Landau Damping

Our starting point is the Vlasov equation for an unmagnetized, collisionless plasma:

where is the ensemble-averaged electron distribution function. The electric field satisfies

where

Here, is the number density of ions (which is the same as the equilibrium number density of electrons).

Because we are dealing with small amplitude waves, it is appropriate to linearize the Vlasov equation. Suppose that the electron distribution function is written

(8.4) |

Here, represents the equilibrium electron distribution, whereas represents the small perturbation due to the wave. Of course, , otherwise the equilibrium state would not be quasi-neutral. The electric field is assumed to be zero in the unperturbed state, so that can be regarded as a small quantity. Thus, linearization of Equations (8.1) and (8.3) yields

and

respectively.

Let us now follow the standard procedure for analyzing small amplitude waves, by assuming that all perturbed quantities vary with and like . Equations (8.5) and (8.6) reduce to

and

respectively. Solving the first of these equations for , and substituting into the integral in the second, we conclude that if is non-zero then we must have

We can interpret Equation (8.9) as the dispersion relation for electrostatic plasma waves, relating the wavevector, , to the frequency, . However, in doing so, we run up against a serious problem, because the integral has a singularity in velocity space, where , and is, therefore, not properly defined.

The way to resolve this problem was first explained by Landau in a very influential paper that was the foundation of much subsequent work on plasma oscillations and instabilities (Landau 1946). Landau showed that, instead of simply assuming that varies in time as , the problem must be regarded as an ``initial value problem'' in which is specified at , and calculated at later times. We may still Fourier analyze with respect to , so we write

(8.10) |

It is helpful to define as the velocity component along (i.e., ), and to also define and as the integrals of and , respectively, over the velocity components perpendicular to . Thus, Equations (8.5) and (8.6) yield

and

respectively, where .

In order to solve Equations (8.11) and (8.12) as an initial value problem, we introduce the Laplace transform of with respect to (Riley 1974):

(8.13) |

If the rate of increase of with increasing is no faster than exponential, then the integral on the right-hand side of the previous equation converges, and defines as an analytic function of , provided that the real part of is sufficiently large.

Noting that the Laplace transform of is (as is easily shown by integration by parts), we can Laplace transform Equations (8.11) and (8.12) to obtain

and

(8.15) |

respectively. The previous two equations can be combined to give

(8.16) |

yielding

where

The function is known as the

According to Equations (8.14) and (8.17), the *Laplace transform* of the distribution function is written

(8.19) |

or

Having found the Laplace transforms of the electric field and the perturbed
distribution function, we must now invert them to obtain
and
as functions of time. The *inverse Laplace transform*
of the distribution function is given by (Riley 1974)

where --the so-called

Rather than trying to obtain a general expression for , from Equations (8.20) and (8.21), we shall concentrate on the behavior of the perturbed distribution function at large times. Looking at Figure 8.1, we note that if has only a finite number of simple poles in the region (where is real and positive) then we may deform the contour as shown in Figure 8.2, with a loop around each of the singularities. A pole at gives a contribution that varies in time as , whereas the vertical part of the contour gives a contribution that varies as . For sufficiently large times, the latter contribution is negligible, and the behavior is dominated by contributions from the poles furthest to the right.

Equations (8.17), (8.18), and (8.20) all involve integrals of the form

Such integrals become singular as approaches the imaginary axis. In order to distort the contour , in the manner shown in Figure 8.2, we need to continue these integrals smoothly across the imaginary -axis. As a consequence of the way in which the Laplace transform was originally defined--that is, for sufficiently large--the appropriate way to do this is to take the values of these integrals when lies in the right-hand half-plane, and to then find the analytic continuation into the left-hand half-plane (Flanigan 2010).

If is sufficiently well-behaved that it can be continued off the real axis as an analytic function of a complex variable , then the continuation of (8.22) as the singularity crosses the real axis in the complex -plane, from the upper to the lower half-plane, is obtained by letting the singularity take the contour with it, as shown in Figure 8.3 (Cairns 1985).

Note that the ability to deform the Bromwich contour into that of Figure 8.3, and so to find a dominant contribution to and from a few poles, depends on and having smooth enough velocity dependences that the integrals appearing in Equations (8.17), (8.18), and (8.20) can be analytically continued sufficiently far into the lower half of the complex -plane (Cairns 1985).

If we consider the electric field given by the inversion of Equation (8.17), then we see that its behavior at large times is dominated by the zero of that lies furthest to the right in the complex -plane. According to Equations (8.20) and (8.21), has a similar contribution, as well as a contribution that varies in time as . Thus, for sufficiently long times after the initial excitation of the wave, the electric field depends only on the positions of the roots of in the complex -plane. The distribution function, on the other hand, has corresponding components from these roots, as well as a component that varies in time as . At large times, the latter component of the distribution function is a rapidly oscillating function of velocity, and its contribution to the charge density, obtained by integrating over , is negligible.

