the electron inertia is the only mechanism that breaks the frozen-in condition, an out-of-plane plasma jet plays a stabilizing role of the tearing mode even if the mode is parallel ($\boldsymbol{k}=k \hat {e}_x$
) (Tassi, Grasso & Comisso Reference Tassi, Grasso and Comisso2014). This is very different from the resistive-MHD regime where the out-of-plane jet only modifies the oblique tearing mode.
Acknowledgements
The author thanks Professor M. Velli and Dr K. Zhang for very helpful suggestions and comments, and the SciPy team for implementing the boundary value solver in Python (Virtanen et al. Reference Virtanen, Gommers, Oliphant, Haberland, Reddy, Cournapeau, Burovski, Peterson, Weckesser and Bright2020).
Editor F. Califano thanks the referees for their advice in evaluating this article.
Funding
The work was supported by NASA HERMES DRIVE Science Center grant no. 80NSSC20K0604.
Declaration of interest
The authors report no conflict of interest.
Open access
