Compressible flow

Compressible flow allows density to change materially and is required for finite-amplitude pressure waves and shocks.

A compressible flow that passes between subsonic and supersonic regions. A smooth steady crossing of a sonic point requires the corresponding regularity condition; a discontinuous transition can instead involve a shock wave.

The temperature reached by bringing a flow to rest adiabatically while converting its kinetic energy to enthalpy. For a calorically perfect gas, $T_{\mathrm{stag}}=T+u^2/(2c_p)=T[1+(\gamma -1)M^2/2]$. It is conserved across an adiabatic normal shock even though entropy increases and stagnation pressure decreases.

Flow whose mass flux is limited by a sonic condition in a controlling section, for fixed upstream thermodynamic conditions. In a smooth isentropic converging-diverging nozzle, the throat reaches Mach number one. Lowering downstream pressure further cannot indefinitely increase that fixed-upstream throat mass flux; shocks and downstream structure can still change.

A model representing a passage by its varying cross-sectional area and approximately uniform section-averaged flow variables. It neglects detailed transverse velocity and thermodynamic structure. In a smooth lossless nozzle, conservation laws then connect area to Mach number; this model does not by itself determine the curved sonic surface of an arbitrary multidimensional passage.

Across a thin steady transition with negligible external impulse, mass conservation and momentum conservation give the displayed relations. If the gas stays isothermal while neutral atomic hydrogen becomes molecular hydrogen, the mean particle mass doubles and $a_2^2=a_1^2/2$ for its isothermal sound speed. Let $J=u_1+a_1^2/u_1$. The downstream subsonic branch has $u_2=[J-\sqrt{J^2-2a_1^2}]/2$, ${\rho }_2/{\rho }_1=u_1/u_2$, and $p_2-p_1={\rho }_1{u}_1(u_1-u_2)>0$.

For $u_1\ll a_1$, the leading behavior is $u_2\simeq u_1/2$, ${\rho }_2\simeq 2{\rho }_1$ and $p_2\simeq p_1$, at unchanged temperature. Radiation of the molecular binding energy prevents that released energy from necessarily heating the gas; it does not, by itself, establish an isothermal thermal closure. If temperature evolution is allowed, an energy balance and a cooling model are required to determine all four changes uniquely.

A self-similar blast wave is an expanding shock wave whose post-shock profiles retain their shape in a rescaled radius or distance. A fixed explosion energy, ambient mass density and geometry set the front scaling through conservation of total energy. The dimensionless coefficient requires the profiles and an energy-normalization convention.

Use the fields of pressure similarity in a spherical blast wave, with front radius $R={\xi }_0\ell$. Since ${\rho }_0{\ell }^5/t^2=E$, changing the radial integration variable to $\xi$ givesBoth are time independent for fixed integrable similarity profiles and constant front coordinate ${\xi }_0$. With explosion energy $E$ and negligible external energy input or losses, their sum equals $E$. This normalization, together with shock matching and regularity, fixes the dimensionless blast coefficient rather than dimensional analysis alone.

For a spherical self-similar blast wave, put $\ell =(E{t}^2/{\rho }_0{)}^{1/5}$, $\xi =r/\ell$, $\rho ={\rho }_0{f}_{\rho }(\xi )$ and $u=(r/t)f_u(\xi )$. The radial Euler equation then requires $p={\rho }_0(\ell /t{)}^2{f}_p(\xi )$, whereThis follows from ${\partial }_t\xi =-2\xi /(5t)$ and ${\partial }_r\xi =1/\ell$. The momentum equation alone leaves a spatially uniform additive pressure undetermined. The scale-free strong-shock boundary data fix that freedom consistently with the displayed similarity scaling; an independently prescribed ambient pressure introduces another scale.

