Category Aerodynamics of Wings and Bodies

Irrotational Flow

Enough has been said about the subject of vorticity, its conservation and generation, that it should be obvious that an initially irrotational, uniform, inviscid flow will remain irrotational in the absence of heat transfer and of strong curved shocks. One important consequence of per­manent irrotationality is the existence of a velocity potential. That is, the equation

f = VxQ = 0 (1-55)

is a necessary and sufficient condition for the existence of a potential Ф such that

Q = V<f>, (1-56)

where Ф(х, у, г, t) or Ф(г, t) is the potential for the velocity in the entire flow. Its existence permits the replacement of a three-component vector by a single scalar as the principle dependent variable or unknown in theo­retical investigations.

Given the existence of Ф, we proceed to derive two important conse­quences, which will be used repeatedly throughout the work which follows.

1. The Bernoulli Equation for Irrotational Flow (Kelvin’s Equation).

This integral of the equations of fluid motion is derived by combining (1-3) and (1-28), and assuming a distant acting force potential:

DQ dQ IQ2′

Dt dt 2 у dt 2 у



Under our present assumptions,





„ /эф

= VW’


so that (1-57) can be rearranged into


^ – ol = 0.

p J


The vanishing of the gradient implies that, at most, the quantity involved will be a function of time throughout the entire field. Hence the least restricted form of this Bernoulli equation is

f + T + ly-11-™- <1Hi0)

In all generality, the undetermined time function here can be eliminated by replacing Ф with

Ф’ = Ф — jF(t) dt. (1-61)

This artificiality is usually unnecessary, however, because conditions are commonly known for all time at some reference point in the flow. For instance, suppose there is a uniform stream I7„ at remote points. There Ф will be constant and the pressure may be set equal to p„ and the force potential to SI» at some reference level.

F(t) = + j ~ — Q, * = const. (1-62)

The simplified version of (1-60) reads


~ + m2 – Vl]+ ^ + [0. – Q] = 0. (1-63)

dt JPtx p

= V — P°° ip*>Ul

2 П~ 7-1 /дФ

~ УМ* IL1 al " dt

Irrotational Flow
image15 image16

In isentropic flow with constant specific heat ratio 7, (1-63) is easily reorganized into a formula for the local pressure coefficient

Подпись: where image18,image19 Подпись: (1-65) (1-66)

Here a is the speed of sound and M = Ux/ax is Mach number. For certain other purposes, it is convenient to recognize that

(Note that, here and below, the particular form chosen for the barotropic relation is isentropic. Under this restriction, dp/dp, (dp/Sp),, and a2 all have the same meaning.)

This substitution in (1-63) provides a convenient means of computing the local value of a or of the absolute temperature T,

a2 – al = -(7 – 1) + i(Q2 – ul) + (Q. – Q)j • (1-67)

Finally, we remind the reader that the term containing the body-force potential is usually negligible in aeronautics.

2. The Partial Differential Equation for Ф. By substituting for p and Q in the equation of continuity, the differential equation satisfied by the velocity potential can be developed:

lf + V. Q_0. (Ь68)

The second term here is written directly in terms of Ф, as follows:

V ■ Q = V • (УФ) = V4,

і V2* =

Irrotational Flow Подпись: Э2Ф dz2 : Подпись: in Cartesians J • Подпись: (1-69)

which we identify as the familiar Laplacian operator,

To modify the first term of (1-68), take the form of Bernoulli’s equation appropriate to uniform conditions at infinity, for which, of course, a special case would be that of fluid at rest, U«, = 0. The body-force term is dropped for convenience, leaving

[ ~ ~ – UQ2 – ul). (1-70)

j Poo P 01

By the Leibnitz rule for differentiation of a definite integral,

Подпись:d fP dp __ 1 dP h* P ~ P

We then apply the substantial derivative operator to (1-71) and make use of the first three members of (1-66),

Подпись: Dp dP Jvoo P J Dt ГР г – ГР ■

dp _ d I dp

Подпись: a2 Dp p Dt Подпись: (1-72) Подпись: 1 dp Dp p dp Dt

Jpc p ~ Up Jp« p.

In view of (1-67) and of the simple relationship between the velocity vector and Ф, this is essentially the desired differential equation. If it is multiplied through by a2, one sees that it is of third degree in the unknown dependent variable and its derivatives. It reduces to an ordinary wave equation in a situation where the speed of sound does not vary signifi­cantly from its ambient values, and where the squares of the velocity components can be neglected by comparison with a2.

It is of interest that Garrick (1957) has pointed out that (1-74) can be reorganized into

v2*=Ml+о -v) (f+«• – v#) – h Ш *■ (i-re>

where the subscript conQc and on the substantial derivative is intended to indicate that this velocity is treated as a constant during the second application of the operators d/dt and (Q • V). Equation (1-75) is just a wave equation (with the propagation speed equal to the local value of a) when the process is observed relative to a coordinate system moving at the local fluid velocity Q.

The question of boundary conditions and the specialization of (1-74) and (1-75) for small-disturbance flows will be deferred to the point where the subject of linearized theory is first taken up.

1-8 The Acceleration Potential

image22,image23 Irrotational Flow

It is of interest that when the equations of fluid motion can be simplified to the form

In a manner paralleling the treatment of irrotational flow, we can conclude that

Подпись: (1-78)a = V¥,

where ¥ (r, t) is a scalar function called the acceleration potential. Clearly,

Подпись:¥ = П – ^ + G(t),


G(t) being a function of time that is usually nonessential.

Подпись: ІЕ I Pm p Подпись: Pm Подпись: (1-80)

The acceleration potential becomes practically useful when disturbances are small, so that

In the absence of significant body forces, we then have

Подпись: (1-81)– V


This differs only by a constant from the local pressure, and doublets of ¥ prove a very useful tool for representing lifting surfaces. The authors have not been able to construct a suitable partial differential equation for the acceleration potential in the general case, but it satisfies the same equation as the disturbance velocity potential in linearized theory.

Qualitative Effects of Viscosity

It is a common feature of most flows of engineering interest that the viscosity of the fluid is extremely small. The Reynolds number, Re = TJJ,/v, which gives an overall measure of the ratio of inertia forces to viscous forces, is in typical aeronautical applications of the order 10® or more. For large ships, Reynolds numbers of the order 109 are common. Viscosity can then only produce significant forces in regions of extremely high shear, i. e., in extremely thin shear layers where there is a substantial variation of velocity across the streamlines. The thickness of the laminar boundary layer on a flat plate of length l is approximately ~ o///Re. For Re = 108, 8/1 ~ 0.005, which is so thin that it cannot even be illustrated in a figure without expanding the scale normal to the plate. A turbulent boundary layer has a considerably greater thickness, reflecting its higher drag and therefore larger momentum loss; an approximate formula given in Schlichting (1960), p. 38, is 8/1 ~ 0.37/Re1/s. For


Re = 10® this gives 6/1 = 0.023, which is still rather small. Viscosity is only important in a very small portion of the turbulent boundary layer next to the surface, in the “viscous sublayer. ”

The transition from a laminar to a turbulent boundary layer is a very complicated process that depends on so many factors that precise figures for transition Reynolds numbers cannot be given. For a flat plate in a very quiet free stream (i. e., one having a rms turbulent velocity fluctua­tion intensity of 0.1% or less) transition occurs at approximately a distance from the leading edge corresponding to Re ~ З X 10®. With a turbulent intensity of only 0.3% in the oncoming free stream the transition Reynolds number decreases to about 1.5 X 10®. These distances are far beyond that for which the boundary layer first goes unstable. Stability calculations show that on a flat plate this occurs at Re ~ 10s. The complicated series of events between the point where instability first sets in until transition occurs has only recently been clarified (see Klebanoff, Tidstrom, and Sargent, 1962). Transition is strongly influenced by the pressure gradient in the flow; a negative (“favorable”) pressure gradient tends to delay it and a positive (“adverse ”) one tends to make it occur sooner. As a practical rule of thumb one can state that the laminar boundary layer can only be maintained up to the point of minimum pressure on the airfoil. On a so-called laminar-flow airfoil one therefore places this point as far back on the airfoil as possible in order to try to achieve as large a laminar region as possible. Laminar-flow airfoils work successfully for moderately high Reynolds numbers (<107) and low lift coefficients but require ex­tremely smooth surfaces in order to avoid premature transition. At very high Reynolds numbers transition starts occurring in the region of favor­able pressure gradient.