As we have already noted, the function is equivalent to the left-hand side of Equation (8.9), provided that is replaced by . Thus, the dispersion relation, (8.9), obtained via Fourier transformation of the Vlasov equation, gives the correct behavior at large times, as long as the singular integral is treated correctly. Adapting the procedure that we discovered using the complex variable , we see that the integral is defined as it is written for , and analytically continued, by deforming the contour of integration in the -plane (as shown in Figure 8.3), into the region . The simplest way to remember how to do the analytic continuation is to observe that the integral is continued from the part of the -plane corresponding to growing perturbations to that corresponding to damped perturbations. Once we know this rule, we can obtain kinetic dispersion relations in a fairly direct manner, via Fourier transformation of the Vlasov equation, and there is no need to attempt the more complicated Laplace transform solution.

In Chapter 5, where we investigated the cold-plasma dispersion relation, we found that
for any given
there were a finite number of values of
, say
,
,
, and a general solution was a linear superposition of
functions varying in time as
,
, et cetera. The set of values of
corresponding
to a given value of
is called the
*spectrum* of the wave. It is clear that the cold-plasma equations yield a discrete wave spectrum.
On the other hand, in the kinetic problem, we obtain contributions
to the distribution function that vary in time as
,
with
taking any real value. In other words, the kinetic equation yields a continuous wave spectrum.
All of the mathematical difficulties of the kinetic
problem arise from the existence of this continuous spectrum (Cairns 1985). At
short times, the behavior is very complicated, and depends on the details
of the initial perturbation. It is only asymptotically that a mode
varying in time as
is obtained, with
determined
by a dispersion relation that is solely a function of the unperturbed state.
As we have seen, the emergence of such a mode depends on the initial velocity
disturbance being sufficiently smooth.

Suppose, for the sake of simplicity, that the background plasma state is a Maxwellian distribution. Working in terms of , rather than , the kinetic dispersion relation for electrostatic waves takes the form

where

Suppose that, to a first approximation, is real. Letting tend to the real axis from the domain , we obtain

where denotes the

Incidentally, because Equation (8.25) holds for any well-behaved distribution function, it follows that

This famous expression is known as the

Suppose that is sufficiently small that over the range of where is non-negligible. It follows that we can expand the denominator of the principal part integral in a Taylor series:

Integrating the result term by term, and remembering that is an odd function, Equation (8.23) reduces to

where is the electron plasma frequency. Equating the real part of the previous expression to zero yields

where is the Debye length, and it is assumed that . We can regard the imaginary part of as a small perturbation, and write , where is the root of Equation (8.29). It follows that

(8.30) |

and so

giving

If we compare the previous results with those for a cold plasma, where
the dispersion relation for an electrostatic plasma wave was found to
be simply
(see Section 5.7), we see, first, that
now depends on
,
according to Equation (8.29), so that, in a warm plasma, the electrostatic plasma
wave is a propagating mode, with a non-zero group-velocity. Such a mode is known as a *Langmuir wave*. Second, we
now have
an imaginary part to
, given by Equation (8.32), corresponding, because
it is negative, to the damping of the wave in time. This damping is generally
known as *Landau damping*. If
(i.e.,
if the
wavelength is much larger than the Debye length) then the imaginary part
of
is small compared to the real part, and the wave is only
lightly damped. However, as the wavelength becomes comparable to the
Debye length, the imaginary part of
becomes comparable to the
real part, and the damping becomes strong.
Admittedly, the approximate solution given previously
is not very accurate in the short wavelength case, but it is nevertheless sufficient to indicate
the existence of very strong damping.

There are no dissipative effects explicitly included in the collisionless Vlasov equation. Thus, it can easily be verified that if the particle velocities are reversed at any time then the solution up to that point is simply reversed in time. At first sight, this reversible behavior does not seem to be consistent with the fact that an initial perturbation dies out. However, we should note that it is only the electric field that decays in time. The distribution function contains an undamped term varying in time as . Furthermore, the decay of the electric field depends on there being a sufficiently smooth initial perturbation in velocity space. The presence of the term means that, as time advances, the velocity space dependence of the perturbation becomes more and more convoluted. It follows that if we reverse the velocities after some time then we are not starting with a smooth distribution. Under these circumstances, there is no contradiction in the fact that, under time reversal, the electric field grows initially, until the smooth initial state is recreated, and subsequently decays away (Cairns 1985).

Landau damping was first observed experimentally in the 1960s (Malmberg and Wharton 1964; Malmberg and Wharton 1966; Derfler and Simonen 1966).