Let a strong shock wave of radius $R$ and speed $V$ sweep a uniform mass density ${\rho }_0$ into a shell of mass $4\pi {\rho }_0{R}^3/3$ and speed $2V/(\gamma +1)$. If the cavity pressure is $\alpha$ times the immediate post-shock pressure, shell momentum balance gives $R\dot{V}=3(\alpha -1)V^2$. Constant total energy gives $R^3{V}^2=\mathrm{constant}$ and hence $\alpha =1/2$. Including shell internal and kinetic energy, and treating the cavity volume as $4\pi R^3/3$ to thin-shell accuracy, yields $R=\xi (E/{\rho }_0{)}^{1/5}{t}^{2/5}$ with ${\xi }^5=75(\gamma -1)(\gamma +1{)}^2/[16\pi (5\gamma -3)]$. This coefficient is approximate; resolving the shell's finite volume changes it.

For an energy-conserving strong spherical shock wave expanding into mass density $C{r}^{-\beta }$, $0\le \beta <3$, dimensional analysis gives the displayed radius law. Neglect ambient pressure, radiation losses and gravity in this blast-dominated approximation. Write $u=\dot{R}U(\xi )$, $\rho =C{R}^{-\beta }D(\xi )$ and $p=C{R}^{-\beta }{\dot{R}}^{2}P(\xi )$, with $\xi =r/R$. The radial continuity equation, momentum equation and adiabatic pressure equation reduce toThe Strong-shock Rankine-Hugoniot conditions give $U(1)=P(1)=2/(\gamma +1)$ and $D(1)=(\gamma +1)/(\gamma -1)$. The explosion energy fixes the radius coefficient through $E=4\pi C{R}^{3-\beta }{\dot{R}}^2{\int }_0^1[D{U}^2/2+P/(\gamma -1)]{\xi }^{2}d\xi$.

The similarity profiles $U=2\xi /(\gamma +1)$, $D=(\gamma +1)\xi /(\gamma -1)$ and $P=2{\xi }^3/(\gamma +1)$ solve the spherical blast equations at the displayed ambient exponent. For $\gamma =5/3$, $\beta =2$ and $R^3=3E{t}^2/(2\pi C)$. The dimensional fields are $u=r/(2t)$, $\rho =8\pi C^{2}r/(3E{t}^2)$ and $p=2\pi C^2{r}^3/(9E{t}^4)$ for $0<r<R$. Their kinetic and internal energies are equal. Substituting these profiles into all three equations and the shock conditions verifies the solution.

For a planar adiabatic explosion with energy per unit area $E$ in a uniform medium of mass density ${\rho }_0$, the conserved energy has form $E=I{\rho }_{0}Z{\dot{Z}}^2$, with a finite positive integral $I$ of the similarity profiles. Hence $Z=C(E/{\rho }_0{)}^{1/3}{t}^{2/3}$. If $E$ is shared by two fronts, its fraction assigned to either side is absorbed into $C$.

A polytropic flow uses a polytropic equation of state. When $K$ and the adiabatic exponent are constant, a smooth streamline has specific enthalpy $h=\gamma K{\rho }^{\gamma -1}/(\gamma -1)$ and adiabatic sound speed $c_s^2=\gamma K{\rho }^{\gamma -1}$. Spatially varying entropy can make $K$ depend on streamline; the constant-$K$ spherical wind is a special case.

A smooth radial compressible flow with a fixed adiabatic exponent conserves $r^2\rho u$ and $p/{\rho }^{\gamma }$. In a prescribed Newtonian gravitational potential, its Bernoulli function is $u^2/2+c^2/(\gamma -1)+\Phi$, with $c$ the adiabatic sound speed. Eliminating the density derivative yields $(u-c^2/u)u^{\prime }=2c^2/r-{\Phi }^{\prime }$.

A smooth transonic branch in a power-law gravitational potential has $u_{\ast }^2=c_{\ast }^2=A\beta /(2r_{\ast }^{\beta })$ at its sonic point. Substitution into the Bernoulli function gives $\mathcal{B}=c_{\ast }^2[1/(\gamma -1)+1/2-2/\beta ]$. The local sonic-point slope discriminant distinguishes a genuine crossing from a degenerate everywhere-sonic scaling solution.