Both laminar and turbulent boundary layers will separate if they have to go through extensive regions of adverse pressure gradients. Separation will always occur for a subsonic flow at sharp corners, because there the pressure gradient would become infinite in the absence of a boundary layer. Typical examples of unseparated and separated flows are shown in Fig. 4-1. In the separated flow there will always be a turbulent wake

behind the body. In principle one could have instead a region of fairly quiescent flow in the wake, separated from the outer flow by a thin laminar shear layer attached to the laminar boundary layer on a body. However, a free shear layer, lacking the restraining effect of a wall, will be highly unstable and will therefore turn turbulent almost immediately. Because of the momentum loss due to the turbulent mixing in the wake the drag of the body will be quite large. On a thin airfoil at a small angle of attack the boundary layer will separate at the sharp trailing edge but there will be a very small wake so that a good model for the flow is the attached flow with the Kutta condition for the inviscid outer flow determining the circulation.

Constant-Density Flow; the Thickness Problem

Having shown how steady, constant-density flow results are useful at all subcritical M, we now elaborate them for the finite wing pictured in Fig. 7-1. As discussed in Section 5-2 and elsewhere above, it is convenient to identify and separate portions of the field which are symmetrical and
antisymmetrical in z, later adding the disturbance velocities and pressures in accordance with the superposition principle. The separation process involves rewriting the boundary conditions (5-30) as

«* = Tfx + etx~a at г = 0+ ,

do. „dh –

<pz — — Гу + в——– a at z = 0—

dx dx

where h(x, y) is proportional to the ordinate of the mean camber surface, while 2i? (ж, у) is proportional to the thickness distribution. The differential equation is, of course, the three-dimensional Laplace equation

V2<p = 0. (7-9)

As a starting point for the construction of the desired solutions, we adapt (2-28) to express the perturbation velocity potential at an arbitrary field point (ж, у, z),

Here n is the normal directed into the field, and the integrals must be carried out over the upper and lower surfaces of S. Dummy variables (xi, 2/i, zi) will be employed for the integration process, so the scalar distance is properly written

r = V(x — xi)2 + (y — yi)2 + (z — zi)2. (7-11)

In wing problems Zj = 0 generally.

Considering the thickness alone, we have

ip{x, y, z) = <p(X, y, —z)

for all z and the boundary condition

Moreover, no discontinuities of <p or its derivatives are expected anywhere else on or off the ж, y-plane. In (7-10), dS = dzj dyb the values of <p(xi, //і, 0+) and <p(xi, yi,0—) appearing in the integrals over the upper and lower surfaces are equal, while the values of д/дп(1/4ят) are equal and opposite. Hence the contributions from the first term in brackets cancel,

where (7-13) and (7-11) have been – employed. Physically, (7-14) states that the flow due to thickness can be represented by a source sheet over the planform projection, with the source strength per unit area being pro­portional to twice the thickness slope dg/dx. [Compare the two-dimensional counterpart, (5-50).]

Examination of (7-14) leads to the conclusion that the thickness prob­lem is a relatively easy one. In the most common situation when the shape of the wing is known and the flow field constitutes the desired information, one is faced with a fairly straightforward double integration. For certain elementary functions g(x, y) this can be done in closed form; otherwise it is a matter of numerical quadrature, with careful attention to the pole singularity at aq = x, у = y, when one is analyzing points on the wing z = 0. The pressure can be found from (5-31) and (7-14) as

There is no net loading, since Cp has equal values above and below the wing. Also the thickness drag works out to be zero, in accordance with d’Alem­bert’s paradox (Section 2-5). Finally, it should be mentioned that, for any wing with closed leading and trailing edges,

avydaq = Рте — !7le — 0. (7-16)

J chord OX

This means that the total strength of the source sheet in (7-14) is zero. As a consequence, the disturbance at long distances from the wing ap­proaches that due to a doublet with its axis oriented in the flight direction, rather than that due to a point source.

Constant-Density Inviscid Flow

1- 1 Introduction

For the present chapter we adopt all the limitations listed in Section 1-1, plus the following:

(1) p = constant everywhere; and

(2) The fluid was initially irrotational.

The former assumption implies essentially an infinite speed of sound, while the latter guarantees the existence of a velocity potential. Turning to (1-74), we see that the flow field is now governed simply by Laplace’s equation

V4 = 0. (2-1)

Associated with this differential equation, the boundary conditions prescribe the values of the velocity potential or its normal derivative over the surfaces of a series of inner and outer boundaries. These conditions may be given in one of the following forms:

(1) The Neumann problem, in which vn = дФ/дп is given.

(2) The Dirichlet problem, in which the value of Ф itself is given.

(3) The mixed (Poincare) problem, in which Ф is given over certain portions of the boundary and дФ/дп is given over the remainder.

A great deal is known about the solution of classical boundary-value problems of this type and more particularly about fluid dynamic applica­tions. Innumerable examples can be found in books like Lamb (1945) and Milne-Thompson (1960). The subject is by no means closed, how­ever, as will become apparent in the light of some of the applications presented in later sections and chapters. A useful recent book, which combines many results of viscous flow theory with old and new develop­ments on the constant-density inviscid problem, is the one edited by Thwaites (1960).

Among the very many concepts and practical solutions that might be considered worthy of presentation, we single out here a few which are especially fundamental and which will prove useful for subsequent work.


2- 2 The Three-Dimensional Rigid Solid Moving Through a Liquid

Let us consider a single, finite solid S moving through a large mass of constant-density fluid with an outer boundary 2, which may for many purposes be regarded as displaced indefinitely toward infinity. For much of what follows, S may consist of several three-dimensional solids rather than a single one. The direction of the normal vector n will now be re­garded as from the boundary, either the inner boundary S or the outer boundary 2, toward the fluid volume.

In the early sections of Chapter 3, Lamb (1945) proves the following important but relatively straightforward results for a noncirculatory flow, which are stated here without complete demonstrations:

(1) The flow pattern is determined uniquely at any instant if the boundary values of Ф or дФ/дп are given at all points of S and 2. One important special case is that of the fluid at rest remote from <S; then Ф can be equated to zero at infinity.

(2) The value of Ф cannot have a maximum or minimum at any interior point but only on the boundaries. To be more specific, the mean value of Ф over any spherical surface containing only fluid is equal to the value at the center of this sphere. This result is connected with the interpreta­tion of the Laplacian operator itself, which may be regarded as a measure of the “lumpiness” of the scalar field; Laplace’s equation simply states that this “lumpiness ” has the smallest possible value in any region.

(3) The magnitude of the velocity vector Q = |Q| cannot have a max­imum in the interior of the flow field but only on the boundary. It can have a minimum value zero at an interior stagnation point.

(4) If Ф = 0 or дФ/дп = 0 over all of S and 2, the fluid will be at rest everywhere. That is, no boundary motion corresponds to no motion in the interior.

Next we proceed to derive some less straightforward results.

1. Green’s Theorem. For the moment, let the outer boundary of the flow field remain at a finite distance. We proceed from Gauss’ theorem for any vector field A, (1-14),

j^A-ndS = – jjjv ■ AdV. (2-2)


(The minus sign here results from the reversal of direction of the normal vector.) Let Ф and Ф’ be two continuous functions with finite, single­valued first and second derivatives throughout the volume V. We do not yet specify that these functions represent velocity potentials of a fluid flow. Let


Constant-Density Inviscid Flow

Equations (2-4) and (2-5) are now substituted into Gauss’ theorem. After writing the result, we interchange the functions Ф and Ф’, obtaining two alternative forms of the theorem:

2. Kinetic Energy. As a first illustration of the application of Green’s theorem, let Ф in (2-6) be the velocity potential of some flow at a certain instant of time and let Ф’ = Ф. Of course, it follows that

Подпись: (2-8)у2Ф = VV = 0.