With reservoir density ${\rho }_{\infty }$ and adiabatic sound speed $c_{\infty }$, put $q=2/\beta -1/2$ and $h=1-q(\gamma -1)>0$. The selected mass accretion rate is $4\pi {\rho }_{\infty }(A\beta /2{)}^{2/\beta }{c}_{\infty }^{1-4/\beta }{h}^{q-1/(\gamma -1)}$. It follows by evaluating the conserved spherical flux at the sonic point.

The dimensionless accretion factor $(1-q\delta {)}^{q-1/\delta }$ tends to $e^q$ as $\delta \to 0$. For $q>0$, it tends to $1$ as $\delta \uparrow 1/q$, although the sonic point moves to the origin. Writing the latter logarithm as $-qh\ln h/(1-h)$ makes the finite limit transparent. These limits connect transonic spherical accretion rate in a power-law potential to the isothermal equation of state and the limiting critical adiabatic exponent.

For $0<\beta <4$, a nondegenerate transonic spherical flow in a power-law potential connected to infinity requires $\gamma <(4+\beta )/(4-\beta )$. For $\beta \ge 4$, every finite $\gamma >1$ satisfies the local crossing condition; there is no finite upper bound. The uniformly valid inequality is $2\beta -(4-\beta )(\gamma -1)>0$.

An isothermal shock joins two states at the same temperature and isothermal sound speed. Mass and momentum conservation imply $u_1{u}_2=c_s^2$ for a nontrivial planar jump, so the compression ratio is the upstream isothermal Mach number squared. Heat exchange replaces adiabatic energy conservation across the gas alone.

Suppose molecular hydrogen is fully dissociated into atoms across a steady planar shock, with the same far-upstream and far-downstream temperature and negligible binding-energy mass correction. The upstream isothermal sound speed obeys $a_1^2=k_{B}T/(2m_H)$, while $a_2^2=2a_1^2$. Conservation of mass gives $\chi ={\rho }_2/{\rho }_1=u_1/u_2$. Momentum conservation then gives $u_1^2+a_1^2=u_1^2/\chi +2a_1^2\chi$, producing the displayed quadratic with ${\mathcal{M}}_1=u_1/a_1$.

The high root approaches ${\mathcal{M}}_1^2/2$ at large upstream Mach number. Its real compressive domain is ${\mathcal{M}}_1\ge 1+\sqrt{2}$. At the endpoint the downstream flow is sonic; above it the high root is subsonic downstream. Completion of dissociation and maintenance of equal temperatures additionally require a consistent chemical and radiative energy budget. The two flux equations alone do not establish that physical realizability.

The specific enthalpy is equal on both sides of an isothermal perfect-gas shock. The energy removed per unit area per unit time is $j(u_1^2-u_2^2)/2$, with $j>0$ the normal mass flux. Net cooling therefore selects compression shocks; an expansion jump requires heating.

For an inviscid fluid with specific internal energy $e$, let $E=\rho (e+|\mathbf{u}{|}^2/2)$. In a prescribed potential $\Phi$, ${\partial }_{t}E+\nabla \cdot [(E+p)\mathbf{u}]=-\rho \mathbf{u}\cdot \nabla \Phi +Q$, where $Q$ is external heat supply. A static gravitational potential can instead be included in the conserved energy.

For steady spherical isentropic flow with $p=K{\rho }^{3/2}$ around a point mass, put $j=r^2\rho v>0$ and $\mathcal{E}=v^2/2+2c^2-GM/r>0$. Eliminating the mass density gives $c^2=[j(3K/2{)}^2{]}^{2/5}{r}^{-4/5}{\mathcal{M}}^{-2/5}$. With $r_s=GM/(4\mathcal{E})$, $x=(r/r_s{)}^{1/5}$, $y={\mathcal{M}}^{2/5}$ and $\lambda =[j(3K/2{)}^2{]}^{2/5}/(2\mathcal{E}{r}_s^{4/5})$, the Bernoulli equation reduces to $\lambda F(y)=F(x)$, where $F(t)=t^4+4/t$. Since $F$ has its unique minimum at $t=1$, a regular sonic point requires $\lambda =1$ and $r=r_s$. One transonic branch is $y=x$, which has $v=\sqrt{2\mathcal{E}}$, $\rho \propto r^{-2}$ and $p\propto r^{-3}$. The decreasing branch has inner free fall and an outer nearly static reservoir, giving Bondi accretion for inward fluid flow. A nonradiative stationary shock wave keeps $j,\mathcal{E},r_s$ fixed but raises $K$, so $\lambda$ increases by the factor $(K_2/K_1{)}^{4/5}$. The inward transonic solution is the spherical $n=2$, $\gamma =3/2$ specialization of transonic accretion in a power-law tube.