We thus obtain a formula for the integral of the square of the fluid particle speed throughout the field

“ -///lvt>’dr – "///«*"• 0«>

S+2 V V

Moreover, if we multiply the last member of (2-9) by one-half the fluid density p and change its sign, we recognize the total kinetic energy T of the fluid within V. There results


Such an integral over the boundary is often much easier to evaluate than a triple integral throughout the interior of the volume. In particular, ЭФ/дп is usually known from the boundary conditions. If the solid S is moving through an unlimited mass of fluid, with Q = 0 at infinity, it is a simple matter to prove that the integral over S vanishes. It follows that (2-10) need be integrated over only the inner boundary at the surface of the solid itself,


3. A Reciprocal Theorem. Another interesting consequence of Green’s theorem is obtained by letting Ф and Ф’ be the velocity potentials of two different constant-density flows having the same inner and outer bounding surfaces. Then, of course, the two Laplacians in (2-6) and (2-7) vanish, and the right-hand sides of these two relations are found to be equal. Equating the left-hand sides, we deduce


S+S S+2

Constant-Density Inviscid Flow Подпись: Cv ~ P«) +^Q2 + P(O« - o) Подпись: (2-13)

4. The Physical Interpretation of Ф. To assist in understanding the significance of the last two results and to give a meaning to the velocity potential itself, we next demonstrate an artificial but nevertheless meaning­ful interpretation of Ф. We begin with Bernoulli’s equation in the form (1-63), assuming the fluid at infinity to be at rest and evaluating the pres­sure integral in consequence of the constancy of p,

Imagine a process in which a system of very large impulsive pressures P = Ґ p dr (2-14)

is applied to the fluid, starting from rest, to produce the actual motion existing at a certain time t. In (2-14), т is a dummy variable of integration. We can make the interval of application of this impulse arbitrarily short, and integrate Bernoulli’s equation over it. [Incidentally, the same result is obtainable from the basic equations of fluid motion, (1-3) and (1-4), by a similar integration over the interval dt.] In the limit, the integrals of the pressure p„, pQ2/2, and p(Sl„ — fi) become negligible relative to that of the very large p, and we derive

pf = “p – (2-15)

J u-bt) от

Hence, P = — рф constitutes precisely the system of impulsive pressures required to generate the actual motion swiftly from rest. This process might be carried out, for example, by applying impulsive force and torque to the solid body and simultaneously a suitable distribution of impulsive pressures over the outer boundary 2. The total impulse thus applied will equal the total momentum in the instantaneous flow described by Ф. Unfortunately, both this momentum and the impulse applied at the outer boundary become indeterminate as 2 spreads outward toward infinity, so

that there are certain problems of physical interpretation when dealing with an externally unbounded mass of liquid.

In the light of this interpretation of the velocity potential, we reexamine the kinetic energy in (2-11), rewriting this result

Подпись: S+2Подпись: S+2image25(2-16)

The work done by an impulse acting on a system which starts from rest is known to be the integral over the boundary of the product of the impulse by one-half the final normal velocity at the boundary. Thus, the starting impulses do a total amount of work given by exactly the last member of (2-16). Since the system is a conservative one, this integral would be expected to equal the change of kinetic energy which, of course, is the final kinetic energy T in the present case. The difficulty in connection with carrying 2 to infinity disappears here, since the work contribution at the outer boundary can be shown to approach zero uniformly. Hence, the kinetic energy of an unbounded mass of constant-density fluid without circulation can be determined entirely from conditions at the inner bound­ary, and it will always be finite if the fluid is at rest at infinity.

The reciprocal theorem, (2-12), can be manipulated, by multiplication with the density, into the form


As such, it becomes a special case of a fairly familiar theorem of dynamics which states that, for any two possible motions of the same system, the sum over all the degrees of freedom of the impulse required to generate one motion multiplied by the velocity in the second motion equals the same summed product taken with the impulses and velocities interchanged.

One final important result is stated without proof: it can be shown that for a given set of boundary conditions the kinetic energy Г of a liquid in a finite or infinite region is a minimum when the flow is acyclical and irrotational, relative to all other possible motions.

Boundary Layer on a Flat Plate

Подпись: (4-1)We shall consider the viscous laminar high Reynolds number flow over a semi-infinite flat plate at zero angle of attack. This is the simplest case of a boundary layer that may serve as a model for the calculation of bound­ary layer effects on a thin airfoil. For incompressible flow the Navier – Stokes equations and the equation of continuity read

The boundary conditions are that the velocity vanishes on the plate and becomes equal to XJx far away from the plate surface. Thus for two­dimensional flow

Подпись:Q(x, 0) = 0 for x > 0, (4-2)

Q иЛ £ог г ^ . (4-3)

We have assumed that the leading edge of the plate is located at the origin (see Fig. 4-2). By assuming that the plate is semi-infinite one avoids the problem of considering upstream effects from the trailing edge. These are actually quite small and do not show up in the boundary layer approximations to be derived.

In the process of introducing nondimensional coordinates a minor difficulty is encountered because there is no natural length in the problem to which spatial coordinates could be referred. We will circumvent this by selecting an arbitrary reference length and thereby implicitly assume that the behavior of the solution for x — 0(1) and z = 0(1) will be studied.

Constant-Density Flow; the Lifting Problem

For the purely antisymmetrical case we have

<p{x, y, z) = —<p{x, y, —z) (7-17)

with corresponding behavior in the velocity and pressure fields. The
principal boundary condition now reads

<Pz = — oc at z = 0± for (ж, y) on S. (7-18)


We must make allowance for a discontinuity in <p not only on the planform projection S but over the entire wake surface, extending from the trailing edge and between the wingtips all the way to x = +oo on the x, y-plane. For this reason, the simplest approach proves to be the use of (7-10) not as a means of expressing <p itself but the dimensionless ж-component of the perturbation velocity

и = <pX! (7-19)

where и is essentially the pressure coefficient, in view of (5-31). Equation (1-81) shows us that we are also working with a quantity proportional to the small-disturbance acceleration potential, and this is the starting point adopted by some authors for the development of subsonic lifting wing theory.

Obviously, и is a solution of Laplace’s equation, since the operation of differentiation with respect to x can be interchanged with V2 in (7-9). We may therefore write


We specify for the moment that S’ encompasses both wing and wake, since the derivation of (2-28) called for integration over all surfaces that are sources of disturbance and made no allowance for circulation around any closed curve in the flow external to the boundary. It is an easy matter to show, however, that the choice of и as dependent variable causes the first term in the (7-20) brackets to vanish except on S and the second term to vanish altogether. Because there can be no pressure jump except through a solid surface, и is continuous through z = 0 on the wake. But

are equal and opposite on top and bottom everywhere over S, so the first-term contributions remain uncanceled only on S. There и jumps by an amount

У(хі, V) = uu — m. (7-22)

(By antisymmetry, щ = —uu.)

As regards the second bracketed term in (7-20), the condition of irrota – tionality reveals that

du _ du _ dw dn dz і dxi

on top and one finds that

on the bottom of S’. Both w and its derivatives are continuous through all of S’, and therefore the upper and lower integrations cancel throughout. One is left with


We have inserted (7-21) and (7-22) here, along with the fact that


dz d(z1 — 2)

d_ dz

when applied to a quantity which is a function of these two variables only in the combination (2 — 21).

The modification of (7-24) into a form suitable for solving lifting-wing problems can be carried out in several ways. Perhaps the most direct is to observe that nearly always w(x, y, 0) is known over S, and (7-24) should therefore be manipulated into an expression for this quantity. This we do by noting that и = рх and w = so that

w(x, y, z) = u(x0, y, 2) dx0, (7-26)

where account has been taken that <p(— ж, у, z) — 0. Inserting (7-26) into (7-24) and interchanging orders of differentiation and integration, we get

w(x, y, z) = — ^.JjУ(хь Vi) s

_____________ dx0____________

V (x0 — Zl)2 + (у — У1)2 + 22′

When deriving (7-28b), an integration by parts* on yx is carried out at finite z. The singularity encountered as z —> 0 is then similar to the one in upwash calculation at a two-dimensional vortex sheet and can be handled by the well-known Cauchy principal value. The integral in (7-28a) containing 7(xj, уi) itself is, however, the more direct and directly useful form. In the process of arriving at it, we find ourselves confronted with the following steps:


V (x0 — Xl)2 + (y — Vi)2 + Z2

For г = 0 all terms here will vanish formally except the one arising from the 2-derivative of the numerator which will give a nonintegrable singu­larity of the form 1 /{y — уi)2. It is precisely with such limits, however, that Mangler’s study of improper integrals [Mangier (1951)] is concerned. Indeed, if we examine Eqs. (33) and (34) of his paper, replacing his £ with our yi, we observe that our ^-integral should be evaluated in accordance
with Mangler’s principal-value technique and thereupon assumes a per­fectly reasonable, finite value. The result implies, of course, that the self­induced normal velocity on a vortex sheet should not be infinite if it is calculated properly.

, if an indefinite integral can be found for the integrand, the answer is obtained simply by inserting the limits y = a and yx = b, provided any logarithm of (y — yi) that appears is interpreted as In |г/ — yx. The validity of Mangler’s principal value depends on the condition that the integrand, prior to letting z —* 0, be a solution of the two-dimensional Laplace equation. It is clear that this is true, in the present case, of the function that causes the singu­larity in the ^-integration of (7-29), since

Equation (7-32) provides confirmation for (5-35).

The question of exact or approximate solution of (7-28) is deferred to later sections. We note here that, when the angle of attack a and camber ordinates h(x, y) of the wing are given and the load distribution is required, (7-28) is a singular double integral equation for the unknown 7. Thus the problem is much more difficult mathematically than the corresponding thickness problem embodied in (7-14)-(7-15). On the other hand, when the loading is given and the shape of the wing to support it desired, the potential and upwash distributions are available by fairly straightforward integrations from (7-31) and (7-28), respectively. Finally, the thickness
shape g(x, y) to generate a desired symmetrical pressure distribution must be determined by solving the rather complicated integral equation which results from г-differentiation of (7-14).

We finish this section with some further discussion of the lift, drag, and nature of the wake. From (5-31), (7-19), and (7-22), we see that the difference in pressure coefficient across the wing is

. Cvi – CPu = 27. (7-33)

Using the definition of Cp,

Vi — Vu = PmUiy. (7-34)

Because the surface slopes are everywhere small, this is also essentially the load per unit plan area exerted on the wing in the positive z-direction. Since (7-34) is reminiscent of the Kutta formula (2-157), we note that 7 can be interpreted as a circulation. As shown in Fig. 7-3, let the circula­tion about the positive ^/-direction be computed around a small rectangular box C of length dx in a chordwise cross section of the wing. Since the con­tributions of the vertical sides cancel, except for terms of higher order in dx, we find that the circulation around C is

UK[ 1 + uu] dx — f7„[l + ui] dx = U«,(uu — щ) dx = f7„7 dx. (7-35)

Hence U„У(х, y) is the spanwise component, per unit chordwise distance, of the circulation bound to the wing in the vicinity of point (x, y).

U^oll + Uu]

Fig. 7-3. Interpretation of 7 by determining circulation around a circuit of length dx along the wing chord.

The lift per unit span at station у on the wing is l(y) = f [Pi — Pu] dx = pxUl ( У(х, y) dx = pxU^T(y), (7-36)


Г being the total bound circulation. If the wing tips are placed at у — ±5/2, the total lift becomes

L = Ґ12 l(y) dy = pxt/o, fbl2 Г(у) dy. (7-37)

J-Ы 2 J-bl 2

The total pitching moment, pitching moment per unit span about an
arbitrary axis, rolling moment, or any other desired quantity related to the loading may be constructed by an appropriate integration of (7-34).

Unlike a two-dimensional airfoil, the finite lifting wing does experience a downstream force (drag due to lift or “induced drag,” Zh, sometimes also called “vortex drag, ” see Chapter 9) in a subsonic inviscid flow. An easy way to compute this resistance is by examining the wake at points remote behind the trailing edge. In fact, if we observe the wing moving at speed Ux through the fluid at rest, we note that an amount of mechanical work DiUx is done on the fluid per unit time. Since the fluid is nondissi­pative and can store energy in kinetic form only, this work must ultimately show up as the value of T (cf. 2-11) contained in a length Ux of the distant wake. The nature of this wake we determine by finding its disturbance velocity potential. Over the wing region S, the ^discontinuity is calcu­lated, as in (7-26), to be

A<p(x, y) = <p(x, y, 0+)


The last line here follows from the definition of 7, (7-22), the lower limit — oo being replaced by the coordinate zle of the local leading edge since there is no w-discontinuity ahead of this point. Beyond the trailing edge on the x, ?/-plane, [uu — щ] = 0 in view of the condition of continuity of pressure. Hence

A<£>wake(2/) ~ / 7(^0, y) dx0 — ‘ (7“39)

J chord ^ qo

Since A^wake is independent of x, the wake must consist of a sheet of trailing vortices parallel to x and having a circulation per unit spanwise distance

A j-p

Ux b(y) = Ux ± (AVwake) = J • (7-40)

The complete vortex sheet simulating the lifting surface, as seen from above, is sketched in Fig. 7-4. It is not difficult to show that the Kutta – Joukowsky condition of smooth flowoff is equivalent to the requirement that the vortex lines turn smoothly into the stream direction as they pass across the trailing edge. At points far downstream the motion produced by the trailing vortices becomes two-dimensional in x, z-planes (the so – called “Trefftz plane”). Although the wake is assumed to remain flat in accordance with the small-perturbation hypothesis, some rolling up and downward displacement in fact occurs [Spreiter and Sacks (1951)]. This rolling up can be shown to have influences on the loading that are only of

third order in в and a. For the plane wake, we use (2-11) and the equality of work and kinetic-energy increment to obtain

Di = — <p dS, (7-41)


where Swake comprises the upper and lower surfaces of unit length of the wake, as seen by an observer at rest in the fluid. Clearly, dS = dy, while

rbl 2

д<Р = = 1 і jTP dyi

дП <Рг 2tU„ J-b/2 dy! (y — yi)

from the properties of infinite vortex lines. Using (7-42) and (7-39) in (7-41),


rbl 2

1 A^wake(y)<Pz{y) dy

2 J



rbl 2

/ T(y)

T—bl 2 dy 1

dy і


J-b/2 •

(y –


A more symmetrical form of (7-43) can be constructed by partial integration with respect to y. In the process, we make use of the Cauchy principal value operation at у = y and note the fact that Г(±Ь/2) = 0 in view of the continuous dropping of load to zero at the wing tips.

It is interesting that (7-37) and (7-44) imply the well-known result that minimum induced drag for a given lift is achieved, independently of the

Г = r(0)Vl – yz/(b/2)2.


Thus, if we represent the circulation as a Fourier sine series

Г = UJ> ^ An sin пв,



b a

2 cos в = у,


we find from (7-37)

L = їряиї, Ь*Аі.


However, (7-43) or (7-44) yields

n p. Ulb2 sr’ л 2

D г 7Г g 7іАП‘


details of camber and planform shape, by elliptic spanwise load distribution

n = l

7- 4 Lifting-Line Theory

The first rational attempt at predicting loads on subsonic, three-dimen­sional wings was a method due to Prandtl and his collaborators, which was especially adapted in an approximate way to the large aspect-ratio, unswept planforms prevalent during the early twentieth century. Although our approach is not the classical one, we wish to demonstrate here how the lifting-line approximation follows naturally from an application of the method of matched asymptotic expansions. Let us consider a wing of the sort pictured in Fig. 7-4, and introduce a second small parameter e.4, which is inversely proportional to the aspect ratio. Then if we write the local chord as

c(y) = eAc(y) (7-49)

and examine the matched inner and outer solutions associated with the process tA —> 0 at fixed span b, we shall be generating a consistent high aspect-ratio theory. As far as the thickness ratio, angle of attack, and camber are concerned, we assert that they are small at the outset and
remain unchanged as eA —> 0. Our starting point is therefore the problem embodied in (5-29)-(5-30), with M = 0; we study the consequences of superimposing a second limit on the situation which they describe.