The speed of sound is the propagation speed of a small pressure disturbance through a medium.

The isothermal sound speed is $a_i^2=(\partial p/\partial \rho {)}_T$. For an ideal gas it is $a_i^2=p/\rho =\mathcal{R}T/\mu$.

The adiabatic sound speed is $v_s^2=(\partial p/\partial \rho {)}_s$. For a perfect gas it is $v_s^2=\gamma p/\rho$.

An isentropic flow has constant entropy along the fluid and contains no entropy-producing shocks or dissipation. For a perfect gas, $p/{\rho }^{\gamma }$ is constant and $c_s^2=\gamma p/\rho \propto {\rho }^{\gamma -1}$.

In one spatial dimension, inviscid isentropic mass and momentum conservation are

A homentropic flow has one spatially uniform entropy value throughout the fluid. It is therefore isentropic along every fluid trajectory, and its pressure can be treated as a single-valued function of density.

A globally isothermal fluid has spatially and temporally constant sound speed $c_s$ and pressure $p=c_s^2\rho$. Maintaining this relation generally requires heat exchange rather than adiabatic evolution.

A sonic point is a point at which the flow speed equals the local sound speed, so the Mach number is one.

In a nonzero steady one-dimensional flow with an isothermal equation of state with $p=c_s^2\rho$ and constant cross section, the continuity equation gives $\rho u=\text{constant}$. Eliminating density from momentum gives $(u-c_s^2/u)u^{\prime }=-{\Phi }^{\prime }$. With Mach number $\mathcal{M}=|u|/c_s$, integration yields ${\mathcal{M}}^2/2-\ln \mathcal{M}+\Phi /c_s^2=C$. The function $f(\mathcal{M})={\mathcal{M}}^2/2-\ln \mathcal{M}$ has minimum $1/2$ at $\mathcal{M}=1$, so a smooth transonic crossing at $x_{\ast }$ requires a local maximum of $\Phi$ and $C=1/2+\Phi (x_{\ast })/c_s^2$. At a nondegenerate maximum, $f=1/2+(\mathcal{M}-1{)}^2+\cdots$ gives ${\mathcal{M}}^{\prime }(x_{\ast })=\pm \sqrt{-{\Phi }^{\prime \prime }(x_{\ast })/(2c_s^2)}$.

For spherically symmetric adiabatic flow in a power-law gravitational potential, let $X=(r_{\ast }/u_{\ast })u_{\ast }^{\prime }$ and $\delta =\gamma -1$. Differentiation at a sonic point gives $(2+\delta )X^2+4\delta X+4\delta -2\beta =0$. Its discriminant is $8[2\beta -(4-\beta )\delta ]$; a positive value gives the two local crossing directions.

For a steady polytropic streamline with Bernoulli function $C_B$, the critical speed is the sound speed at which the scalar mass flux $\rho u$ is maximal. It satisfies

A transonic branch is a solution that passes continuously between subsonic and supersonic flow through a regular sonic point.