Placing the wing as close as possible to the у-axis in the x, у-plane, we define new independent variables

Fig. 7-5. Coefficients of lift plotted vs. angle of attack a and total drag coeffi­cient Cd for several rectangular wings (values of A indicated on the figures). In parts (b) and (d), the same data are adjusted to a reference A = 5 by formulas based on lifting-line theory and elliptic loading. [Adapted from Prandtl, Wiesels – berger, and Betz (1921).]

The Representation of Ф in Terms of Boundary Values

Let us consider a motion of the type treated in the foregoing sections and examine the question of finding Фр at a certain point (x, y, z) arbi­trarily located in the flow field. For this purpose we turn to the reciprocal theorem (2-12), making Ф the velocity potential for the actual flow and


r лДзГ— Xi)2 + (у — Ух)2 + (z — Zx)2

Here (xi, yi, Z) is any other point, one on th; boundary for instance. See Fig. 2-1. By direct substitution in Laplace’s equation, it is easy to prove that

Подпись:image28Подпись: (2-19)vV = v2

everywhere except in the immediate vicinity of r = 0, where the La – placian has an impulsive behavior corresponding to the local violation of the requirement of continuity.

Подпись: П Fig. 2-і. Arbitrary point P in liquid flow produced by general motion of an inner boundary S.

Подпись: Small spherical surface a

If we are to use the reciprocal theorem, which requires that the Laplacians of both members vanish, we must exclude the point P from the volume V. This we do by centering a small spherical surface a around the point, and we obtain

§i%ds + §*%^-§i’!ids + §t,!iAr- <2-20>

<f S-f-2 a

Inserting the value of Ф’, we rewrite this

a S+2

-#*£0) <2-21)

S+2 <r

If we let <7 become a very small sphere, then

da = r2 da, (2-22)

where dQ is an element of solid angle such that

Over the surface of a,


Подпись: -ФР image32 Подпись: —47гФр. (2-25)

By the mean value theorem of integration, it is possible to replace the finite, continuous Ф by Фр in the vicinity of the point P to an acceptable degree of approximation. These considerations lead to


<r—*0 J J P ^/mean r—»0 J J

The Representation of Ф in Terms of Boundary Values Подпись: (2-26)

Moreover, although дФ/дп varies rapidly over <r, it will always be possible to find a bounded average value (дФ/дп)т(.ап such that

Подпись: Hence
* = -§£s;ds + §*iQ*)ds-

S+2 S+S

In the limit as the outer boundary goes to infinity, with the liquid at rest there, one can show that the integrals over 2 vanish, leaving

All the manipulations in (2-27) and (2-28) on the right-hand sides should be carried out using the variables aq, уb zi. In particular, the normal derivative is expressed in terms of these dummy variables.

The reader will be rewarded by a careful examination of some of the deductions from these results that appear in Sections 57 and 58 of Lamb (1945). For instance, by imagining an artificial fluid motion which goes on in the interior of the bounding surface S, it is possible to reexpress (2-28) entirely in terms of either the boundary values of Ф or its normal derivative. Thus, the determinate nature of the problem when one or the other of these quantities is given from the boundary conditions be­comes evident.

The quantities —1/47гг and д/дп(1/4тгг), which appear in the integrands of (2-27) and (2-28), are fundamental solutions of Laplace’s equation that play the role of Green’s functions in the representation of the velocity potential. Their names and physical significances are as follows.



The Point Source

where r may be regarded as the radial coordinate in a set of spherical coordinates having the origin at the center of the source. The radial velocity component is whereas the other velocity components vanish. Evidently we have a spherically symmetric outflow with radial streamlines. The volume efflux from the center is easily shown to equal unity. The equipotential surfaces are concentric spheres.

A negative source is referred to as a point sink and has symmetric inflow. When the strength or efflux of the source is different from unity, a strength factor H [dimensions (length)3 (time)-1] is normally applied in (2-29) and (2-30).

2. The Doublet. The derivative of a source in any arbitrary direction s is called a doublet,


In particular, a doublet centered at the point (aq, ylt zi) and oriented in the г-direction would have the velocity potential


By examining the physical interpretation of the directional derivative, we see that a doublet may be regarded as a source-sink pair of equal strengths, with the line between them oriented in the direction of the doublet’s axis, and carried to the limit of infinitesimal separation between them. As this limit is taken, the individual strengths of the source and sink must be allowed to increase in inverse proportion to the separation.

Sources and doublets have familiar two-dimensional counterparts, whose potentials can be constructed by appropriate superposition of the three­dimensional singular solutions. There also exist more complicated and more highly singular solutions of Laplace’s equation, which are obtained by taking additional directional derivatives. One involving two differenti­ations is known as a quadrupole, one with three differentiations an octu – pole, and so forth.

Thin-wing Theory

4- 1 Introduction

In this chapter we shall derive the equations of motion governing the sub – or supersonic flow around thin wings. Although this problem may be handled by simpler regular perturbation methods, we shall here instead use the method of matched asymptotic expansions for basically two reasons. First, the present method gives a clearer picture of the usual linearized formulation as that for an outer flow for which the wing in the limit collapses onto a plane (taken to be z = 0). Secondly, the similarities and dissimilarities between the airfoil problem and that for a slender body of revolution, which will be treated in the following chapter, will be more readily apparent; in fact, as will later be seen, the slender-body theory formulation can be obtained from the two-dimensional airfoil case by means of a simple modification of the inner solution.

The procedure will be carried out in detail for two-dimensional flow only, but the extension to three dimensions is quite straightforward. The first term in the series expansion leads to the well-known linearized wing theory. In this chapter, the solutions for two-dimensional flow are given as examples. Three-dimensional wings will be considered in later chapters. The case of transonic flow requires special treatment and will be deferred to Chapter 12.

4- 2 Expansion Procedure for the Equations of Motion

Anticipating the three-dimensional wing case, for which the standard choice is to orient the wing in the x, y-plane, we shall this time consider two-dimensional flow in the x, «-plane around a thin airfoil located mainly along the ж-axis with the free stream of velocity [/„ in the direction of the positive ж-axis as before. Thus, referring to Fig. 5-1, we let the location of the upper and lower airfoil surfaces be given by

zu = tfu(x) = тд(х) + вЦх) — ax zi = ifi(x) = —тд(х) + вЛ{ x) — ax,

where e is a small dimensionless quantity measuring the maximum cross­wise extension of the airfoil, r its thickness ratio, a the angle of attack, and 0 is a measure of the amount of camber. The functions g(x) and h(x)


define the distribution of thickness and camber, respectively, along the chord. It will be assumed that g and h are both smooth and that g’ and h1 are of order of unity everywhere along the chord. A blunt leading edge is thus excluded. In the limit of e —» 0 the airfoil collapses to a segment along the z-axis assumed to be located between x = 0 and x = c. We will seek the leading terms in a series expansion in e of Ф to be used as an approximation for thin airfoils with small camber and angle of attack.

For two-dimensional steady flow the differential equation (1-74) for Ф simplifies to

(a2 – Ф2Х)ФХХ + (a2 – Ф2)Фгг – 2ФхФгФхг = 0, (5-2)

where the velocity of sound is given by (1-67), which, for the present case, simplifies to

a2 = a2 – {Ф2Х + Ф2 – Vl). (5-3)

From Ф the pressure can be obtained using (1-64):

Подпись: Cv чМ21 •——– ^ (ф| + ф l-Ul) – 1 ■ (5-4)

2 al J )

The boundary conditions are that the flow is undisturbed at infinity and tangential to the airfoil surface. Hence

Thin-wing Theory

Фг/Фх = 6 0П г = «/u Фг/Фх = on z=eji.

Additional boundary conditions required to make the solution unique for a subsonic flow are that the pressure is continuous at the trailing edge (Kutta-Joukowsky condition) and also everywhere outside the airfoil. In a supersonic flow, pressure discontinuities must satisfy the Rankine – Hugoniot shock conditions. However, to the approximation considered here, these are always automatically satisfied. Strictly speaking, a super­

sonic flow with curved shocks is nonisentropic, and hence nonpotential, but the effects of entropy variation will not be felt to within the approxi­mation considered here, provided the Mach number is moderate.

We seek first an outer expansion of the form

Ф° = Г/„[Фо(ж, z) + €Ф?0, z) + • • •]. (5-6)

The factor U„ is included for convenience; in this manner the first partial derivatives of the Фп-terms will be nondimensional. Since the airfoil in the limit of € —> 0 collapses to a line parallel to the free stream, the zeroth – order term must represent parallel undisturbed flow. Thus

Ф°о = x. (5-7)

By introducing the series (5-6) into (5-2) and (5-3), and using (5-7), we obtain after equating terms of order e

(1 – М2)Ф°1хх-+ Ф?« = 0. (5-8)

The only boundary condition available for this so far is that flow perturba­tions must vanish at large distances,

Подпись: (5-9)і Фіг —» 0 for sjx2 + г2

The remaining boundary conditions belong to the inner region and are to be obtained by matching. The inner solution is sought in the form

Подпись:£/„[Фо(ж, z) + еФі(а;, z) + б2Ф’2(х, «) + •••], (5-10)

z = г/t. (5-11)

Подпись: FIG. 5-2. Airfoil in stretched coordinate system. This stretching enables us to study the flow in the immediate neighborhood of the airfoil in the limit of e —> 0 since the airflow shape then remains independent of € (see Fig. 5-2) and the width of the inner region becomes of order unity. The zeroth-order inner term is that of a parallel flow, that is, Фо = x, because the inner flow as well as the outer flow must be parallel in the limit t —> 0.

(This could of course also be obtained from the matching procedure.) From the expression for the IF-component,

Подпись: WФг = – еФІ = Ux[$Ux, z)

Подпись: (5-12)«ФггОг, г) +

we see directly that Ф} must be independent of z, say Ф = (ji(x) (which in general is different above and below the airfoil), otherwise W would not vanish in the limit of zero e. Hence (5-10) may directly be simplified to

Ф1 = U<t,[x + — є2Ф2(ж, z) + • • • ]. (5—13)

By substituting (5-13) into the differential equation (5-2) and the asso­ciated boundary condition (5-5) we find that

Фк* = 0, (5-14)

Фг = for г = Ju(x)


Фіі = for z = Ji(x).

Thus, the solution must be linear in z,

ФІ = Ї§+Мї), (5-16)

for z >7u, with a similar expression for 2 < fi. This result means that to lowest order the streamlines are parallel to the airfoil surface throughout the inner region.

The inner solution cannot, of course, give vanishing disturbances at infinity since this boundary condition belongs to the outer region. We

therefore need to match the inner and outer solutions. This can be done in two ways; either one can use the limit matching principle for W or the asymptotic matching principle for Ф. From (5-16) it follows that W’ is independent of z to lowest order, hence in the outer limit 2 = oo

Wi0 = Ut(x, oo) = e^“- (5-17)


W° = еФ°1г. (5-18)

Equating the inner limit (z = 0+) to (5-17), we obtain the following boundary condition:

*u(x, 0+)=J> (5-19)

and in a similar manner

Фї.(*,0-) =§• (5-20)

By matching the potential itself we find that

Viu(x) = Фї(*,0+). (5-21)

To determine it is necessary to go to a higher order in the outer solution.

To illustrate the use of the asymptotic matching principle we first express the two-term outer flow in inner variables,

Ф° = TJ*[x + ez) + • • • ], (5-22)

and then take the three-term inner expansion of this, namely

Ф0 = U„[x + еФ°г(х, 0+) + е2гФ°1г(х, 0+) + • ■ • ], (5-23)

which, upon reexpression in outer variables yields

Ф0 = U*[z + еФІ(х, 0+) + егФЇЛх, 0+) + • • • ]■ (5-24)

The three-term inner expansion, expressed in outer variables, reads

Подпись: x + Є$1и(х) + €2^f + e2d2u(x) +Подпись: (5-25)Подпись:

Thin-wing Theory Thin-wing Theory Thin-wing Theory


with the three-term inner expansion of the two-term outer expansion as given by (5-23) leads to

as before.

From the velocity components we may calculate the pressure coefficient by use of (5-4). Expanding in e and using (5-21) we find that the pressure on the airfoil surface is given by

Cp = ~2еФ°1х(х, 0±), (5-27)

where the plus sign is to be used for the upper surface and the minus sign for the lower one.

An examination of the expression (5-25) makes clear that such a simple inner solution cannot hold in regions where the flow changes rapidly in the x-direction as near the wing edges, or near discontinuities in airfoil surface slope. For a complete analysis of the entire flow field, these must be considered as separate inner flow regions to be matched locally to the outer flow. The singularities in the outer flow that are usually encountered at, for example, wing leading edges, do not occur in the real flow and should be interpreted rather as showing in what manner the perturbations due to the edge die off at large distances. For a discussion of edge effects on the basis of matched asymptotic expansions, see Van Dyke (1964).

In the following we will use the notation

Подпись: (5-28)6ФЇ = <p,

Thin-wing Theory Подпись: (5-29) (5-30) (5-31)

where <p is commonly known as the velocity perturbation potential (the factor Ux is sometimes included in the definition). The above procedure is easily extended to three dimensions, and for ip we then obtain the follow­ing set of linearized equations of motion and boundary conditions:

where S is the part of the x, у-plane onto which the wing collapses as e —> 0.

Fig. 5-3. Separation of thickness and lift problems.

Since the equations of motion and the boundary conditions are linear, solutions may be superimposed linearly. It is therefore convenient to write the solution as a sum of two terms, one giving the flow due to thickness and the other the flow due to camber and angle of attack (see Fig. 5-3). Thus we set


on S. It follows from (5-33) and (5-34) that <p‘ is symmetric in z whereas <pl is antisymmetric. As indicated in Fig. 5-3, represents the flow around a symmetric airfoil at zero angle of attack whereas <pl represents that around an inclined surface of zero thickness, a “lifting surface.” It will be apparent from the simple examples to be considered below that the lifting problem is far more difficult to solve than the thickness problem, at least for subsonic flow.

Another consequence of the lineariza­tion of the problem is that complicated solutions can be built up by superposition of elementary singular solutions. The ones most useful for constructing solu­tions to wing problems are those for a source, doublet and elementary horseshoe vortex. For incompressible flow the first two have already been discussed in Chap­ter 2. The elementary horseshoe vortex consists of two infinitely long vortex fila – pIG 5.4 – p^g eiementary ments of unit strength but opposite signs horseshoe vortex, located infinitesimally close together

image72along the positive ж-axis and joined together by an infinitesimal piece along the y-axis, also of unit strength (see Fig. 5-4).

The solution for this can be obtained by integrating the solution for a doublet in the ж-direction. Thus /•00

г___ 1 / г dxi _ 1 г Л x

* ~ J* (xj + y + z2)3/2 47r У2 + г2 y/x*


To extend these solutions to compressible flow the simplest procedure is to notice that by introducing the stretched coordinates

Подпись: (5-36)У = 0у, г = j8z,

Подпись: (5-37)

where /3 = /l — M2, the differential equation (5-29) transforms to the Laplace equation expressed in the coordinates x, y, z. Hence any solution to the incompressible flow problem will become a solution to the compres­sible flow if у and z are replaced by у and z. This procedure gives as a solution for the simple source

This solution could be analytically continued to supersonic flow (M > 1). However, the solution will then be real only within the two Mach cones (•y/M2 — l)r < x (see Fig. 5-5), and for physical reasons the solution within the upstream Mach cone must be discarded. Hence, since we must discard half the solution it seems reasonable that the coefficient in front of it should be increased by a factor of two in order for the total volume
output to be the same. Thus the solution for a supersonic source would read


Подпись: Z

J___________ 1_________

2ir Vz2 – (М2 – l)r2

A check on the constant will be provided in Chapter 6.

The use of the elementary solutions to construct more complicated ones is a method that will be frequently employed later in connection with three-dimensional wing theories. This method is particularly useful for developing approximate numerical theories. However, in the two-dimen­sional cases that will be considered next as illustrations of the linearized wing theory a more direct analytical method is utilized.