Consider steady isentropic flow toward a Newtonian gravitational potential $-GM/r$ through a tube of cross-sectional area $A(r)=C{r}^n$. Write $v>0$ for the inward speed and use a polytropic equation of state with specific-heat ratio $\gamma >1$. Mass conservation and the Euler equations for an inviscid fluid givewhere $c_s$ is the adiabatic sound speed. A regular sonic point therefore has $v_s=c_s$ and $r_s=GM/(n{c}_s^2)$. The Bernoulli equation, matched to a nearly stationary reservoir with sound speed $c_0$, givesFor $n>1/2$, a finite positive sonic point requires $1<\gamma <(2n+1)/(2n-1)$. Differentiating the flow equation at that point, with $x=r_s{v}_s^{\prime }/c_s$, givesIts discriminant is $4n[(2n+1)-(2n-1)\gamma ]$, so the same bound permits real regular slopes. The branch on which the Mach number rises inward selects the negative sign inFor a spherical tube $n=2$, this recovers the $5/3$ threshold of Bondi accretion; for a dipolar flux-tube area, $n=3$ gives $7/5$. The tube approximation must remain valid between the reservoir matching region and the accretor.

The Mach number is the ratio of fluid flow speed to local sound speed.

More precisely, the Mach number is the dimensionless quantitywhere $U$ is the magnitude of the body's relative velocity through the surrounding gas, and $c>0$ is the local sound speed measured in the gas rest reference frame. For a fluid flow past a stationary body, the same Mach number uses the local fluid speed in the body's rest reference frame. A Mach number below, equal to, or above one means subsonic flow, sonic flow, or supersonic flow, respectively. The denominator is a property of the local gas: its sound speed depends on temperature and composition.^[ref]

The relative velocity matters even when the body is at rest in a wind tunnel. A uniform change of reference frame subtracts the same velocity from body and gas, so their relative velocity and the Mach number stay unchanged under a Galilean transformation. Ground speed alone cannot supply $U$ when there is wind. For an original collinear example, a body moving at $680\mathrm{m}{\mathrm{s}}^{-1}$ over the ground through gas moving in the same direction at $170\mathrm{m}{\mathrm{s}}^{-1}$ has $U=510\mathrm{m}{\mathrm{s}}^{-1}$. With a stipulated sound speed $c=340\mathrm{m}{\mathrm{s}}^{-1}$, its Mach number is $1.5$, rather than $2$. A moving observer changes apparent acoustic wave arrival times, as in the radial Mach number, without changing the body's Mach number relative to the gas.

Why is the threshold one so useful? Acoustic waves carry small pressure disturbances at the sound speed. In subsonic flow, such information can travel upstream relative to the body; in a uniform supersonic flow, it is swept downstream. The Mach cone turns this difference in communication into visible geometry. Matching Mach number between a wind tunnel and flight is therefore essential to matching compressible flow effects.^[ref]

An actual supersonic wind tunnel experiment makes the distinction between wavefronts and shock waves concrete. In NASA Glenn's 1-by-1-foot wind tunnel, interchangeable converging-diverging nozzle blocks set the Mach number. A strut-mounted exhaust-nozzle model had a nosecone, an air plenum and transparent sidewalls for schlieren photography. Before installing the model, the researchers photographed the empty test section: two shock wave patterns originated at a wall joint after the nozzle block. This control guided model placement to keep reflected shock waves away from the plume-study region. The rig included internal total pressure and static pressure sensing and a traversing external static pressure probe. A baseline nozzle and previous pressure data supplied a planned validation comparison; the report distinguishes this plan from the observed empty-tunnel control. This establishes an apparatus origin for those visible shock waves, not a measurement of the Mach cone from an infinitesimal source. The reported finite shock wave angle cannot simply be inserted into the Mach angle formula to calibrate $M$: shock wave strength and geometry also matter.^[ref]

In conventional schlieren photography, a bright slit source and two concave mirrors send parallel light rays through the wind tunnel and focus them at a knife edge before an optical camera. Spatial changes in mass density cause refraction; the knife edge converts the deflection into image contrast, making shock waves visible. This is the general optical principle, rather than a specification of every component in the preceding rig. A photograph projects a three-dimensional fluid flow into two dimensions; contrast alone does not provide a calibrated pressure jump or acoustic wave amplitude.^[ref]

Take a steady moving point source in uniform resting gas, with constant sound speed $c$ and source velocity $U{\mathbf{e}}_x$. Place the source at the origin at the current time $t=0$. In the linear acoustics model, an emission at an earlier time $\tau <0$ has a spherical wavefront centered at $(U\tau ,0,0)$ with radius $-c\tau$. An axial cross-section is the circle in the displayed equation. It is centered at the old release position, not the source's present position. The circles describe the geometry of weak acoustic waves, not a measured shock wave or its pressure jump.

For $M=0.5$, every old circle contains the present source because its center is $U|\tau |$ from the origin while its radius is the larger distance $c|\tau |$. It reaches ahead of the source. For $M=2$, the rightmost point of every old circle is $(U-c)\tau <0$: the source outruns all of its old wavefronts. The continuously emitted wavefronts have a common geometric envelope, the Mach cone. The plotted discrete release times sample this continuous family; rotating the axial section about the motion axis recovers the three-dimensional construction.

The radial Mach number is the source velocity projected toward the observer, divided by the sound speed. It enters the moving-surface retarded Jacobian through $d|x-y(\tau )|/d\tau =-\hat{r}\cdot v$. The sign is directional: a source moving toward the observer has positive $M_r$ and compressed arrival times.

A subsonic flow has Mach number less than one, so its speed is below the local sound speed.

For uniform gas moving at $U{\mathbf{e}}_x$ past a resting body, an upstream-directed acoustic wave has body-frame velocity $(U-c){\mathbf{e}}_x$. When $0\le M<1$, this velocity points upstream. This is why the body can influence the upstream fluid flow through small pressure disturbances.

A sonic flow has Mach number exactly one: the local fluid speed relative to the body equals the local sound speed. In uniform gas flowing at $U=c$ past a resting body, an upstream-directed acoustic wave has zero body-frame velocity. A sonic point is a location where this equality holds. In the moving-point acoustic wavefronts construction, old spheres touch the source at the origin; the Mach angle tends to $90^{\circ }$ as the supersonic flow approaches this threshold.

A supersonic flow has Mach number greater than one, so the flow speed exceeds the local sound speed.

For uniform gas moving at $U{\mathbf{e}}_x$ past a resting body, even an acoustic wave directed against the fluid flow has body-frame velocity $(U-c){\mathbf{e}}_x$ downstream. More generally, an acoustic wave travelling along a unit vector $\mathbf{n}$ has body-frame velocity $U{\mathbf{e}}_x+c\mathbf{n}$; its $x$ component is at least $U-c>0$. All such weak signals move downstream and their possible directions fit inside a Mach cone.

For one-dimensional homentropic ideal-gas flow,is constant along the characteristic $dx/dt=u\pm c$.

For $U_x+A(U)U_y=(1,\sqrt{v}{)}^T$ with $A=\left(\begin{array}{cc}u& 2\\ 2v& u\end{array}\right)$ and $v>0$, the variables $R=u-2\sqrt{v}$ and $S=u+2\sqrt{v}$ decouple into an unforced and a uniformly forced Inviscid Burgers equation. Smooth data for $u,v$ need not be smooth data for these variables when $v$ vanishes. For example $u(0,y)=2y$, $v(0,y)=y^2$ gives $R_0=4\min (y,0)$ and $S_0=4\max (y,0)$. Their derivative corners travel along different characteristic families. In the positive quadrant, $R=0$ and $S=2x$ below $y=x^2$, whereas $S=2(x+2x^2+2y)/(1+4x)$ above. Reconstructing $u=(R+S)/2$, $v=(S-R{)}^2/16$ gives a continuous piecewise classical continuation whose $y$ derivative jumps at the parabola. Treating $\sqrt{y^2}$ as $y$ for all real $y$ would miss this obstruction. The degeneracy at zero density and the regularity class are essential parts of the Cauchy problem.

For a smooth hyperbolic system $U_t+A(U)U_x=S(U)$, suppose $\nabla R$ is a left eigenvector of $A$ with eigenvalue $\lambda$. The chain rule gives $({\partial }_t+\lambda {\partial }_x)R=\nabla R\cdot S$. Hence the homogeneous invariant remains an exact Riemann invariant when its gradient annihilates the source. For $U=(u,v)$, $A=\left(\begin{array}{cc}v& u\\ u^{-3}& v\end{array}\right)$, $S=(-u^3,u)$ and $u>0$, the function $R=v-1/u$ has $\nabla R=(u^{-2},1)$, eigenvalue $v+1/u$, and $\nabla R\cdot S=0$. On its zero level, the other characteristic speed vanishes and $v_t=1/v$ along constant spatial positions. Source terms need not destroy every invariant.