Interference and Nonplanar Lifting-Surface Theories

10- 1 Introduction

For a general discussion of interference problems and linearized theo­retical methods for analyzing them, the reader is referred to Ferrari (1957). His review contains a comprehensive list of references, and although it was editorially closed in 1955 only a few articles of fundamental importance seem to have been published since that date.

The motivation for interference or interaction studies arises from the fact that a flight vehicle is a collection of bodies, wings, and tail surfaces, whereas most aerodynamic theory deals with individual lifting surfaces, or other components in isolation. Ideally, one would like to have theoretical methods of comparable accuracy which solve for the entire combined flow field, satisfying all the various boundary conditions simultaneously. Except for a few special situations like cascades and slender wing-body configurations, this has proved impossible in practice. One has therefore been forced to more approximate procedures, all of which pretty much boil down to the following: first the disturbance flow field generated by one element along the mean line or center surface of a second element is calculated; then the angle-of-attack distribution and hence the loading of the second element are modified in such a way as to cancel this “inter­ference flow field” due to the first. Such interference effects are worked out for each pair of elements in the vehicle which can be expected to inter­act significantly. Since the theories are linear, the various increments can be added to yield the total loading.

There are some pairs of elements for which interference is unidirectional. Thus a supersonic wing can induce loading on a horizontal stabilizer behind it, whereas the law of forbidden signals usually prevents the stabilizer from influencing the wing. In such cases, the aforementioned procedure yields the exactly correct interference loading within the limits imposed by linearization. When the interaction is strong and mutual, as in the case of an intersecting wing and fuselage, the correct combined flow can be worked out only by an iteration process, a process which seems usually to be stopped after the first step.

Interference problems can be categorized by the types of elements involved. The most common combinations are listed below.

(1) Wing and tail surfaces.

(2) Pairs or collections of wings (biplanes or cascades).

(3) Nonplanar lifting surfaces (T – and V-tails, hydrofoil-strut combi­nations).

(4) Wing or tail and fuselage or nacelle.

(5) Lifting surface and propulsion system, especially wing and propeller.

(6) Tunnel boundary, ground and free-surface effects.

It is also convenient to distinguish between subsonic and supersonic steady flight, since the flow fields are so different in’the two conditions.

In the present discussion, only the first three items are treated, and even within this limitation a number of effects are omitted. Regarding item 4, wing-fuselage interference, however, a few comments are worth making. Following Ferrari, one can roughly separate such problems into those with large aspect-ratio, relatively unswept wings and those with highly-swept, low aspect-ratio wings. Both at subsonic and (not too high) supersonic speeds, the latter can be analyzed by slender-body methods along the lines described in Chapter 6. The wings of wider span need different approaches, depending on the Mach number [cf. Sections C, 6-11 and C, 35-50 of Ferrari (1957)]. For instance, subsonically it appears to be satisfactory to replace the fuselage with an infinite cylinder and work with two-dimensional crossflow methods in the Trefftz plane.[7] At super­sonic speeds, however, the bow wave from the pointed body may have a major influence in modifying the spanwise load distribution.

Further Examination of the Rigid, Impermeable Solid

Moving Through a Constant-Density Fluid Without Circulation

This section reviews some interesting results of incompressible flow theory which highlight the similarities between the dynamics of a rigid body and constant-density flow. They are particularly useful when cal­culating the resultant forces and moments exerted by the fluid on one or more bodies moving through it. Consider a particular solid body S of the type discussed above. Let there be a set of portable axes attached to the body with origin at O’; г is the position vector measured instantaneously from O’. Let u(<) and w(t) denote the instantaneous absolute linear and angular velocity vectors of the solid relative to the fluid at rest at infinity. See Fig. 2-2. If n is the outward normal to the surface of S, the boundary condition on the velocity potential is given in terms of the motion by


T – = [u + w X r] • n = u • n —(— и ■ (г X n) on S = 0. (2-33)

In the second line here the order of multiplication of the triple product has been interchanged in standard fashion.

Now let Ф be written as

Подпись: FIG. 2-2. Rigid solid in motion, showing attached coordinate system and linear and angular velocities relative to the liquid at rest. Ф = u – ф + ы-Х. (2-34)

Assuming the components of the vectors ф and X to be solutions of Laplace’s equation dying out at in­finity, the flow problem will be solved if these coefficients of the linear and angular velocity vectors are made to satisfy the following boundary conditions on S:

P – = г X n. (2-36) dn

Evidently ф and X are dependent on the shape and orientation of S, but not on the instantaneous magnitudes of the linear and angular veloci­ties themselves. Thus we see that the total potential will be linearly de­pendent on the components of u and w. Adopting an obvious notation for the various vector components, we write

Turning to (2-11), we determine that the fluid kinetic energy can be expressed as

T = ~ f + г)] dS

s s

= + Bv2 + Cw2 + 2 A’vw + 2B’wu + 2C’uv

+ Pp2 + Qq2 + Rr2 + 2 P’qr + 2Q’rp + 2 R’pq + 2 p[Fu + Gv + Hw) + 2q[F’u + G’v + H’w]

+ 2r[F"u + G"v + H"w]}. (2-38)

Careful study of (2-37) and the boundary conditions reveals that A, B, C, . . . are 21 inertia coefficients directly proportional to density p and dependent on the body shape. They are, however, unaffected by the instantaneous motion. Therefore T is a homogeneous, quadratic function of u, v, w, p, q, and r.

By using (2-35) and (2-36) and the reciprocal theorem (2-12), the following examples can be worked out without difficulty:

Подпись: A = A' = image37(2-39)


Here 71 and 72 are the direction cosines of the unit normal with respect to the x – and у-axes, respectively.

Further Examination of the Rigid, Impermeable Solid Подпись: (2-41)

Some general remarks are in order about the kinetic energy. First, we observe that the introduction of more advanced notation permits a sys­tematization of (2-38). Thus, in terms of matrices,

The first and last factors here are row and column matrices, respectively, while the central one is a 6 X 6 symmetrical square matrix of inertia coefficients, whose construction is evident from (2-38). In the dyadic or tensor formalism, we can express T

Ї1 — • M • u “T u • S • ш – j – 2^ • I * w. (2—42)

Here M and I are symmetric tensors of “inertias ” and “moments of inertia, ” while S is a nonsymmetrical tensor made up of the inertia coefficients which couple the linear and angular velocity components. For instance, the first of these reads

M = Aii + 5jj + Ckk + A'[ jk + kj] + B'[lsi + ik] + C'[ij + ji],


(Of course, the tensor summation notation might be used in place of dyadics, but the differences are trivial when we are working in Cartesian frames of reference.)

It is suggestive to compare (2-38) with the corresponding formula for the kinetic energy of a rigid body,

Ti = mu2 + v2 + w2] + %[p2Ixx + q2Ivy + r2Izz]

— [Iyzqr – F Izxrp + IxyPq]

+ m{x(vr — ivq) + 1j(wp — ur) + z(uq — vp)}. (2-44)

Here standard symbols are used for the total mass, moments of inertia, and products of inertia, whereas x, y, and z are the coordinates of the center of gravity relative to the portable axes with origin at O’. For many purposes, it is convenient to follow Kirchhoff’s scheme of analyzing the motion of a combined system consisting of the fluid plus the solid body. It turns out to be possible to derive Lagrangian equations of motion for this combined system, which are little more complex than those for the solid alone.

Another important point concerns the existence of principal axes. It is well known that the number of inertia coefficients for the rigid body can be reduced from ten to four by working with principal axes having their origin at the center of gravity. By analogy with this result, or by reference to the tensor character of the arrays of inertia coefficients, one can show that an appropriate rotation of axes causes the off-diagonal terms of M to vanish, while a translation of the origin O’ relative to S converts I to a diagonal, thus reducing the 21 to 15 inertia coefficients. Sections 124 to 126 of Lamb (1945) furnish the details.

For present purposes, we focus on how the foregoing results can be used to determine the forces and moments exerted by the fluid on the solid.