For a barotropic pressure law with $c^2=dp/d\rho$, the invariantsobey $[{\partial }_t+(u\pm c){\partial }_x]R_{\pm }=0$.

An initially resting perfect gas with adiabatic index $\gamma >1$ next to a piston $X(t)=-f{t}^2/2$ develops a smooth rarefaction. The incoming invariant gives $c_p=c_0-(\gamma -1)ft/2$. Pressure falls as $p_p=p_0(1-t/t_v{)}^{2\gamma /(\gamma -1)}$ until $t_v$, when the piston speed is $2c_0/(\gamma -1)$. Thereafter the gas edge moves at that limiting constant speed while the piston accelerates away, opening a vacuum gap.

A simple wave is a region of a hyperbolic flow in which one Riemann invariant is constant throughout, leaving the other invariant to determine all fluid variables.

For positive smooth tube area $A$, the equation $q_Z-q{q}_{\theta }+q(\log A{)}^{\prime }/2=0$ becomes $Q_{\zeta }-Q{Q}_{\theta }=0$ under the displayed transformation. Indeed, $g^{\prime }/g=-A^{\prime }/(2A)$ cancels geometrical spreading and ${\zeta }^{\prime }=g$ makes the remaining nonlinear and transport coefficients equal. Starting with $q(\theta ,Z_0)=f(\theta )$ gives $Q=f(\xi )$ and $\theta =\xi -f(\xi )\zeta$; characteristic crossing first occurs when $\zeta \sup f^{\prime }=1$, if that supremum is positive.

For area $A(R)=R^2$ and smooth data $q(\theta ,R_0)=f(\theta )$ at $R_0>0$, the transformed distance is $\zeta =R_0\log (R/R_0)$. If $M=\sup f^{\prime }>0$, the first shock wave occurs at the displayed radius; if $M\le 0$, there is no forward characteristic crossing. Nontrivial bounded data cannot instead be imposed at $R=0$ in this outgoing simple-wave model: along every characteristic, $Rq$ is constant, so a bounded limit at the origin forces that constant to vanish. With distance measured from a finite emitting surface, use $A(Z)=(R_0+Z{)}^2$ and subtract $R_0$ from the shock radius.

A characteristic field is linearly degenerate if the directional derivative of its speed along its right eigenvector vanishes. Its simple waves have constant propagation speed along the state curve, so they do not steepen by amplitude-dependent transport.

An ideal-MHD simple wave traces a one-parameter state-space curve whose tangent is a right eigenvector of the characteristic matrix. Its parameter obeys $q_t+v(q)q_x=0$ and its characteristic speed obeys the Inviscid Burgers equation. Nonlinear Alfvén waves are a linearly degenerate example.

For a one-dimensional velocity perpendicular to a transverse magnetic field, $c_f^2=\gamma p/\rho +B^2/({\mu }_0\rho )$. The primitive-variable characteristic matrix has two advective eigenvalues and the two speeds $u\pm c_f$. Integrating the fast eigenvectors keeps $p/{\rho }^{\gamma }$ and $B/\rho$ constant. The state parameter obeys ${\rho }_t+(u\pm c_f){\rho }_x=0$, so the characteristic speed satisfies the Inviscid Burgers equation.

With $B_x=0$, a fast simple wave in magnetohydrodynamics has constant $u_y$ and $B_y/\rho =K$. Its longitudinal velocity satisfies $d{u}_x/d\rho =\pm c_f/\rho$, and ${\rho }_t+(u_x\pm c_f){\rho }_x=0$. Along this state curve the speed derivative is $\pm [c_f/\rho +K^2/(2{\mu }_0{c}_f)]$, which is nonzero. Compressive profiles therefore exhibit shock formation by characteristic intersection, unlike a linearly degenerate characteristic field.

For $u_t+(c_0+bu)u_x=-\alpha u$, $b,\alpha >0$, the characteristic curves are $u=e^{-\alpha t}{u}_0(\xi )$ and $x=c_{0}t+\xi +(b/\alpha )(1-e^{-\alpha t})u_0(\xi )$. If $m=\max [-u_0^{\prime }]>0$, finite-time gradient blow-up occurs precisely for $\alpha <bm$, at $t_s=-{\alpha }^{-1}\log (1-\alpha /(bm))$. Equality $\alpha =bm$ still gives a positive Jacobian at each finite time; only the limiting map degenerates.

For a piston receding from initially stationary homentropic gas, the right-moving simple wave has $R_{-}=0$, so $c=c_0+(\gamma -1)u/2$. A $C_{+}$ characteristic emitted by the piston at time $s$ carries $u={\dot{X}}_p(s)$ and is the straight line

The gas state at a receding piston reaches vacuum when ${\dot{X}}_p=-2c_0/(\gamma -1)$, because then the sound speed and mass density both vanish. If the piston subsequently recedes faster, a vacuum gap opens between it and the gas boundary, which continues with velocity $-2c_0/(\gamma -1)$.

For $X_p(t)=-t^{1+\alpha }/(1+\alpha )$ with $0<\alpha \ll 1$, a point on the ray $x=(c_0-a)t$, $0<a<(\gamma +1)/2$, receives the characteristic emitted at the exponentially small timeAlthough $s\to 0$, the piston speed $-s^{\alpha }$ has the finite limit $-2a/(\gamma +1)$, producing the centered-rarefaction profile of the singular limit $\alpha =0$.

In a right-moving simple wave entering gas initially at rest with sound speed $c_0$, the other Riemann invariant is constant, soThe isentropic relationconverts the velocity disturbance into its nonlinear pressure disturbance.

A smooth compressive simple wave forms a shock when characteristics first intersect. If characteristics are parametrized by their emission time $\tau$, the first shock occurs at the minimum positive time for which $\partial x(\tau ,t)/\partial \tau =0$.

A smooth solution undergoes a gradient catastrophe when a spatial derivative becomes unbounded while the solution itself remains bounded. For $u_t+q^{\prime }(u)u_x=0$, characteristics have $x=s+t{q}^{\prime }(u_0(s))$ and $u=u_0(s)$; thus $u_x=u_0^{\prime }(s)/[1+t{q}^{\prime \prime }(u_0(s))u_0^{\prime }(s)]$. Vanishing of the characteristic projection derivative gives the first such time. A weak solution with an entropy condition can continue after this breakdown of the smooth description.

A normal shock is a discontinuity perpendicular to the flow direction across which mass, momentum, and total energy fluxes are conserved.

Property ratios for this discontinuity are tabulated in normal shock tables.

A tube in which releasing an initial pressure difference produces a shock, often accompanied by an expansion wave and a contact surface. Spatially separated pressure probes can measure shock arrival speed and jump size. The shock and the contact surface must be identified separately when interpreting the downstream state.

For a perfect-gas normal shock wave reflected from a stationary rigid wall, apply the Rankine-Hugoniot conditions for a perfect gas using signed velocity in each shock frame. The once-shocked state is the common left state for both applications. Its two relative speeds are roots of $v^2-(\gamma +1)u_{s}v/2-\gamma p_s/{\rho }_s=0$. Their product and the pressure jump eliminate the velocity, giving the displayed relation. A strong incident shock wave gives $p_1/p_s\to (3\gamma -1)/(\gamma -1)$.

For a moving level surface, $n=\nabla S/|\nabla S|$, $v_n=-S_t/|\nabla S|$, and the surface delta distribution is ${\delta }_s=|\nabla S|\delta (S)$. Piecewise classical conservation laws acquire a distributional flux defect $[b]\cdot n-v_n[a]$. Vanishing of this defect is the Rankine-Hugoniot condition.