Further Examination of the Rigid, Impermeable Solid Further Examination of the Rigid, Impermeable Solid

It is a well-known theorem of dynamics, for a system without potential energy, that the instantaneous external force F and couple M are related to the instantaneous linear momentum £ and angular momentum Л by

Further Examination of the Rigid, Impermeable Solid Подпись: (2-47)

Furthermore, these momenta are derivable from the kinetic energy by equations which may be abbreviated



The notation for the derivatives is meaningful if the kinetic energy is written in tensor form, (2-42). To assist in understanding, we write out the component forms of (2-47),

£1 = dTi/du, £2 = dTi/dv, £3 = dTi/dw. (2-49a, b,c)

Lagrange’s equations of motion in vector notation are combining (2-45) through (2-48), as follows:

derived by




— !(£)•


The time derivatives here are taken with respect to a nonrotating, non­translating system of inertial coordinates. Some authors prefer these results in terms of a time rate of change of the momenta as seen by an observer rotating with a system of portable axes. If this is done, and if we recognize that there can be a rate of change of angular momentum due to a translation of the linear momentum vector parallel to itself, we can replace (2-45) and (2-46) with



The subscript p here refers to the aforementioned differentiation in portable axes. Corresponding corrections to the Lagrange equations are evident.

Since the foregoing constitute a result of rigid-body mechanics, we attempt to see how these ideas can be extended to the surface S moving through an infinite mass of constant-density fluid. For this purpose, we associate with £ and X the impulsive force and impulsive torque which would have to be exerted over the surface of the solid to produce the motion instantaneously from rest. (Such a combination is referred to as a “wrench.”) In view of the relationship between impulsive pressure and velocity potential, the force and torque can be written

These so-called “Kelvin impulses” are no longer equal to the total fluid momenta; the latter are known to be indeterminate in view of the non­vanishing impulses applied across the outer boundary in the limit as it is taken to infinity. Nevertheless, we shall show that the instantaneous force and moment exerted by the body on the liquid in the actual situation are determined from the time rates of change of £ and X.

To derive the required relationship, we resort to a partially physical reasoning that follows Chapter 6 of Lamb (1945). Let us take the linear force and linear impulse for illustration purposes and afterwards deduce the result for the angular quantities by analogy.

Consider any flow with an inner boundary S and an outer boundary 2, which will later be carried to infinity, and let the velocity potential be Ф. Looking at a short interval between a time t0 and a time t, let us imagine that just prior to t0 the flow was brought up from rest by a system of impulsive forces (—рф0) and that, just after t1: it was stopped by —(—рФі). The total time integral of the pressure at any point can be broken up into two impulsive pieces plus a continuous integral over the time interval.

The first integral on the right here vanishes because the fluid is at rest prior to the starting impulse and after the final one. Also these impulses make no appreciable contribution to the integral of Q2/2.

Suppose now that we integrate these last two equations over the inner and outer boundaries of the flow field, simultaneously applying the unit normal vector n so as to get the following impulses of resultant force at these boundaries:

We assert that each of these impulses must separately be equal to zero if the outer boundary 2 is permitted to pass to infinity. This fact is obvious if we do the integration on the last member of (2-57),

The integral of the term containing the constant in (2-57) vanishes since a uniform pressure over a closed surface exerts no resultant force. Further­more, the integral of Q2/2 vanishes in the limit because Q can readily be shown to drop off at least as rapidly as the inverse square of the distance from the origin. The overall process starts from a condition of rest and ends with a condition of rest, so that the resultant impulsive force exerted over the inner boundary and the outer boundary must be zero. Equation (2-58) shows that this impulsive force vanishes separately at the outer boundary, so we are led to the result

f[nf‘1+ pdtdS = 0. (2-59)


In view of (2-59), the right-hand member of (2-56) can then be inte­grated over S and equated to zero, leading to

Jjnj*1 pdtdS = f^^dt = jj[—рФі — (—рФ0)]пй5 = ki ~ 5o-

s ° ° s (2-60)

Finally, we hold t0 constant and differentiate the second and fourth mem­bers of (2-60) with respect to fi. Replacing <i with t, since it may be regarded as representing any given instant, we obtain

F = f – (2-61)

This is the desired result. It is a simple matter to include moment arms in the foregoing development, thus working with angular rather than

Подпись: (2-62)

linear momentum, and derive

where Л is the quantity defined by (2-55). The instantaneous force and instantaneous moment about an axis through the origin O’ exerted by the body on the fluid are, respectively, F and M.

If we combine with the foregoing the considerations which led to (2-52) and (2-53), we can reexpress these last relationships in terms of rates of change of the Kelvin linear and angular impulses seen by an observer moving with the portable axes,



Подпись: and Further Examination of the Rigid, Impermeable Solid Further Examination of the Rigid, Impermeable Solid

As might be expected by comparison with rigid body mechanics, fj and can be obtained from the fluid kinetic energy T. The rather artificial and tedious development in Lamb (1945) can be bypassed by carefully examin­ing (2-38). The integrals there representing the various inertia coefficients are taken over the inner bounding surface only, and it should be quite apparent, for example, that the quantity obtained by partial differentiation with respect to и is precisely the ж-component of the linear impulse defined by (2-54); thus one finds

The dyadic contraction of these last six relations is, of course, similar to (2-47) and (2-48).

Further Examination of the Rigid, Impermeable Solid Further Examination of the Rigid, Impermeable Solid

Suitable combinations between (2-65)-(2-66) and (2-61)-(2-64) can be regarded as Lagrange equations of motion for the infinite fluid medium bounded by the moving solid. For instance, adopting the rates of change as seen by the moving observer, we obtain

In the above equations, Fbody and Mbody are the force and moment exerted by the fluid on the solid, which are usually the quantities of interest in a practical investigation.

2-5 Some Deductions from Lagrange’s Equations of Motion in Particular Cases

In special cases we can deduce a number of interesting results by exam­ining (2-67)-(2-68) together with the general form of the kinetic energy, (2-38).

1. Uniform Rectilinear Motion. Suppose a body moves at constant velocity u and without rotation, so that и = 0. The time derivative terms in (2-67) and (2-68) vanish, and they may be reduced to

If body = 0, (2-69)

Mbody = – u X ~ ■ (2-70)

The fact that there is no force, either drag or lift, on an arbitrary body moving steadily without circulation is known as d’Alembert’s paradox. A pure couple is found to be experienced when the linear Kelvin impulse vector and the velocity are not parallel. These quantities are obviously parallel in many cases of symmetry, and it can also be proved that there are in general three orthogonal directions of motion for which they are parallel, in which cases the entire force-couple system vanishes. Similar results can be deduced for a pure rotation with u = 0.

2. Rectilinear Acceleration. Suppose that the velocity is changing so that u = u(t), but still w = 0. Then a glance at the general formula for kinetic energy shows that

дТ^.дТ. ЗГ ЭГ

5u 1 du ^ dv dw

= і [Am + C’v + B’w] + }[Bv – f A’w + C’u]

+ к[Cw + A’v + B’u] = M u. (2-71)

This derivative is a linear function of the velocity components. Therefore,

Fbody = – M ~ ■ (2-72)

For instance, if the acceleration occurs entirely parallel to the ж-axis, the force is still found to have components in all three coordinate directions:

рьjC’g-M’g. (2-73)

It is clear from these results why the quantities A, B, . . . , are referred to as virtual masses or apparent masses. In view of the facts that the

virtual masses for translation in different directions are not equal, and that there are also crossed-virtual masses relating velocity components in different directions, the similarity with linear acceleration of a rigid body is only qualitative.

3. Hydrokinetic Symmetries. Many examples of reduction of the system of inertia coefficients for a body moving through a constant-density fluid will be found in Chapter 6 of Lamb (1945). One interesting specialization is that of a solid with three mutually perpendicular axes of symmetry, such as a general ellipsoid. If the coordinate directions are aligned with these axes and the origin is taken at the center of symmetry, we are led to

T = ЦАи2 + Bv2 + Cw2 + Pp2 + Qq2 + Rr2]. (2-74)

The correctness of (2-74) can be reasoned physically because the kinetic energy must be independent of a reversal in the direction of any linear or angular velocity component, provided that its magnitude remains the same. It follows that cross-product terms between any of these components are disallowed. Here we can see that a linear or angular acceleration along any one of the axes will be resisted by an inertia force or couple only in a sense opposite to the acceleration itself. This is a result which also can be obtained by examination of the physical system itself.